This is what we have and what we want to do: I'll discuss the Netgear DG834Gv3 and WG602v3 as those are what I've used; your mileage may vary with different devices. The DG834G is a wireless router with built-in ADSL modem; I have version 3 and you'll need recent firmware (I've used v4.01.28) for this form of 'Wireless Distribution Services' (WDS) to work. The WG602 is a wireless access point; I have version 3 and again you'll need recent firmware (I've used v1.2.5).
The only system I could find that fitted (almost) all the above criteria, was that of 'Wireless Repeating'. Not Point-to-Point Bridging (with that the laptops can only associate with the WG602 not the DG834G). The counter-intuitive bit that had me foxed for a while was that you need to tell the DG834G and the WG602 to repeat with each other. The 'almost' in the above sentence is because with the above firmware version this only works with WEP encryption, which everyone knows can be cracked by a determined hijacker. NB: With firmware version v4.01.40, a reader has informed me that the DG834G with WPA over WDS works, please see the 'Loose ends' below.
Firstly set this up without encryption, MAC address filtering and the like. Those bits can be added later; you need to start with a clean slate. Set the name (SSID) of the wireless network the same on both the DG834G and the WG602. NB: It seems this may not be essential and you can use different SSIDs if you wish, please see the 'Loose ends' below.
Set the 'WDS Mode' on the DG834G to 'Enable Wireless Bridging and Repeating' and select the radio button 'Repeater with Wireless Client Association'. On one of the lines for the remote MAC addresses underneath, write the MAC address of the WG602.
Set the 'Access Point Mode' on the WG602 to 'Repeater with Wireless Client Association' and on the line underneath for the remote MAC address put in the MAC address of the DG834G.
Basically, that's it. I never thought this would work because if both the DG834 and WG602 exclusively talk to each other then it wasn't obvious to me how the laptops would access anything or how the desktop would be able to access the internet… Judging by the various questions that come up on the Netgear forums regarding this though it seems I'm not the only one who wanted this to work but couldn't get it to. If this article helps just one person, it was worth writing.
As I said above, get this working with no encryption first and start from the simplest setup possible. At the time this article was first written, the only encryption that the DG834G supported with repeating was WEP but a reader has informed me that WDS now works with WPA encryption using firmware v4.01.40 [thanks Simon, most appreciated]. I no longer own these devices and run this configuration so can't personally verify this, but one reader has taken the trouble to advise that he was unable to get WPA to work using the 'Repeating with Wireless Client Association' mode [thanks Thomas, most appreciated]. As always, test it with a really simple encryption key that you can't fail to type in wrongly before immediately changing that to a properly random key. A great source of random data for keys can be found on the grc.com site: Perfect Passwords.
I then set up access controls based on MAC addresses to only allow authorised machines to access the network. Finally, I turned off the SSID broadcasting on both the DG834G and the WG602. Both these are 'security by obfuscation' steps, which are not really security, but will stop Fred Bloggs next door from 'seeing' my wireless network appear automatically on his computer as an available network to use.
If you have a laptop that has a habit of changing networks, a reader has informed me that you can set the SSIDs to be different. I can't quite get my head around why this should work but it does (or at least did - I've only had one positive report about this so far, the rest haven't been able to make it work) and allows you to be certain which device you're actually connecting to [thanks Mike, most appreciated].
Can I emphasise a point about using WPA encryption if possible? Almost all devices support WPA, including for the configuration above if the latest firmware is used so there's no excuse. WEP really is the Wireless Eavesdropping Protocol now and the world should stop using it!
Thanks for reading. Hope this helped. Any questions are welcome by email at the address in the footer.
Thanks to each and every one of the following for taking the time to write about their experiences, its good to know I wasn't the only one who found this to be a problem and I'm very pleased to have been of some assistance.
A word of thanks for your posting [this article]. It [was the only one that helped] resolve problems with a WG602 I bought today as a repeater but then thought, because of various misinformed or outdated postings that DG838v3 could not be used with it unless I purchased a second 602! I only had to upgrade the firmware from netgear's site to a later version to make it provide a WDS menu (not possible before). In any case, it's working now even though it is not exactly a whopper but that's another story.
Just a quick but genuine thanks for your article on repeating with the DG834 and WG602. A great help, with both my choice of hardware and with setting it up.
Thanks for your article. It helped lots.
Just wanted to send a thank you your way. That guide is spot on and didn't take me long at all to setup the router and repeater as instructed in your tutorial.
Many thanks for the useful article. I had a number of DG834Gs lying around and wanted to set them up as repeaters, but couldn't find any information on the internet about how to do this. Your article gave me some good ideas. I've documented what needs to be done for DG834Gs at http://www.unix.ms/netgear (wireless repeating with multiple DG834Gs).
Just to let you know that you can use WPA+WPA2 wireless encryption with WDS on the DG834v3 (and Sky DG934G) routers using firmware V4.01.40.
I just wanted to say thanks for posting [this] article; I've got nothing useful to add, but it really saved the day after hours of headaches trying to get those devices doing what I wanted them to. Thanks!
Many thanks for your guide to setting up the WG602. I'm using a WG602 as a wirless range extender for a DG834Gv4.
One variation from your guide however - I have set different SSID's in the 2 Netgear boxes. It's working fine, and this makes more sense to me as I can be certain which box I'm connected to. My laptop can see both boxes, with the DG834 being very weak, but possible to connect to. I want to be certain I'm connected to the WG602.
I've set WEP encryption on both boxes with the same WEP key in both.
Also, you didn't talk about the (apparent) need to first configure a PC with a suitable TCP/IP address so it can talk to the WG602 for configuring, then resetting the PC's TCP/IP address again when configured, to use DHCP or whatever.
Many thanks for your article; it helped me wade through the misinformation from Netgear. I have a similar setup using an Netgear branded modem/router made for Optus Communications hence the designation, DG834GUv5 and a Netgear branded WG602v4 access point.
It's important to note both devices must be configured with identical details- SSID broadcast (or not), security key, MAC filtering, etc. FYI, I tried using different SSIDs but could not connect to the internet via the access point for me.
I also unsuccessfully tried WPA over WDS with Repeating with Wireless Client Association. According to 2 Australian Netgear senior techs, WPA over WDS is only allowed with Point-to-Point and Multi-point Bridging be it the DG834Gv5 or DG834GUv5 and the WG602; WPA is not currently available in Repeating with Wireless Client Association.
The DG834g I had was a v1, which didn't have the repeater function. I bought a WG602 and did the repeater set-up to a DG834N (master) as you have described on your website. Bingo! it is working well, though at the moment I am using WEP security. I will try to improve this.
Your one page instruction is so useful, compared to the dozens of manual pages. I am a novice in networks and wireless set-ups, without your information I probably would still be working on it, or given up!


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All sales are classified as a California sale and subjected to California Sales Tax.
No shopping cart means no chance of hackers getting your personal information.
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machinery, and may cause health problems.
Counter Punch Wines are sold direct to the consumer.
Why no shopping cart? See the Fine Print Department below.
deep, benefitting from moderate tannins to balance out its richness."
it was sufficient to rejuvenate the vines and power them up for the 2016 growing season.
2012 with some of the balanced power one associates with the 2014 vintage.
not be surprised if you see some sediment!
some great barrels and stayed out of the way!


Watch Iggy drive around L.A. singing before wedding dress shopping.
If you want to see Iggy Azalea perform any time soon, here's your chance.
Iggy Azalea cruised around Los Angeles singing "Fancy" and "Black Widow" then shopped for wedding dresses with James Corden as part of his Carpool Karaoke series on the Late Late Show. It's been a tough few months for Azalea, who recently canceled her Great Escape tour after pour ticket sales and creative differences with production. But to make up for it, Iggy appeared on the late night show to sing a few of her biggest hits before trying on wedding dresses.
The Australian rapper revealed her real name is Amethyst, Demi Lovato will be a bridesmaid at her wedding, she wants R. Kelly to sing at it and, if and when Corden becomes ordained, can marry her and fiance Nick Young.
we need to invent some new snacks.
im cheating on target with amazon prime.


text:
An Introduction to Mark Catesby
The Earliest Watercolor Drawings of North America's Flora and Fauna, 1712-1726
Jul 9 2016 - Sep 11 2016
Oglethorpe University Museum of Art
OUMA presents a selection of 16 limited edition reproductions of original watercolor drawings by Mark Catesby (ca.1682-1749). This is the first natural history of North America and was completed 100 years before those of John James Audubon. The 265 original watercolors were purchased by King Henry III (sic) [should be George III] in 1768 for his Royal Library at Windsor Castle. They have been kept in storage to this day, hence the clarity of color in the originals from which the facsimiles are made. Along with the 16 framed works, wall text, video and one of the four volume publications will be on view.
Still Life / Botanic

redpajama_set_name: RedPajamaCommonCrawl


text:
The Purabi are a social group of India. Along with the Pachhimi, they make up the two branches of the Tharu caste. The cast is divided into the Barhka Purabi or "upper," and the Chhutka Purabi or "lower" eastern.

Customs
An 1880 record notes:

References

Indian castes

redpajama_set_name: RedPajamaWikipedia


Nel ASA has been granted more than NOK 7.5m ($930,000) in funding for the establishment of hydrogen (H2) refuelling stations in Sønderborg and Herning.
The station in Sønderborg will service an initial fleet of fuel cell electric vehicles (FCEVs) owned by Sønderborg Forsyning, and the station in Herning will service a fleet of vehicles for Herning Kommune as well as other vehicles operated locally.
In addition to the infrastructure funding, there is also funding available for FCEVs, among those a taxi fleet. After completion of the stations, Denmark will have 12 fuelling stations in operation.
Nel ASA is increasing its ownership of the joint venture Nel-Deokyang Co. Ltd to 100%, making it a fully-owned subsidiary of Nel, and renaming it Nel Korea Co., Ltd.


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The end development will provide a great rental return or high demand selling!
YOUR MONEY IS SAFE HERE!


Foreign expedition rate was 2.4 expeditions per country.
More than one third peaks attempted by foreigners were unnamed.
Most of the foreign expeditions attempted only one peak per expedition.
The mountaineers liked single-country expeditions.
Jammu Kashmir (JK) and Uttarakhand (UK) cornered more than 80% share of the foreign expedition market.
Jammu and Kashmir and Uttarakhand also accounted for maximum (1) number of expeditions and (2) maximum number of peaks attempted.
No foreign expedition attempted peaks in Sikkim.
Market for mountaineering, a land-based adventure activity, is too small.
The 2014 witnessed a neck and neck race between Jammu and Kashmir and Uttarakhand for the share of small international mountaineering market.
Forty one (41) foreign expeditions attempted thirty three (33) peaks in the Indian Himalayas during 2014. This mountaineering business contributes a miniscule to the hard adventure revenue earned by India.
Slightly more than one third of the total attempted peaks (36.36%) were unnamed. The altitude of the unnamed peaks varied from 5,200 to 6,335 meters.
More than 90% expeditions attempted just one peak per expedition. Five percent (5%) expeditions tried two peaks. Only two percent (2%) expeditions attempted three peaks.
Europeans organized the highest number of 1-peak expeditions (28). Only two (2) expeditions planned climbing of two peaks.
Only Americans went on one 3-peak expedition.
Jammu and Kashmir recorded the highest number of one-peak expeditions (18 | 47%), followed by Uttarakhand (17 | 45%).
Mountaineers from seventeen (17) different countries organized expeditions in the Himalayas. Expedition rate was thus very low, 2.4 per country. The country-specific expeditions accounted for 80% of the total expeditions. The remaining expeditions were a joint effort of more than one country. For instance, the mountaineers from The Netherlands, Singapore, and Turkey did not attempt any peak individually.
Europeans preferred one peak per expedition. The share of two peaks per expedition was nominal. Europeans were the only mountaineers to attempt two peaks per expedition.
In Europe, Britain is the largest market for the Indian Himalayas. The country sent ten (10) expeditions. Out of which, nine (9) were 1-peak expeditions. The remaining one attempted two peaks.
Britons attempted nine peaks during ten expeditions. They attempted Mt Satopanth (Uttarakhand) twice in 2014.
They liked both Jammu and Kashmir and Uttarakhand. In both the states, Britons attempted four peaks each. However, Uttarakhand received five expeditions and JK one less.
Their preferred peak altitude ranged from 5,924 to 7,077 meters.
Americans contributed 4 (10%) expeditions to the Himalaya.
Although Americans focused on one peak per expedition, a few of them also attempted three peaks per expedition.
Joint Asian and European expeditions were 3 (7%).
Expeditions from Asia excluding India were just two (2) i.e. from Japan. These were 1-peak expeditions.
Single-country expeditions were the highest (80%). The remaining expeditions were multi-country (joint) expeditions.
Jammu and Kashmir received half of joint expeditions (4). Uttarakhand recorded 25% of joint expeditions.
Seven peaks received joint expeditions. The remaining twenty-six (26) peaks received single-country expeditions.
However, none of three Himalayan states received expeditions from all these countries. Jammu and Kashmir and Uttarakhand in fact received fifteen (15) and sixteen (16) expeditions respectively. But JK received expeditions from ten (10) countries, whereas UK from eight (8) countries.
The common foreign markets for JK and UK were France, Germany, United Kingdom, and the United States of America.
Himachal Pradesh catered to just two expeditions: one each from Japan and the United Kingdom.
The British expeditions accounted for the highest number of single-country expeditions (10 | 30%). The remaining expeditions were from the thirteen countries. Each of these countries contributed 1-3 expeditions only.
Jammu and Kashmir with 19 expeditions (46%) was the most popular Himalayan state. Uttarakhand with 18 expeditions (44%) was not very far from JK. The share of Himachal Pradesh was seven percent. Data for one expedition was not available.
As per number of peaks attempted, Jammu and Kashmir accounted for 45.45% (15) peaks. Uttarakhand accounted for 39.39% (13) peaks. The share of Himachal Pradesh was just 12.12% (4). Data for one peak was not available.
Jammu and Kashmir was the most popular state in terms of (1) number of expeditions and (2) number of peaks attempted. Uttarakhand was the second most popular state as per these two parameters. Although Himachal Pradesh was the third largest market, its share was too small. Sikkim did not receive any foreign expeditions.
Neither of the peaks received double digit expeditions.
Kun East Ridge accounted for the highest number of expeditions (5 | 11%), followed by Shivling (4 | 9%).
The remaining twenty-six (26) peaks received just one expedition each.
During 2014, there was no direct correlation between the peak altitude and the number of expeditions. We divided the peak altitude in five groups. The 6000-6499 meter group received the maximum number of expeditions (18 | 40%). The share of the 7000-7500 meter group was 28.88% (13) and that of the 6500-6999 meter group was 24.44% (11).
The 5000-5499 meter group with just one expedition was the least popular.
Surprised to see that JK nos are higher than UK.


text:
\section{\label{introduction}Introduction}

The existence of Dark Matter (DM) has been strongly suggested in various astrophysical and cosmological observations \cite{Sofue2000,PlanckCollaboration2013,Clowe2006}. Despite the fact that about $80\%$ of the matter content of the Universe is composed of DM and many experiments have been searching for possible DM signals including the Direct Detection (DD), Indirect Detection (ID) and collider searches, no confirmed non-gravitational properties of DM has been established.

Given the fact that the dominant component in proton and neutron are gluons and the light quarks, the null confirmed signal from the DD experiments motivate the possibility that DM may couple weakly to the first generation quarks if DM is a colorless particle in the first place, the origin of which may be attributed to loop nature and/or destructive cancellation. The preference of DM couplings to specific fermion flavors in the Standard Model (SM) has also been suggested in the explanation of the Galactic Center gamma-ray Excess (GCE) \cite{Goodenough_0910.2998,Hooper_1010.2752,Hooper_1110.0006, Abazajian_1207.6047,Gordon_1306.5725,Abazajian_1402.4090,Hooper_1302.6589,Hooper_1402.6703,Calore_1409.0042,Murgia-Fermi}. The gamma-ray spectrum fitting of DM annihilations into different SM final states showed that some fermion species are more suitable than others \cite{Calore2014-1411.4647}.
The sensitivities of collider searches for DM also rely on whether and/or how strongly DM couples to different quark
flavors, which would affect both the DM production cross section via light quark fusions and the visible signal types in the final states associated with the DM.

The above observations have led to the proposal of flavor structure in the interaction between DM and the SM, the so-called quark-philic (or leptophilic)
\cite{Lopez-Honorez2013,Jackson2009,Agrawal2015a,Baek2016,Bai2013,Baek2015,Chao2016-1606.07174,Hamze2014,Kile2014,Kile2013,Agrawal2014b,Agrawal2014,Yu2014,Chang2014,Lee2014,Calibbi2015,Agrawal2015,Cheung2013,DiFranzo2013,Freitas2014,Cohen2009,Schmidt2012,Chao2010,Baltz2002,Cai2014,Isidori2012,Isidori2010,Agrawal2011,Kamenik2011,Batell2011,Kile2011,Abe2016-FDM,Okawa2017-FDM,Bhattacharya:2015xha-FDM,Kumar2013,Kilic2015,Arina2016,Gomez2014,Zhang2012,Batell2013,Baek2016a}. Extensive
studies have been carried out with different assumptions of the quantum numbers of the particles participating in the
interaction including flavor charge assignment, the particle spin configurations as well as the Lorentz structure of the
interaction vertex. Many studies imposed the assumption of Minimal Flavor Violation (MFV) to evade the  strong
constraints from various flavor measurements, which assumes that the origin of flavor mixing comes only from
the SM Yukawa interaction \cite{DAmbrosio2002}. However, implementation of more general flavor structures beyond the MFV
assumption have also been discussed \cite{Chen2015,Agrawal2014a}.

Given the special role played by the top quark in the SM, top-philic DM has been an interesting scenario receiving special attention \cite{Kumar2013,Kilic2015,Arina2016,Gomez2014,Zhang2012,Batell2013,Baek2016a}.
Due to the heavy top mass and the wide mass window below the top threshold, the calculations related to the DM thermal relic are more complicated for DM lighter than the top quark, in which case co-annihilations are usually important and one needs a more complete calculation without neglecting the top quark mass. On the contrary, for light quark- or lepton-philic DM scenario, DM mass is easily above the light fermion threshold if DM is a Weakly Interacting Massive Particle (WIMP). The large top quark mass is also advantageous if one considers the scalar DM scenario, especially the real scalar DM $S$, which receives chiral suppression in both $s-$ and $p-$wave components of $\langle \sigma v\rangle_{SS\to \bar{f}f}\simeq s + p\,v^2$, where $f$ are SM fermions and $s,p\propto m_f^2$.

Another interesting point worth emphasizing is that the absence of valence top quark in nucleons makes DM-gluon scattering at loop level crucial and the future improvement of the DD sensitivities very important.
In our previous work \cite{Baek2016a} we studied a phenomenologically motivated top-philic DM model in which the DM is a SM singlet real scalar and interacts exclusively with the right-handed (RH) top quark through a vector-like fermion mediator. We found that the projected XENON1T
experiment may test the new Yukawa coupling to $\mathcal{O}(1)$ and the future LZ project may further test to
$\mathcal{O}(0.5)$. The loop nature of DM-nucleon scattering in the top-philic DM scenario manifests itself in the challenges in the direct detection.

Interestingly, in an opposite direction, \cite{Giacchino:2015hvk} has presented a comprehensive study on a scalar DM which interacts only with light quarks $-y S \overline{\psi_L} q_R + h.c.$, where $q=u,d$ and $\psi$ is a vector-like mediator carrying the same quantum number as $q_R$. Considering that the DM self-annihilation $\langle \sigma v\rangle_{SS\to \bar{q}q}\simeq d\, v^4 $ is $d$-wave suppressed due to $m_q\to 0$, \cite{Giacchino:2015hvk} studied two classes of higher-order processes: the internal Bremsstrahlung of a gauge boson $V=\gamma, Z, g$ in $SS\to q\bar{q}V$, and $SS \to VV'$ induced via 1-loop box diagram. Furthermore, non-perturbative Sommerfeld corrections to the co-annihilation processes $\psi\psi, \overline{\psi}\,\overline{\psi}, \overline{\psi}\psi \to {\rm SM}$ are also calculated. Other phenomenology including DM direct/indirect detections and collider signals are also investigated in \cite{Giacchino:2015hvk}.

In this work we extend our previous study \cite{Baek2016a} focusing on DM-top quark interaction and \cite{Giacchino:2015hvk} focusing on DM-light quark interactions, by including the DM-charm quark interaction in the following Lagrangian
\begin{eqnarray}
\mathcal{L} \supset -y_1 S \overline{\psi_L} u_R -y_2 S \overline{\psi_L} c_R  -y_3 S \overline{\psi_L} t_R +  h.c.,
\label{eq-L}
\end{eqnarray}
where $S$ is the SM singlet real scalar DM and $\psi$ is a vector-like mediator carrying the same quantum number as $u_R,c_R,t_R$. We use $y_1,y_2,y_3$ to denote the new Yukawa couplings.

Note that we will not repeat the quite complete analyses in \cite{Giacchino:2015hvk} for $S-u$ coupling $y_1$, especially the complicated higher-order calculations of DM annihilations. Nevertheless we highlight the scale effects in calculating DM-nucleon scattering rate $\sigma^{\rm SI}_{p}$ due to wide mass gaps existing in $u_R,c_R,t_R$, including the renormalization group equation (RGE) and heavy quark threshold matchings. Our observations that $S-u$ coupling $y_1$ receives strong constraints from the current DM direct detection experiments agrees well with \cite{Giacchino:2015hvk}, while we regard our analysis on scale effects as a complementarity to \cite{Giacchino:2015hvk} in which destructive interferences in effective DM-nucleon coupling will be moderately shifted.

On the other hand, since we are considering a scalar DM, apart from the phenomenology-irrelevant DM self interaction $\lambda_S S^4$, the Higgs portal interaction $\lambda_{SH}S^2H^2$ would naturally exist \cite{Silveira:1985rk,McDonald:1993ex,Burgess:2000yq}. The Higgs portal has been extensively studied in the literature and strong constraints on it have been found from the measurement of invisible Higgs decay and DM direct detection experiments (see e.g. \cite{Cline2013,Beniwal2015-HP}). Both \cite{Baek2016a} and \cite{Giacchino:2015hvk} assumed the Higgs portal to be small and focused on the new interactions mediated by the vector-like $\psi$ portal. However, $\lambda_{SH}$ can receive high order corrections starting at 1-loop level from Eq.(\ref{eq-L}). In this work we upgrade our previous analysis by calculating the radiatively generated Higgs portal $\lambda_{SH}^{\rm 1PI}$ at $\mathcal{O}(y_3^2 \,y_t^2)$ with $y_t$ being the SM top Yukawa. This correction $\lambda_{SH}^{\rm 1PI}$ is negligible in \cite{Giacchino:2015hvk} for $S-u$ interaction due to $y_q\to 0$, while it turns out to be significant with $S-t$ interaction and puts strong constraints on $y_3$ if $\lambda_{SH}=0$ at tree-level is assumed. The general case with free parameter $\lambda_{SH}\neq 0$ at tree-level, however, will bring the total Higgs portal contributions back to the well-studied scenario. Therefore, except for using $\lambda_{SH}^{\rm 1PI}$ to illustrate the interesting interplay between Higgs portal and vector-like portal, we will still assume that the total Higgs portal strength up to 1-loop level is small compared to the vector-like portal which can be easily realized by turning the tree-level free parameter $\lambda_{SH}$.

After the general analysis and observing that both the $S-u$ coupling $y_1$ and the Higgs portal have been well discussed in the existing literature and strongly constrained, especially from the DM direct detection results, their existence would make it difficult to observe the effects of DM-heavy quark interactions $y_2,y_3$ through the new vector-like $\psi$ portal. In the latter part of this work we will focus on the DM-heavy quark $c,t$ interactions by imposing
\begin{eqnarray}
\lambda_{SH}^{\rm ren.}=0,
\end{eqnarray}
which is the total renormalized Higgs portal up to 1-loop level and
\begin{eqnarray}
y_1=0.
\end{eqnarray}
In this case the new Lagrangian we consider is reduced to
\begin{eqnarray}
\mathcal{L} \supset -y_2 S \overline{\psi_L} c_R  -y_3 S \overline{\psi_L} t_R +  h.c..
\label{eq-L-tc}
\end{eqnarray}

This paper is organized as follows. In Section \ref{section-Modeldescription} we describe our model in more detail. In Section \ref{section-DD-review} we briefly review the general framework of calculating the DM-nucleon scattering rate. In Section \ref{section-DD} we perform a careful discussion of the direct detection signals in this model including both scalar-type and twist-2 operator contributions, which are missed in some papers.
Starting from Section \ref{section-DM-heavy-quark} we concentrate on DM interactions with heavy quarks $c,t$ with the alleviation of direct detection constraints. In Section \ref{section-omg} we explore the model parameter space to produce a correct DM thermal relic. In Section \ref{section-DD-heavy} we investigate the XENON1T constraints on the model's mass plane.
In Section \ref{section-ID} we consider the constraints from DM indirect detection signals from Fermi gamma-ray observations of dwarf galaxies. In Section \ref{section-FCNC} we discuss the predictions of top quark FCNC observables in this model. In Section \ref{section-collider} we study the collider signals of this model using the $36\, fb^{-1}$ data at LHC.
We combine the various results in Section \ref{section-combined} and present our conclusion in Section \ref{section-conclusion}.

\section{Model description}
\label{section-Modeldescription}

We extend the SM with a real scalar singlet DM $S$ which couples exclusively to the $SU(2)_L$ singlet up-type quarks $u_R,c_R,t_R$ via a vector-like fermion mediator $\psi$. The new interactions beyond the SM are
\begin{eqnarray}
\label{eq-L-full}
{\mathcal L_{\rm new}} &=&  {\mathcal L_{\rm fermion}} + {\mathcal L_{\rm scalar}} + {\mathcal L_{\rm Yukawa}},\\ \nonumber
{\mathcal L_{\rm fermion}} &=& \bar{\psi}(i \slashed D-m_\psi)\psi,\\ \nonumber
{\mathcal L_{\rm scalar}} &=& \frac{1}{2}\partial^\mu S \partial_\mu S - \frac{1}{2}m_S^2 S^2 - \frac{1}{4!} \lambda_S S^4 - \frac{1}{2}\lambda_{SH} S^2 H^2, \\ \nonumber
{\mathcal L_{\rm Yukawa}} &=& -y_1 S \overline{\psi_L} u_R -y_2 S \overline{\psi_L} c_R -y_3 S \overline{\psi_L} t_R +  h.c.,
\end{eqnarray}
where $\psi$ carries the same gauge quantum number as $u_R,c_R,t_R$ and $D_\mu$ is the $SU(3)\times SU(2)\times U(1)$ covariant derivative in the SM. We assume $m_S < m_\psi$ to meet the DM scenario and impose an odd $Z_2$ parity to $S,\psi$ to stabilize the DM. Because of this $Z_2$ parity and the chosen gauge quantum numbers for $S,\psi$, only $u_R,c_R,t_R$ sector in the SM are involved in the new Yukawa interaction indicated by ${\mathcal L_{\rm Yukawa}}$ in Eq.(\ref{eq-L-full}). This $Z_2$ parity also forbids the mass mixing between DM $S$ and SM Higgs $h$. Similarly, the mixings between $\psi$ and SM quarks $u,c,t$ are also forbidden which implies that the results of LHC searches for exotic heavy quarks $Q'$ do not apply, as the targeted decay channels $Q' \to Q+h/Z$ in those searches are based on the mass mixing (see e.g. \cite{CMSCollaboration2017-VLQuark-1,CMSCollaboration2017a-VLQuark-2,TheATLASCollaboration2017-VLQuark-1,ATLASCollaboration2017-VLQuark-2}).

Compared to our previous study \cite{Baek2016a} focusing on a phenomenologically motivated top-philic DM, Eq.(\ref{eq-L-full}) including all $u_R,c_R,t_R$ introduces several interesting changes:
\begin{itemize}
\item  DM-nucleon scattering in the direct detection is a low energy non-relativistic process occurring at $\mu_{\rm had}\sim 1\, {\rm GeV}$, while other DM phenomenology can happen at much higher energy scales such as the thermal freeze-out in the early Universe at $\mathcal{O}(T_f)$ with $T_f\simeq m_{\rm DM}/20$ and collider detections at $\mathcal{O}(1)$ TeV. Although one has the freedom of choosing at which scale to match the ultraviolet (UV)-complete Lagrangian Eq.(\ref{eq-L-full}) to an effective field theory (EFT) between DM and the SM, this EFT at $\mu_{\rm EFT}$ still needs to be matched to another one at $\mu_{\rm had}\sim 1\, {\rm GeV}$ with only a few active degree of freedom ({\it d.o.f}) including DM, light quarks and gluons. If there is a sizable gap between $\mu_{\rm EFT}$ and $\mu_{\rm had}$, e.g. $\mu_{\rm EFT}\sim m_Z$, the renormalization group equation (RGE) effects and heavy quark threshold matching with respect to (w.r.t) $c_R,t_R$ may be important.
\item Since gluon is the most abundant constituents inside nucleon, DM-gluon interactions generated through loops may also be important despite the loop factor suppression. In this case, loops connecting DM and gluon involving $u_R,c_R,t_R$ should be distinguished in the sense of short-distance and long-distance parts characterized by different energy scales \cite{Hisano2010-Gluon}.
\item More annihilation channels $\overline{U_i}U_j+\overline{U_j}U_i$ with $U_{i,j}=u,c,t$ and $i,j=1,2,3$ are available to produce the observed DM relic abundance $\Omega_{\rm DM}h^2\sim0.12$, which are helpful when DM mass is below top quark threshold and $SS\to\bar{t}t$ is kinematically difficult to open.
\item Flavor Changing Neutral Current (FCNC) processes of top quark can be generated at both tree level $t\to \psi^{(*)}S \to u/c+SS$ and loop level $t\to u/c+\gamma/g/Z$, which can provide additional tests of this model.
\end{itemize}

\section{Brief review of DM-nucleon effective theory}
\label{section-DD-review}
We start the discussion with the DM-nucleon scattering rate in direct detection experiment. We follow the effective theories of DM-nucleon interaction constructed in \cite{Hisano2015-DMEFT,Hisano2015-GluonWino} which carefully distinguishes long-distance (LD) and short-distance (SD) components in the loop-generated DM-gluon coupling \cite{Hisano2010-Gluon}.

To calculate the DM-nucleon scattering rate at $\mu_{\rm had}\sim 1\, {\rm GeV}$, the following steps are included:
\begin{itemize}
\item First, one integrates out the mediator particles and constructs the effective theory for the DM, quarks, and gluons at a certain scale $\mu_{\rm EFT}$. The effective interactions usually consist of higher-dimensional operators.
\item Second, the Wilson coefficients (WC) of the effective operators are evolved according to the RGE down to $\mu_{\rm had}\sim 1\, {\rm GeV}$ where the nucleon matrix elements of the operators are evaluated. In this process one usually needs to pass several heavy quark thresholds where the corresponding heavy quarks are also integrated out.
\item Finally, one expresses the DM-nucleon effective coupling in terms of the Wilson coefficients and the nucleon matrix elements. Then the scattering cross sections are obtained from this DM-nucleon effective coupling.
\end{itemize}
In the first step, the effective theory for the DM, quarks, and gluons can be expressed as
\begin{equation}
\label{eq-Leff-master}
\mathcal{L}_{\rm eff}
=\sum_{p=q,g}C^p_S\mathcal{O}^p_S
+\sum_{p=q,g}C^p_{T_2}\mathcal{O}^p_{T_2}~,
\end{equation}
with
\begin{align}
\label{eq-twist-0}
\mathcal{O}^q_S&\equiv   S^2 m_q\bar{q}q~, \quad\quad\quad\quad\quad \,\,  \mathcal{O}^g_S\equiv
 S^2  \frac{\alpha_s}{\pi}G^{A\mu\nu}G^A_{\mu\nu}~,\nonumber\\
\mathcal{O}^q_{T_2}&\equiv\frac{1}{m_S^2} S i \partial^\mu i
 \partial^\nu S \mathcal{O}^q_{\mu\nu}~, \quad \mathcal{O}^g_{T_2}\equiv\frac{1}{m_S^2} S i\partial^\mu i\partial^\nu S \mathcal{O}^g_{\mu\nu}~.
\end{align}
In the above, $\mathcal{O}^q_S,\mathcal{O}^g_S$ are scalar-type operators while ${\cal O}^q_{\mu\nu}$ and ${\cal O}^g_{\mu\nu}$ are the twist-2 operators of quarks and gluon defined by
\begin{align}
\label{eq-twist-2}
 {\cal O}^q_{\mu\nu}&\equiv \frac{1}{2}\overline{q}i\biggl(
D_\mu^{}\gamma_\nu^{} +D_\nu^{}\gamma_\mu^{}-\frac{1}{2}g_{\mu\nu}^{}
\slashed{D}\biggr)q~,\nonumber \\
{\cal O}^g_{\mu\nu}&\equiv
G^{A\rho}_{\mu} G^{A}_{\nu\rho}-\frac{1}{4}g_{\mu\nu}^{}
G^A_{\rho\sigma}G^{A\rho\sigma}~.
\end{align}
Then one performs the evolutions of Wilson coefficients according to RGE. For scalar-type interactions we have
\begin{align}
C^q_S(\mu)&= C^q_S(\mu_0)-4C^g_S(\mu_0)
\frac{\alpha^2_s(\mu_0)}{\beta(\alpha_s(\mu_0))}
(\gamma_m(\mu)-\gamma_m(\mu_0)) ~,\label{eq:cqsrge} \\
C^g_S(\mu)&=\frac{\beta(\alpha_s(\mu))}{\alpha_s^2(\mu)}
\frac{\alpha_s^2(\mu_0)}
 {\beta (\alpha_s(\mu_0))}C^g_S(\mu_0)
~,
\end{align}
where the beta-function of $\alpha_s$ and the anomalous dimension of quark mass operator are given by
\begin{align}
 \beta(\alpha_s)=(2b_1)\frac{\alpha_s^{2}}{4\pi}+(2b_2)\frac{\alpha_s^3}{(4\pi)^2} ~, \quad\quad
  \gamma_m&=-6C_F \frac{\alpha_s^{}}{4\pi}  ~,
\end{align}
with
\begin{equation}
 b_1=-\frac{11}3N_c+\frac23 N_f~, \quad\quad
  b_2=-\frac{34}{3}N_c^2 +\frac{10}{3}N_cN_f+2C_F N_f~.
\label{b1b2}
\end{equation}
In the above expressions, $N_c=3$ is the number of $SU(3)$ colors, $N_f$ denotes the number of active quark
flavors in an effective theory and $C_F$ is the quadratic Casimir invariant  $C_F\equiv \frac{N_c^2-1}{2N_c}$.

For the twist-2 type interactions the RGEs are
\begin{equation}
 \mu \frac{d}{d\mu}(C^q_{T_i}, C^g_{T_i})=
(C^q_{T_i}, C^g_{T_i})~ \Gamma_{\rm T}~,
\end{equation}
with $i=1,2$ in Eq.(\ref{eq-twist-0}) and $\Gamma_{\rm T}$ being a $(N_f+1)\times (N_f+1)$ matrix:
\begin{equation}
 \Gamma_{\rm T}=
\begin{pmatrix}
 \gamma_{qq}&0&\cdots&0&\gamma_{qg}\\
 0 &\gamma_{qq}&&\vdots&\vdots\\
\vdots&&\ddots&0&\vdots\\
0&\cdots&0&\gamma_{qq}&\gamma_{qg}\\
\gamma_{gq}&\cdots&\cdots&\gamma_{gq}&\gamma_{gg}
\end{pmatrix}
~,
\end{equation}
where
\begin{align}
 \gamma_{qq}&=\frac{16}{3}C_F \, \frac{\alpha_s}{4\pi}, \quad  \gamma_{qg}=\frac{4}{3} \, \frac{\alpha_s}{4\pi}~,\nonumber \\
 \gamma_{gq}&=\frac{16}{3}C_F \, \frac{\alpha_s}{4\pi}, \quad  \gamma_{gg}=\frac{4}{3}N_f \, \frac{\alpha_s}{4\pi}~.
\end{align}
The solution of twist-2 operator RGEs can be found in \cite{Hill2014-I,Hill2014a-II}.

Before reaching $\mu_{\rm had}\sim 1\, {\rm GeV}$, we need to pass the heavy quark thresholds such as $\mu_b\sim m_b$ and $\mu_c\sim m_c$. Taking bottom quark as an example, the matching condition we took are
\begin{equation}
 \frac{1}{\alpha_s(\mu_b)|_{N_f=4}}
= \frac{1}{\alpha_s(\mu_b)|_{N_f=5}}
+\frac{1}{3\pi}\ln\biggl(\frac{\mu_b}{m_b}\biggr) ~,
\end{equation}
and
\begin{align}
 C^q_S(\mu_b)|_{N_f=4}&= C^q_S(\mu_b)|_{N_f=5}~,\nonumber \\
[\alpha_s C^g_S](\mu_b)|_{N_f=4}&=
-\frac{\alpha_s(\mu_b)}{12}
\left[
1+\frac{\alpha_s(\mu_b)}{4\pi}
\left(11+\frac{2}{3}\ln\frac{m_b^2}{\mu_b^2}\right)
\right]
C^b_S(\mu_b)|_{N_f=5}
\nonumber \\ &
+
\left[
1+\frac{\alpha_s(\mu_b)}{4\pi}
\frac{2}{3}
\ln\frac{m_b^2}{\mu_b^2}
\right]
[\alpha_s C^g_S](\mu_b)|_{N_f=5}~,\nonumber \\
C^q_{T_i}(\mu_b)|_{N_f=4}&= C^q_{T_i}
(\mu_b)|_{N_f=5}~,\nonumber \\
 C^g_{T_i}(\mu_b)|_{N_f=4}&=
\left[
1+\frac{\alpha_s(\mu_b)}{4\pi}
\frac{2}{3}
\ln\frac{m_b^2}{\mu_b^2}
\right]
C^g_{T_i}(\mu_b)|_{N_f=5}
+\frac{\alpha_s(\mu_b)}{4\pi}\frac{2}{3}\ln \frac{m_b^2}{\mu_b^2}
C^b_{T_i}(\mu_b)|_{N_f=5}
~,
\label{eq:quarkthmatch}
\end{align}
where we have included the next-to-leading order $\alpha_s$ contribution to $C_S^g$ since the effect is known to be large \cite{Djouadi2000-GluonDM}, especially for charm quark.
In our calculation, we take $\mu_b=m_b$ and $\mu_c=m_c$ for simplification.

Finally, the effective interaction between scalar DM and a nucleon is defined by
\begin{eqnarray}
\mathcal{L}^{(N)}_{{SI}}=f_N S^2 \overline{N}N~,
\end{eqnarray}
where
\begin{align}
\label{eq-fN}
 f_N/m_N&= \sum_{q=u,d,s}C^q_S(\mu_{\text{had}}) f^{(N)}_{T_q}
-\frac{8}{9}C^g_S(\mu_{\text{had}}) f^{(N)}_{T_G}
\nonumber \\
&+\frac{3}{4}\sum_{q}^{N_f}
C^q_{T_2}(\mu)[q(2;\mu)+\overline{q}(2;\mu)]
-\frac{3}{4}C^g_{T_2}(\mu)g(2;\mu)~,
\end{align}
and
\begin{align}
\label{eq-form-factor}
f_{T_q}^{(N)} &\equiv \langle N|m_q \bar{q}q|N\rangle/m_N ~, \nonumber \\
q(2;\mu) = \int^1_0 dx~ x\ q(x,\mu)~, ~~ \bar{q}(2;\mu) &= \int^1_0 dx~ x\ \bar{q}(x,\mu)~, ~~ g(2;\mu) = \int^1_0 dx~ x\ g(x,\mu)~,
\end{align}
with $q(x,\mu)$, $\bar{q}(x,\mu)$ and $g(x,\mu)$ being the PDFs of quarks, anti-quarks and gluon at factorization scale $\mu$, respectively.

Note that the light quark mass fractions $f_{T_q}^{(N)}$ belong to the non-perturbative QCD region and are obtained from the lattice QCD simulations. Accordingly, one needs to use the Wilson coefficients $C^q_S(\mu_{\text{had}})$ at $\mu_{\rm had}\sim 1\, {\rm GeV}$. On the contrary, twist-2 matrix elements $q(2;\mu), \bar{q}(2;\mu), g(2;\mu)$ can be calculated perturbatively using parton PDFs to other scales \cite{Hisano2015-DMEFT,Hisano2015-GluonWino,Hisano2010-Gluon,Hill2014-I,Hill2014a-II,Drees1993-DM}, thus one can choose a convenient scale applied to both the twist-2 Wilson coefficients and matrix elements, e.g. simply the scale $\mu_{\rm EFT}$ of the first-step EFT is defined. In this work we use the same the nucleon matrix elements as those given in \cite{Hisano2010-Gluon,Hill2014a-II}.

Since the DM particle is a scalar in our model, only spin-independent (SI) DM-nucleus interaction is generated. The cross section can be expressed by
\begin{equation}
 \sigma =\frac{1}{\pi}\biggl(\frac{m_T}{m_S+m_T}\biggr)^2
|n_pf_p+n_nf_n|^2~,
\end{equation}
where $m_T$ is the mass of the target nucleus and $n_p,n_n$ are the numbers of protons and neutrons in the nucleus, respectively.

\section{Direct detection in our model}
\label{section-DD}

In this section we choose $\mu_{\rm EFT}=m_Z$ and extract the Wilson coefficients in Eq.(\ref{eq-Leff-master}) from the UV-complete Lagrangian in Eq.(\ref{eq-L-full}) up to $\mathcal{O}(y_i^2)$. The active SM {\it d.o.f}s at this scale include $u,c,g$, while top quark and vector-like fermion $\psi$ are integrated out. Then we include RGE effects and heavy quark threshold matching to calculate DM-nucleon scattering rate at $\mu_{\rm had}\sim 1\, {\rm GeV}$.

\begin{figure}[h]
  \centering

  \includegraphics[width=13cm]{fig-DD-Diagram.pdf}
  \caption{Feynman diagrams used for calculating the Wilson coefficients, at the order of $\mathcal{O}(y_i^2)$, of the effective operators in Eq.(\ref{eq-Leff-master}) when choosing $\mu_{\rm EFT}=m_Z$. We refer to diagrams mediated by the SM Higgs $h$ as Higgs portal, while denoting others as vector-like $\psi$ portal.}
  \label{fig-DD-diagram}
\end{figure}

Apart from the conventional $t$-channel Higgs-mediated DM-nucleon scattering at tree-level for scalar DM, the new Feynman diagrams we consider in our model are presented in Fig.\ref{fig-DD-diagram}, all of which generate Wilson coefficients in Eq.(\ref{eq-Leff-master}) at the order of $\mathcal{O}(y_i^2)$ in terms of the new Yukawa couplings in Eq.(\ref{eq-L-full}).  We will still identify the $t$-channel Higgs-mediated diagrams to be the Higgs portal (HP), while calling the others the vector-like $\psi$ portal (VLP). Note that the HP diagrams only generate scalar-type interactions in Eq.(\ref{eq-Leff-master}) while VLP generates both scalar-type and twist-2 operators \cite{Hisano2015-GluonWino}.

\subsection{Higgs portal at $\mathcal{O}(y_i^2)$}
\label{WC-HP}

Since we are considering a scalar DM, apart from the phenomenologically irrelevant DM self interaction $\lambda_S S^4$, the Higgs portal interaction $\lambda_{SH}S^2H^2$ would naturally exist \cite{Silveira:1985rk,McDonald:1993ex,Burgess:2000yq}. The Higgs portal has been extensively studied in the literature and strong constraints on it have been found from the measurements of invisible Higgs decay and DM direct detection experiments (see e.g. \cite{Cline2013,Beniwal2015-HP}). Even if one sets $\lambda_{SH}=0$ in the Lagrangian Eq.(\ref{eq-L-full}), it will receive loop corrections starting from 1-loop level in our model, as shown in the 1PI triangle vertex in Fig.\ref{fig-DD-diagram}.
\begin{eqnarray}
\label{eq-lamSH-1PI}
\lambda_{SH}^{\rm1PI} &=& \sum_{k=1,2,3} (-8) y_3^2 \, N_c \, (\frac{m_{U_k}}{v})^2  \frac{1}{16\pi^2} \Big( 4 C_{00} + m_S^2(C_{11}+C_{22}) - 2 m_S^2 C_{12}  \Big), \\
C_{ij} &\equiv& C_{ij}(m_S^2, 0, m_S^2; m_\psi^2, m_{U_k}^2, m_{U_k}^2),
\end{eqnarray}
where $N_c=3, U_{k=1,2,3}=\{u,c,t\}$ and $v\simeq 246$ GeV is the SM vacuum expectation value (vev).
In the more general case with $\lambda_{SH}\neq 0$ in the Lagrangian Eq.(\ref{eq-L-full}), one would need a more complete renormalization for $\lambda_{SH}$ at 1-loop level
\begin{eqnarray}
\label{eq-lamSH-FullRen}
\lambda_{SH}^{\rm ren.} &=& \lambda_{SH} + \delta \lambda_{SH} + \lambda_{SH}^{\rm1PI}  + \lambda_{SH} (\frac{1}{2} \delta Z_h + \delta Z_S + \delta v),
\end{eqnarray}
where $\delta Z_h, \delta Z_S$ are field renormalization strength constants in on-shell renormalization scheme for Higgs and scalar DM, while $\delta v$ is the counterterm for the SM vev. The full calculations in usual Higgs portal for $\delta Z_h, \delta Z_S, \delta v$ can be found in, e.g. \cite{Kanemura2004-REN1,Kanemura2016-REN2}, while $\delta Z_S$ in our model receives the following additional contributions
\begin{eqnarray}
\label{eq-lamSH-ZS}
\delta Z_S^{\rm new} &=& \sum_{k=1,2,3} 2\, y_k^2\, N_c\, \frac{1}{16 \pi^2} \Big(  4 B'_{00} + (B_{11}+B_1) + m_S^2 (B'_{11}+B'_1)  \Big),\\
B^{(\prime)}_{i,ij} &\equiv& B^{(\prime)}_{i,ij} (m_S^2; m_\psi^2, m_{U_k}^2).
\end{eqnarray}

In the numerical calculations we impose modified minimal subtraction ($\overline{{\textrm{MS}}}$) scheme by choosing $\delta \lambda_{SH} = (...) \Delta$ to eliminate all $\Delta = 1/\epsilon -\gamma_E + \ln 4 \pi$ in $\lambda_{SH}^{\rm ren.}$. Note that the minimal subtraction scheme is also a common choice, e.g. as utilized in \cite{Kanemura2004-REN1,Kanemura2016-REN2}. We also set the artificial scale originating from dimensional regularization (DR) to be $\mu_{\rm DR}=m_\psi$ throughout this work. By denoting the finite parts in $\lambda_{SH}^{\rm1PI}, \delta Z_h, \delta Z_S$ using the notation $Fin.[...]$, one can rewrite $\lambda_{SH}^{\rm ren.}$ as
\begin{eqnarray}
\label{eq-lamSH-FullRen-Finite}
\lambda_{SH}^{\rm ren.} &=& \lambda_{SH} \Big(1+ Fin.[\frac{1}{2} \delta Z_h + \delta Z_S + \delta v] \Big) + Fin.[\lambda_{SH}^{\rm1PI}].
\end{eqnarray}
When combined with the Higgs-quark and Higgs-gluon interactions in Fig.\ref{fig-DD-diagram}, the Wilson coefficients in Eq.(\ref{eq-Leff-master}) from Higgs portal are:
\begin{eqnarray}
\label{eq-WC-HP}
C_{S,{\rm HP}}^{q} = \frac{\lambda_{SH}^{\rm ren.}}{m_h^2}, \quad \quad C_{S,{\rm HP}}^{g} = -\frac{\lambda_{SH}^{\rm ren.}}{12 m_h^2},
\end{eqnarray}
where $q=u,d,s,c,b$ and we use $m_h=125$ GeV. Note that there is no $\frac{\alpha_s}{\pi}$ factor in $C_{S,{\rm HP}}^{g}$ which has been absorbed in the definition of scalar-type gluon operator $\frac{\alpha_s}{\pi} G^{A\mu\nu}G^A_{\mu\nu}$ in Eq.(\ref{eq-twist-0}), as clarified in \cite{Hisano2015-GluonWino}.

In this work we will consider two special choices for $\lambda_{SH}$ in Eq.(\ref{eq-lamSH-FullRen-Finite}).
\begin{itemize}
\item Case-I:
\begin{eqnarray}
\label{eq-WC-lamSH-Case-I}
\lambda_{SH} &=& 0,\\
\lambda_{SH}^{\rm ren.} &=& Fin.[\lambda_{SH}^{\rm1PI}].
\end{eqnarray}
This choice can provide a direct comparison to the vector-like $\psi$ portal contribution to DM-nucleon scattering at the same order $\mathcal{O}(y_i^2)$,
\begin{eqnarray}
\label{eq-WC-HP-Case-I}
C_{S,{\rm HP}}^{q} = \frac{Fin.[\lambda_{SH}^{\rm 1PI}]}{m_h^2} \propto y_3^2, \quad \quad C_{S,{\rm HP}}^{g} = -\frac{Fin.[\lambda_{SH}^{\rm 1PI}]}{12 m_h^2}  \propto y_3^2,
\end{eqnarray}
where we ignored the $U_1=u, U_2=c$ contributions in Eq.(\ref{eq-lamSH-1PI}) for $\lambda_{SH}^{\rm 1PI}$ due to the small $m_u, m_c$.
\item Case-II:
\begin{eqnarray}
\label{eq-WC-lamSH-Case-II}
\lambda_{SH} &=& - \frac{Fin.[\lambda_{SH}^{\rm1PI}]}{1+ Fin.[\frac{1}{2} \delta Z_h + \delta Z_S + \delta v]}, \\
\lambda_{SH}^{\rm ren.} &=& 0.
\end{eqnarray}
In this case, the Higgs portal at 1-loop level vanishes in the effective theory defined at $\mu_{\rm EFT}= m_Z$
\begin{eqnarray}
\label{eq-WC-HP-Case-II}
C_{S,{\rm HP}}^{q} =  C_{S,{\rm HP}}^{g} = 0.
\end{eqnarray}
\end{itemize}

\subsection{Vector-like $\psi$ portal at $\mathcal{O}(y_i^2)$}
\label{WC-VLP}

In Fig.\ref{fig-DD-diagram}, the DM-quark Wilson coefficients in Eq.(\ref{eq-Leff-master}) from vector-like $\psi$ portal at $\mathcal{O}(y_i^2)$ have been given in \cite{Hisano2015-DMEFT} as follows:
\begin{eqnarray}
\label{eq-WC-VLP-tree-matching}
C_{S,{\rm VLP}}^{u,c} &=& \frac{y_{1,2}^2}{4} \frac{2 m_\psi^2 - m_S^2}{\Big(m_\psi^2 - (m_S+m_{u,c})^2 \Big)^2},\\
C_{T_2,{\rm VLP}}^{u,c} &=& y_{1,2}^2 \frac{m_S^2}{\Big(m_\psi^2 - (m_S+m_{u,c})^2 \Big)^2},
\end{eqnarray}
where we have explicitly kept the small quark masses $m_{u,c}$ in the denominators for the convenience of the following discussion.

To calculate DM-gluon Wilson coefficients $C_{S,{\rm VLP}}^{g}$ in Fig.\ref{fig-DD-diagram}, one needs to distinguish diagram (a) from (b,c), in which (a) belongs to long distance (LD) effect characterized by $q^{\rm loop}\sim m_{u,c,t}$ while diagram (b,c) belong to short distance (SD) effects characterized by $q^{\rm loop}\sim m_\psi$ \cite{Hisano2010-Gluon}. Only those effects coming from energy scales above $\mu_{\rm EFT}= m_Z$ should be integrated into $C_S^g$, which include all SD effects as well as top quark LD\footnote{Although the case $m_\psi<m_Z$ will slightly change the argument, a colored fermion particle lighter than $m_Z$ will be strongly constrained given the current LHC data. We do not consider such light colored mediator in this work.}.
\begin{eqnarray}
\label{eq-CSg-mZ}
C_{S,{\rm VLP}}^{g} &=& C_{S,{\rm VLP}}^{g}|_t^{\rm LD+SD} + C_{S,{\rm VLP}}^{g}|_{u,c}^{\rm SD}~,\\ \nonumber
C_{S,{\rm VLP}}^{g}|_t^{\rm LD+SD} &=& \frac{1}{4}\frac{y_3^2}{2} \Big(f^{(a)}_+ + f^{(b)}_+ + f^{(c)}_+\Big)(m_S;m_t,m_{\psi})~,\\ \nonumber
C_{S,{\rm VLP}}^{g}|_{u,c}^{\rm SD} &=&  \frac{1}{4}\frac{y_{1,2}^2}{2} \Big(f^{(b)}_+ + f^{(c)}_+\Big)(m_S;m_{u,c},m_{\psi})~,
\end{eqnarray}
where the expressions of $f^{(a,b,c)}_{+}(m_S;m_{u,c,t},m_\psi)$ can be found in \cite{Hisano2015-DMEFT}.

Furthermore, one needs to distinguish diagram (a) containing $u$ and $c$ quarks characterized by $q^{\rm loop}\sim m_{u,c}$. Tiny up quark mass $m_u<\Lambda_{\rm QCD}$ is in the non-perturbative QCD energy region, therefore diagram (a) for up quark should be included in the matrix element $\langle N |m_u \bar{u}u| N\rangle$ obtained from lattice QCD simulations, but not in the perturbative evaluation of $C_{S,{\rm VLP}}^{g}$.
On the contrary, $m_c>\Lambda_{\rm QCD}$ can be included in perturbative calculations and the LD diagram (a) for charm quark comes into place when one performs charm quark threshold matching near $m_c$, i.e. by integrating out charm quark into DM-gluon operator. More interestingly, an explicit calculation of $C_{S,{\rm VLP}}^{g}|_c^{\rm LD}$ from LD diagram (a) for charm quark at $\mu_{\rm EFT}= m_Z$ can be related to DM-quark Wilson coefficient $C_{S,{\rm VLP}}^{c}$ using the following relation \cite{Shifman1978-QQGG}
\begin{eqnarray}
m_Q \overline{Q}Q=-\frac{\alpha_s}{12\pi} G^{A\mu\nu}G^A_{\mu\nu},
\end{eqnarray}
where $Q$ denotes heavy quarks $c,b,t$. This implies
\begin{eqnarray}
\label{eq-WC-VLP-loop-matching}
C_{S,{\rm VLP}}^{c} =\Big(-\frac{1}{12} \Big)^{-1} \,C_{S,{\rm VLP}}^{g}|_c^{\rm LD},
\end{eqnarray}
which can be regarded as loop-diagram matching for $C_{S,{\rm VLP}}^{c}$. It has been shown that \cite{Gondolo2013-pole}, contrary to the pole at $m_\psi=m_S+m_c$ in Eq.(\ref{eq-WC-VLP-tree-matching}) obtained from tree-diagram matching using left-bottom diagram in Fig.\ref{fig-DD-diagram}, Eq.(\ref{eq-WC-VLP-loop-matching}) is regular at $m_\psi=m_S+m_c$. However, our numerical calculations do not cover this small mass parameter region $\sim m_c$. Therefore both matching results are applicable.

As for the Wilson coefficient $C_{T_2,{\rm VLP}}^{g}$ for the twist-2 gluon operator, it turns out to be of order $\mathcal{O}(y_i^2 \alpha_s^1)$ and thus suppressed by $\frac{\alpha_s}{\pi}$. As clarified in \cite{Hisano2010-Gluon,Hisano2015-DMEFT,Hisano2015-GluonWino} and also exhibited in \cite{Hill2014-I,Hill2014a-II}, for a chosen loop diagram generating DM-gluon interaction (which is up to 1-loop considered in this work), the $\frac{\alpha_s}{\pi}$ difference in the definitions of scalar-type gluon operator $\frac{\alpha_s}{\pi} G^{A\mu\nu}G^A_{\mu\nu}$ in Eq.(\ref{eq-twist-0}) and twist-2 gluon operator $G^{A\rho}_{\mu} G^{A}_{\nu\rho}-\frac{1}{4}g_{\mu\nu}G^A_{\rho\sigma}G^{A\rho\sigma}$ in Eq.(\ref{eq-twist-2}) results in an $\frac{\alpha_s}{\pi}$ suppression in $C_{T_2,{\rm VLP}}^{g}$ compared to $C_{S,{\rm VLP}}^{g}$. Considering that the matrix element of scalar and twist-2 gluon operators inside nucleon are at the same order $\mathcal{O}(1)$ \cite{Hisano2010-Gluon,Hill2014a-II}, the twist-2 gluon operator contribution is negligible and we ignore it in Eq.(\ref{eq-fN}) for the DM-nucleon effective coupling $f_N$.

\subsection{Differences between $\mu_{\rm EFT}=m_Z$ and $\mu_{\rm EFT}=\mu_{\rm had}$}
\label{section-difference-EFT-scale}

Before presenting the numerical results, we point out the main differences if one directly chooses $\mu_{\rm EFT}=\mu_{\rm had}\sim 1$ GeV instead of $\mu_{\rm EFT}=m_Z$ for the DM direct detection calculation.
\begin{itemize}
\item When choosing $\mu_{\rm EFT}=\mu_{\rm had}\sim 1$ GeV, calculation of DM-nucleon scattering rate is simplified and no RGE running and heavy quark threshold effects is applied to the Wilson coefficients.
\item Eq.(\ref{eq-fN}) for the DM-nucleon effective coupling $f_N$ only includes $C_{S,T_2}^{u,g}$ at $\mu_{\rm had}\sim 1$ GeV for this model, without the additional $d,s$ quark operators induced by RGE running and mixing in the $\mu_{\rm EFT}=m_Z$ case.
\item $u$-quark twist-2 operator matrix element $u(2;\mu)+\overline{u}(2;\mu)$ should take the value at $\mu=\mu_{\rm had}\sim 1$ GeV if one uses the Wilson coefficient $C_{T_2}^u(\mu_{\rm had})$ matched at $\mu_{\rm EFT}=\mu_{\rm had}$.
\item Both LD and SD effects of $c,t$ quark loops in Fig.\ref{fig-DD-diagram} are integrated into $C_{S,{\rm VLP}}^{g}$ as follows compared to Eq.(\ref{eq-CSg-mZ})
\begin{eqnarray}
\label{eq-CSg-1GeV}
C_{S,{\rm VLP}}^{g} &=& C_{S,{\rm VLP}}^{g}|_{c,t}^{\rm LD+SD} + C_{S,{\rm VLP}}^{g}|_{u}^{\rm SD}~,\\ \nonumber
C_{S,{\rm VLP}}^{g}|_{c,t}^{\rm LD+SD} &=& \frac{1}{4}\frac{y_{2,3}^2}{2} \Big(f^{(a)}_+ + f^{(b)}_+ + f^{(c)}_+\Big)(m_S;m_{c,t},m_{\psi})~,\\ \nonumber
C_{S,{\rm VLP}}^{g}|_{u}^{\rm SD} &=&  \frac{1}{4}\frac{y_{1}^2}{2} \Big(f^{(b)}_+ + f^{(c)}_+\Big)(m_S;m_{u},m_{\psi})~.
\end{eqnarray}
\end{itemize}

\subsection{Numerical results}
\label{section-DD-numerical}

After combining the Higgs portal in Section \ref{WC-HP} and vector-like $\psi$ portal in Section \ref{WC-VLP}, in Fig.\ref{fig_DDySQ-Psi} we present the constraints on a single $y_1,y_2,y_3$ when the other two are set to be zero on the plane of $m_S$ versus $y_i, (i=1,2,3)$, from the latest XENON1T results on spin independent DM-nucleon scattering rate. We choose three benchmark mediator masses $m_\psi = 100, 500, 1000$ GeV indicated by red, green and blue color, and use solid (dashed) lines to denote the choice of $\mu_{\rm EFT}=m_Z$ ( $\mu_{\rm had}$) with (without) RGE and heavy quark threshold effects. Since Eq.(\ref{eq-WC-HP-Case-I}) indicates that only top quark contributes sizably to $Fin.[\lambda_{SH}^{\rm1PI}]$ at $\mathcal{O}(y_3^2)$, we use blue and black lines to distinguish bounds on $y_3$ for the two benchmark choices of $\lambda_{SH}$ in Eq.(\ref{eq-WC-lamSH-Case-I}) and Eq.(\ref{eq-WC-lamSH-Case-II}).

\begin{figure}[h]
  \centering

  \includegraphics[width=8.5cm]{fig_DDySQ-Psi-1.pdf}
  \includegraphics[width=8.5cm]{fig_DDySQ-Psi-2.pdf}
  \includegraphics[width=8.5cm]{fig_DDySQ-Psi-3.pdf}
  \caption{The constraints on a single $y_1,y_2,y_3$ when the other two are set to be zero on the plane of $m_S$ versus $y_i$, from the latest XENON1T results on spin independent DM-nucleon scattering rate.}
  \label{fig_DDySQ-Psi}
\end{figure}

We have the following interesting observations:
\begin{itemize}
\item RGE and heavy quark threshold effects are important if one chooses $\mu_{\rm EFT}=m_Z$ while calculating the DM-nucleon scattering rate at $\mu_{\rm had}\sim 1$ GeV.
We regard the bounds on $y_1$ embedding the scale effects in Fig.\ref{fig-DD-diagram} as a complementarity to \cite{Giacchino:2015hvk} which performed a comprehensive study for scalar DM interactions with light quarks through a vector-like mediator, including a similar part $-y_1 S \overline{\psi_L} u_R$ in this work.
\item When focusing only on the vector-like $\psi$ portal (red, green, black), DM-up quark coupling $y_1$ generally receives the strongest constraints, followed by the weaker bounds on $y_2$ and $y_3$, except for a peak structure coming from the destructive cancellations between DM-gluon and DM-up quark interactions (observed also in \cite{Giacchino:2015hvk,Hill2014a-II}). Choices of $\mu_{\rm EFT}=m_Z$ and $\mu_{\rm EFT}=\mu_{\rm had}$ result in different peak locations.
\item Constraints on $y_3$ with $\lambda_{SH}^{\rm ren.}=Fin.[\lambda_{SH}^{\rm1PI}]$ (blue) is about $\mathcal{O}(10^1)$ stronger than the case of $\lambda_{SH}^{\rm ren.}=0$ (black), and can be even stronger than the bounds on $y_1$ with $m_\psi \gtrsim 200$ GeV. This implies that $y_3$ can contribute significantly through the Higgs portal $Fin.[\lambda_{SH}^{\rm1PI}]/m_h^2$ in Eq.(\ref{eq-lamSH-1PI}), but very little through the vector-like $\psi$ portal $C_{S,{\rm VLP}}^{g}|_{t}$ in Eq.(\ref{eq-CSg-mZ}). Note that strong constraints from DM direct detection experiments on the Higgs portal have already been well studied in, e.g. \cite{Cline2013,Beniwal2015-HP} and references therein.
\end{itemize}

\section{DM-heavy quark $c,t$ interactions}
\label{section-DM-heavy-quark}

Since both the $S-u$ coupling $y_1$ and the Higgs portal have been well discussed in the existing literature and strongly constrained, especially from the DM direct detection results, their existence would make it difficult to observe the effects of DM-heavy quark interactions $y_2,y_3$ through the vector-like $\psi$ portal. In the latter part of this work we will focus on the DM-heavy quark $c,t$ interactions by imposing
\begin{eqnarray}
\lambda_{SH}^{\rm ren.}=0,
\end{eqnarray}
and
\begin{eqnarray}
y_1=0.
\end{eqnarray}
In this case the new Lagrangian we consider is reduced to
\begin{eqnarray}
\mathcal{L} \supset -y_2 S \overline{\psi_L} c_R  -y_3 S \overline{\psi_L} t_R +  h.c..
\label{eq-L-tc}
\end{eqnarray}

\subsection{Relic Abundance}
\label{section-omg}
The most relevant DM annihilation processes for DM-$\{c,t\}$ interactions are given in Fig.\ref{fig-omgPlot}. The thermally averaged DM annihilation cross sections can be decomposed into partial waves as $\langle \sigma v\rangle\simeq s + p\,v^2$. We implement our model with {\bf FeynRules}
\cite{Alloul2014} and use {\bf micrOMEGAs} \cite{Belanger2015,Belanger2014} to calculate the DM thermal relic density which systematically takes into account co-annihilations, threshold and resonant effects \cite{Griest1991-omg}. Before presenting numerical results, here we give the analytic results obtained using {\bf FeynCalc} \cite{Mertig1991-FeynCalc,Shtabovenko2016-FeynCalc}. We found good agreement with the numerical calculations.

\begin{figure}[h]
  \centering

  \includegraphics[height=3cm,width=12cm]{omg.pdf}
  \caption{Most relevant DM annihilation channels in this work.}
  \label{fig-omgPlot}
\end{figure}

\begin{eqnarray}
\sigma v(SS\to t\bar{t})_s &=& \Big(y_3^4\Big) \frac{3}{4 \pi   m_t^2 }  \frac{(r_S^2-1)^{3/2}}{r_S^3 (r_S^2 (r_\psi^2+1)-1)^2	}~,\\ \nonumber
\sigma v(SS\to t\bar{t})_p &=& \Big(y_3^4\Big)  \frac{1}{32 \pi   m_t^2 }  \frac{(r_S^2-1)^{1/2}}{r_S^3 (r_S^2 (r_\psi^2+1)-1)^4}\\ \nonumber
&\quad&\times (-16 r_S^6 (2 r_\psi^2+1)+r_S^4 (r_\psi^2+1) (9 r_\psi^2+41)-2 r_S^2 (9 r_\psi^2+17)+9)~,
\end{eqnarray}
\begin{eqnarray}
\sigma v(SS\to t\bar{c}+c\bar{t})_s &=& \Big(y_2^2 y_3^2\Big) \frac{3}{16 \pi   m_t^2}  \frac{(1-4 r_S^2)^2 }{ r_S^4 (1-2 r_S^2 (r_\psi^2+1))^2}~,\\ \nonumber
\sigma v(SS\to t\bar{c}+c\bar{t})_p &=& \Big(y_2^2 y_3^2\Big)  \frac{1}{32 \pi   m_t^2 }  \frac{1-4 r_S^2}{r_S^4 (1-2 r_S^2 (r_\psi^2+1))^4}\\ \nonumber
&\quad&\times (64 r_S^6 (2 r_\psi^2+1)-4 r_S^4 (3 r_\psi^4+16 r_\psi^2+17)+2 r_S^2 (7   r_\psi^2+11)-3)~,
\end{eqnarray}
and
\begin{eqnarray}
\sigma v(S\psi\to gt)_s &=& \Big(g_s^2 y_3^2\Big) \frac{1}{24 \pi   m_t^2 } \frac{ r_S^2 (r_\psi+1)^2-1}{r_S^4 r_\psi^2 (r_\psi+1)^3}~,\\ \nonumber
\sigma v(S\psi\to gt)_p &=& \Big(g_s^2 y_3^2\Big) \frac{1}{144 \pi  m_t^2 }\frac{1}{r_S^4 r_\psi^2 (r_\psi+1)^5(r_S^2 (r_\psi+1)^2-1)}\\ \nonumber
&\quad&\times (r_S^4 \Big(13 r_\psi^2-7 r_\psi-4) (r_\psi+1)^4-2 r_S^2 (r_\psi^2-6 r_\psi-4) (r_\psi+1)^2~, \\ \nonumber
&\quad& \quad + 5 r_\psi^2-5 r_\psi-4\Big)
\end{eqnarray}
where $g_s$ is the SM $SU(3)$ gauge coupling and the mass ratios are defined by
\begin{eqnarray}
r_S = \frac{m_S}{m_{t}}, \quad r_\psi = \frac{m_\psi}{m_S}~.
\end{eqnarray}
Results for $SS\to c\bar{c}$ and $S\psi \to gc$ can be obtained by replacing $t\to c$ and $y_3 \to y_2$ in those of $SS\to t\bar{t}$ and $S\psi\to gt$.
Apart from the tree level annihilations, we also have $SS\to gg$ induced by the loop coupling $C_S^g$ and
\begin{eqnarray}
\sigma v(SS\to gg)_s &=& \frac{64}{\pi} |\frac{\alpha_s}{\pi}C_S^g|^2 m_S^2~,\\ \nonumber
\sigma v(SS\to gg)_p &=& \frac{16}{\pi} |\frac{\alpha_s}{\pi}C_S^g|^2 m_S^2~.
\end{eqnarray}

In the left panel of Fig.\ref{fig-Omg} we show the proper masses $m_S,m_\psi$ to produce $\Omega_{\rm DM} h^2\simeq 0.12$ \cite{PlanckCollaboration2015} (with an uncertainty of $10\%$ in numerical calculation) for $y_3=0.5$ which can provide a moderate contribution from top quark sector, and $y_2=0.5, 1, 3$ to cover the perturbative but sizable range of DM-charm interaction considering the chiral suppression from $m_c$. We also show the contributions from different annihilation channels for $y_2=y_3=0.5$ in the right panel. Since co-annihilation is only relevant when $(m_\psi/m_S-1)\lesssim 1/x_f$ with $x_f=m_S/T_f\sim 25$ \cite{Griest1991-omg}, the channels with dominant contribution are $SS\to  f\bar{f'}$ with $f,f'=t,c$, except when $m_S \gtrsim 1$ TeV in which case the co-annihilations $\overline{T}T\to {\rm SM}$ and $SS\to gg$ become important. Also note that for $m_S \gtrsim 2m_t$ when both the top and charm quark appear to be massless with respect to $m_S$, the contribution from $SS\to t\bar{c}+c\bar{t}$ approaches that of $SS\to t\bar{t}$ because of $y_2=y_3=0.5$ we choose for the illustration. Also, we remind ourselves that $C_S^g$ obtained in \cite{Hisano2010-Gluon,Hisano2015-DMEFT} assumes small gluon momenta expansion and thus may not sufficiently describe the annihilation process where  full loop calculations may be more suitable. However, the right panel of Fig.\ref{fig-Omg} indicates that tree level annihilations into $c,t$ quarks with $y_2,y_3\sim \mathcal{O}(1)$ can already produce the observed DM relic abundance. Therefore we leave the more complete exploration for the loop-induced DM annihilations in future works (see \cite{Colucci:2018vxz-1804} for a recent discussion and references therein).

\begin{figure}[h]
  \centering

  \includegraphics[height=6.5cm,width=7.5cm]{fig_Omg.pdf}
  \includegraphics[height=6.5cm,width=7.5cm]{fig_fc0505.pdf}
  \caption{{\bf Left}: Mass spectrum producing correct relic abundance ($10\%$ uncertainty included); {\bf Right}: Contributions from different annihilation channels for $y_2=y_3=0.5$.}
  \label{fig-Omg}
\end{figure}

\subsection{Direct detection}
\label{section-DD-heavy}

From the discussions in Section \ref{section-DD}, it is straightforward to obtain the constraints from XENON1T on the mass plane $m_S,m_\psi$ for the benchmark couplings $y_3=0.5$ and $y_2=0.5, 1, 3$ chosen for the thermal relic discussion. We show the contour bounds in Fig.\ref{fig-DD} for $\mu_{\rm EFT}=m_Z$ to be consistent with the phenomenology scales discussed in this section.

\begin{figure}[h]
  \centering

  \includegraphics[width=10cm]{fig_DD.pdf}
  \caption{Constraints from XENON1T results on the mass plane $m_S,m_\psi$ for $y_3=0.5$ and $y_2=0.5, 1, 3$ with $\mu_{\rm EFT}=m_Z$ to be consistent with the phenomenology scales discussed in this section. Regions below solid lines are excluded.}
  \label{fig-DD}
\end{figure}

\subsection{Indirect detection}
\label{section-ID}
Analogous to the method introduced in our previous work \cite{Baek2016a} (also in \cite{Bringmann:2012vr,Giacchino:2015hvk}), we impose the bounds set by Fermi observation of dwarf galaxy on DM annihilation into certain SM final states to further constrain our model \cite{Ackermann:2015zua}. In today's Universe with DM velocity $v\sim 10^{-3}c$, the relevant DM annihilations are the $s$-wave components of $SS\to  f\bar{f'}$ with $f,f'=t,c$ and $SS\to gg$. We construct the following function of the number of photons produced in a single DM pair annihilation event:
\begin{eqnarray}
N_\gamma^{th} &=&2N_{\gamma,t}(m_S) \times \sigma v(SS\to t\bar{t})_s + 2N_{\gamma,c}(m_S) \times \sigma v(SS\to c\bar{c})_s \\ \nonumber
&\quad& + \Big(N_{\gamma,t}(E_t)+N_{\gamma,c}(E_c)\Big) \times \sigma v(SS\to t\bar{c}+c\bar{t})_s + 2N_{\gamma,g}(m_S) \times \sigma v(SS\to gg)_s~, \\ \nonumber
N_\gamma^{exp} &=& 2N_{\gamma,b}(m_S) \times \sigma v_{b\bar{b}}(m_S),
\end{eqnarray}
where $N_{\gamma,f}(E_f)$ denotes the number of photons produced by a single SM fermion $f$ with energy $E_f$
\begin{eqnarray}
N_{\gamma,f}(E_f)\,= \, \int^{E_{max}}_{E_{min}}  dE_\gamma \frac{dN_{\gamma,f}}{dE_\gamma} (E_f)~,
\end{eqnarray}
and
\begin{eqnarray}
E_t =  \frac{4m_S^2+m_t^2-m_c^2}{4m_S}~, \quad E_c =  \frac{4m_S^2+m_c^2-m_t^2}{4m_S} ~.
\end{eqnarray}
We choose the following photon energy range according to Fermi's report
\begin{eqnarray}
E_{max} \, = \, min\{E_f, 300 \, {\rm GeV}\}~, \quad \quad E_{min} \, = \, 0.3 \, {\rm GeV}.
\end{eqnarray}
We use the photon number spectra $\frac{dN_{\gamma,f}}{dE_\gamma} (E_f)$ calculated by {\bf PPPC4DM} \cite{Buch2015-PPPC4DM,Cirelli2010-PPPC4DM} which contain the electroweak corrections. $\sigma v_{b\bar{b}}(m_S)$ is also taken from Fermi's report. We set the following requirement
\begin{eqnarray}
N_\gamma^{th} \lesssim N_\gamma^{exp}.
\end{eqnarray}

In the left panel of Fig.\ref{fig-ID} we show $N_{\gamma,f}(E_f)$ for $f=t,c,b,g$ taken from {\bf PPPC4DM}, from which one can see that the top quark is capable of producing more photons than others, meanwhile photons from charm quark are approximately the same as the those from gluon. This observation suggests that the constraints from Fermi dwarf observation should be stronger when on-shell top quark(s) is produced.
In the right panel we show the bounds on the plane of $m_\psi, m_S$ for $y_3=0.5$ and $y_2=0.5, 1, 3$. As expected, the exclusion becomes stronger when $m_S\gtrsim m_t/2$ in which case a single top quark can be produced. The bounds are even more stringent when $SS\to t\bar{t}$ is kinematically open and the lower bound of the mediator mass can reach $\{400,600,1000\}$ GeV, respectively. We also notice that Fermi dwarf observation is a good complementarity in $m_S>m_t/2$ to the direct detection constraints from XENON1T, which is significant for $m_S<m_t/2$ as shown in Fig.\ref{fig-DD}.

\begin{figure}[h]
  \centering

  \includegraphics[height=6.5cm,width=7.5cm]{fig_ID_Nphoton.pdf}
  \includegraphics[height=6.5cm,width=7.5cm]{fig_ID.pdf}
  \caption{{\bf Left}: Photon numbers produced by different SM fermions taken from {\bf PPPC4DM} \cite{Buch2015-PPPC4DM,Cirelli2010-PPPC4DM}; {\bf Right}: Constraints from current Fermi dwarf observation \cite{Ackermann:2015zua}, regions below solid lines are excluded.}
  \label{fig-ID}
\end{figure}

\subsection{FCNC of top quark}
\label{section-FCNC}
The FCNC processes for top quark in DM-$\{c,t\}$ interactions can be generated at both tree level $t\to T^{(*)}S \to cSS$ and loop level $t\to c+\gamma/g/Z$, with the diagrams shown in Fig.\ref{fig-FCNCPlot}.

\begin{figure}[h]
  \centering

  \includegraphics[height=3cm,width=12cm]{FCNC.pdf}
  \caption{FCNC processes of top quark in this model.}
  \label{fig-FCNCPlot}
\end{figure}

The widths of tree-level processes can be expressed by
\begin{eqnarray}
\Gamma(t\to cSS) &=&  \Big(y_2^2 y_3^2\Big) \frac{1}{1024 \pi^3 m_t} \int^{m^2_{12,max}}_{m^2_{12,min}} dm^2_{12} \int^{m^2_{23,max}}_{m^2_{23,min}} dm^2_{23} \frac{1}{(m_{12}^2-m_\psi^2)^2(m_{23}^2-m_\psi^2)^2}   \\ \nonumber
&\quad& \times \Big( m_{12}^6 (-m_{23}^2)+2 m_{12}^4 m_{23}^4+m_{12}^4 m_{23}^2 m_c^2-m_{12}^2 m_{23}^6+m_{12}^2 m_{23}^4 m_c^2-4 m_{12}^2 m_{23}^2\\ \nonumber
&\quad&\quad  m_c^2 m_\psi^2+m_{12}^2 m_c^2 m_\psi^4+m_t^2 (m_{12}^4 (m_{23}^2+m_c^2-m_S^2)+m_{12}^2 (m_{23}^4+2 m_c^2 (m_{23}^2-2 m_\psi^2) \\ \nonumber
&\quad&\quad  -4 m_{23}^2 m_\psi^2+2 m_S^2 m_\psi^2+m_\psi^4)+m_{23}^4
   m_c^2-4 m_{23}^2 m_c^2 m_\psi^2 -m_S^2
   (m_{23}^2-m_\psi^2)^2\\ \nonumber
&\quad&\quad +m_{23}^2 m_\psi^4+4 m_c^2
   m_\psi^4-m_S^2 m_\psi^4)+m_S^2 (m_{12}^4
   (m_S^2-m_c^2)+m_{12}^2 (2 m_c^2 m_\psi^2-2
   m_{23}^2 m_S^2)\\ \nonumber
&\quad&\quad  +m_{23}^4 m_S^2-m_c^2
   m_\psi^4)  -m_{23}^4 m_c^2 m_S^2+2 m_{23}^2 m_c^2
   m_S^2 m_\psi^2+m_{23}^2 m_c^2 m_\psi^4-m_c^2 m_S^2
   m_\psi^4 \Big)~,
\end{eqnarray}
where we have kept $m_c$ in the calculation and
\begin{align}
m^2_{12,min} &= (m_S+m_c)^2, \quad m^2_{12,max} = (m_t-m_S)^2~, \\ \nonumber
m^2_{23,min} &= (E^*_2+E^*_3)^2-\left(\sqrt{E^{*2}_2-m_c^2} + \sqrt{E^{*2}_3-m^2_S} \right)^2~, \\ \nonumber
m^2_{23,max} &= (E^*_2+E^*_3)^2-\left(\sqrt{E^{*2}_2-m_c^2} - \sqrt{E^{*2}_3-m^2_S} \right)^2~, \\ \nonumber
E^*_2 &=\frac{m_{12}^2-m_S^2+m_c^2}{2m_{12}}, \quad E^*_3 =\frac{m_t^2-m_{12}^2-m_S^2}{2m_{12}}~.
\end{align}

For the loop level processes $t\to c+\gamma/g/Z$, the generic amplitudes can be written in the following form:
\begin{equation}
i {\mathcal M}_{tcV}=\bar u(p_2)\, \Gamma^\mu  u(p_1)\, \epsilon_\mu(k,\lambda)~,
\label{eq:deltaM}
\end{equation}
where $V=\gamma/g/Z$ and $p_1, p_2$, and $k$ are the 4-momenta of the incoming top quark,
outgoing charm quark and outgoing gauge boson, respectively.
$\epsilon_\mu(k,\lambda)$ is the polarization vector of the outgoing gauge
boson. The vertices $\Gamma^\mu$ with on-shell external particles can be decomposed as \cite{Lopez1997-FCNC}
\begin{eqnarray}
\Gamma^\mu_{tcZ}&=&\gamma^\mu (P_L A_{Z1}+ P_R B_{Z1})
+ i\sigma^{\mu\nu}k_\nu (P_L A_{Z2} + P_R B_{Z2})\, ,
\label{eq:VtcZ}\\
\Gamma^\mu_{tc\gamma}&=&i\sigma^{\mu\nu}k_\nu (P_L A_{\gamma2} + P_R B_{\gamma2})\, ,
\label{eq:Vtcgamma}\\
\Gamma^\mu_{tcg}&=& T^a i\sigma^{\mu\nu}k_\nu (P_L A_{g2} + P_R B_{g2})\,,
\label{eq:Vtcg}
\end{eqnarray}
where $P_{R,L}={1\over2}(1\pm\gamma_5)$, $\sigma^{\mu\nu}={i\over 2}[\gamma^\mu,\gamma^\nu]$ and $T^a$ are the generators
of $SU(3)$ color. Then the widths are
\begin{eqnarray}
\Gamma(t\rightarrow cZ)&=&{m_t^3 \over 32 \pi m_Z^2}(1- {m_Z^2\over m_t^2})^2
\Bigl[(1+2{m_Z^2\over m_t^2})(|A_{Z1}|^2+|B_{Z1}|^2) \nonumber\\
&&-6{m_Z^2 \over m_t}\,Re\Big (A_{Z1}B_{Z2}^*+
A_{Z2} B_{Z1}^*\Big) +m_Z^2(2+
{m_Z^2 \over m_t^2})(|A_{Z2}|^2+|B_{Z2}|^2)\Bigr]\, ,\\
\Gamma(t\rightarrow c\gamma)&=&{m_t^3\over 16 \pi } (|A_{\gamma2}|^2+|B_{\gamma2}|^2)\, ,\\
\Gamma(t\rightarrow cg)&=&C_F\,{m_t^3\over 16 \pi } (|A_{g2}|^2+|B_{g2}|^2)\, ,
\end{eqnarray}
where we have used $m_c=0$ for simplification. After calculating the loop diagram in Fig.\ref{fig-FCNCPlot} and extracting the coefficients \cite{Soares1989,Deshpande1982} we can obtain
\begin{align}
A_{Z1}&=-D_Z\, m_c m_t \,\Big( g_{ZL}\frac{B_t-B_c}{m_t^2-m_c^2} + g_{ZR} (C_{0} + 2 C_{1} + C_{11} + 2C_{12} + 2 C_{2} + C_{22}) \Big)~, \nonumber\\
B_{Z1}&=-D_Z\, g_{ZR} \quad \Big( \frac{m_t^2 \,B_t- m_c^2\, B_c}{m_t^2 - m_c^2} + m_\psi^2 C_0 - 2 C_{00} + m_Z^2 C_{12} + \frac{1}{2}\Big)~, \nonumber\\
A_{Z2}&=D_Z\, g_{ZR}\,  m_t \, (C_{12}+C_2+C_{22}) ,\quad\quad B_{Z2}= D_Z\, g_{ZR}\,  m_c \, (C_{1}+C_{11}+C_{12})~, \nonumber\\
A_{\gamma2}&= D_\gamma\, g_{\gamma R}\,  m_t\,  (C_{12}+C_2+C_{22}),\quad\quad \, \,B_{\gamma2}= D_\gamma\, g_{\gamma R}\,  m_c \, (C_{1}+C_{11}+C_{12}) ~, \nonumber\\
A_{g2}&=D_g\, g_{gR}\,  m_t \, (C_{12}+C_2+C_{22}) ,\quad\quad \,\,\, B_{g2}=D_g\, g_{gR}\,  m_c \, (C_{12}+C_2+C_{22}) ~,
\end{align}
where
\begin{align}
B_t &= (B_0+B_1)\Big(m_t^2;m_S^2,m_\psi^2\Big)~, \nonumber\\
B_c &= (B_0+B_1)\Big(m_c^2;m_S^2,m_\psi^2\Big)~, \nonumber\\
C_{i,ij}\, &=C_{i,ij}\Big(m_c^2, m_V^2, m_t^2; m_S^2, m_\psi^2, m_\psi^2\Big)~, \nonumber\\
D_Z\, &=-y_2 y_3\,  \frac{g_2}{c_w}\, \frac{i}{16\pi^2}, \quad D_\gamma\, =-y_2 y_3\,  e\,  \frac{i}{16\pi^2}, \quad D_g\, =-y_2 y_3\,  g_s\,  \frac{i}{16\pi^2}~, \nonumber\\
g_{ZL} &=\frac{1}{2}-\frac{2}{3}s_w^2, \quad\quad\quad~~g_{ZR} =-\frac{2}{3}s_w^2, \quad\quad\quad~~~ g_{\gamma R} =g_{gR}=\frac{2}{3},
\end{align}
with $s_w\equiv \sin\theta_w,c_w\equiv \cos\theta_w$ and $\theta_w$ is the Weinberg angle. We checked the loop divergence cancellation in the widths and use {\bf LoopTools} \cite{Hahn1999-LoopTools} based on {\bf FF} package \cite{Hahn1999-FF} to perform numerical calculations.

In Fig.\ref{fig-FCNC} we choose $y_2=y_3=0.5$ as a benchmark point and show the predictions of $Br(t\to cSS)$ and $Br(t\to c+\gamma/g/Z)$ in the left and right panel, respectively. We found that $Br(t\to cSS)$ can exceed $10^{-7}$ with a mediator lighter than 1 TeV. The loop FCNC branching fractions can reach $10^{-8}$ with light mediator $m_\psi\lesssim 300$ GeV but are smaller than $10^{-10}$ with $m_\psi \sim 1$ TeV. Benefiting from the strong coupling strength and color index, $Br(t\to cg)$ is about one order of magnitude larger than the other two.

\begin{figure}[h]
  \centering

  \includegraphics[height=6.5cm,width=7.5cm]{fig_FCNC_tcSSnew1.pdf}
  \includegraphics[height=6.5cm,width=7.5cm]{fig_FCNC_smallscalenew.pdf}
  \caption{FCNC of top quark at tree ({\bf left} panel) and loop level ({\bf right} panel).}
  \label{fig-FCNC}
\end{figure}

\subsection{Collider detection}
\label{section-collider}

The typical collider signals for DM-$\{c,t\}$ interactions are $\slashed{E}_T+\{t\bar{t},jj,tj\}$ from the pair production of colored mediator $pp \to \psi\overline{\psi}$ followed by $\psi\to S+t/c$. In this work we consider the $t\bar{t},jj$ signals which are analogous to the supersymmetry (SUSY) searches for stop and first two generation squarks $pp\to \tilde{t}^*\tilde{t}, \tilde{q}^*\tilde{q}$, but with larger $\psi\overline{\psi}$ production cross section compared to $\tilde{t}^*\tilde{t}, \tilde{q}^*\tilde{q}$ with the same mass due to the internal spin degree of freedom. In our previous work \cite{Baek2016a} we performed a detailed study of $\slashed{E}_T+t\bar{t}$ signal for top-philic DM scenario based on ATLAS 13.2 $fb^{-1}$ data at 13 TeV and found that the excluded mediator mass can reach 1150 GeV compared to 800 GeV in stop search \cite{ATLAS-CONF-2016-050}. In the SUSY case, the bounds on first two generation squarks are stronger than stop, which have currently reached 1.6 TeV \cite{ATLAS-CONF-2017-022} for squarks with $36\, fb^{-1}$ data at LHC compared to 950 GeV for stop \cite{ATLAS-CONF-2017-037} using $1\ell + jets + \slashed{E}_T$ signal. As a result, $\slashed{E}_T+jj$ should also be an important signal at collider when we allow DM to couple to both charm and top quark in this work.

The production cross sections of $\slashed{E}_T+t\bar{t},jj$ depend on $Br(\psi\to St,Sc)$. The partial decay width of the mediator is
\begin{eqnarray}
\Gamma(\psi \to SQ)&=& y_Q^2 \frac{m_\psi}{32\pi} (1+x_Q^2-x_S^2)\Big[\Big(1-(x_Q+x_S)^2\Big)\Big(1-(x_Q-x_S)^2\Big)\Big]^{1/2}\\ \nonumber
 &\equiv& y_Q^2 F(m_Q,m_S,m_\psi)~,
\end{eqnarray}
where $Q=c,t$ are on-shell, $y_Q=y_2,y_3$ and $x_Q=m_Q/m_\psi, x_S=m_S/m_\psi$. Since decaying into top quark contains larger phase space suppression compared to the charm quark case, $F(m_t,m_S,m_\psi)$ will be smaller than $F(m_c,m_S,m_\psi)$ except for $\Delta m=m_\psi-m_S\gg m_t$ in which case $F(m_t,m_S,m_\psi)\simeq F(m_c,m_S,m_\psi)$. Then the branching fractions are
\begin{eqnarray}
\label{Eq-BrPsi}
Br(\psi \to St)&=& \frac{1}{1+\frac{y_2}{y_3}\frac{F(m_c,m_S,m_\psi)}{F(m_t,m_S,m_\psi)}}~, \\ \nonumber
Br(\psi \to Sc)&=& \frac{1}{1+\Big(\frac{y_2}{y_3}\frac{F(m_c,m_S,m_\psi)}{F(m_t,m_S,m_\psi)}\Big)^{-1}}~,
\end{eqnarray}
and the signal production cross sections are
\begin{eqnarray}
\label{Xsection}
\sigma\Big(pp \to \psi\overline{\psi} \to \slashed{E}_T+t\bar{t} \Big) = \sigma\Big(pp \to \psi\overline{\psi}\Big) \, Br^2(\psi \to St)~, \\ \nonumber
\sigma\Big(pp \to \psi\overline{\psi} \to \slashed{E}_T+jj \Big) = \sigma\Big(pp \to \psi\overline{\psi}\Big) \, Br^2(\psi \to Sc)~.
\end{eqnarray}

In Fig.\ref{fig-BrXs} we set $y_2=y_3$ to show $Br(\psi\to Sc)$ on the plane of $m_S, m_\psi$ with $\Delta m>m_t$. One can easily use $1-Br(\psi\to Sc)$ to obtain the corresponding plot for $Br(\psi\to St)$. We can see that $\Delta m\sim m_t$ makes $Br(\psi\to Sc)$ easily exceed $90\%$, while $\Delta m\gtrsim 400$ GeV would make $F(m_t,m_S,m_\psi)\simeq F(m_c,m_S,m_\psi)$ and result in $Br(\psi\to St)\simeq Br(\psi\to Sc)\simeq 0.5$. $\sigma(pp \to \psi\overline{\psi})$ at $\sqrt{s}=13$ TeV can be found in \cite{ATLASCollaboration2017-VLQuark-2} which uses Top++ v2.0 \cite{Czakon:2011xx} and the MSTW 2008 NNLO PDF set. Next-to-next-to-leading-order (NNLO) QCD corrections and the soft gluon resummation to NNLL accuracy \cite{Cacciari:2011hy,Beneke:2011mq,PhysRevLett.109.132001,Czakon:2012zr,Czakon:2012pz,Czakon:2013goa} are also included.

\begin{figure}[h]
  \centering

  \includegraphics[height=6cm,width=7cm]{fig_BrTSc.pdf}
  \caption{$Br(\psi\to Sc)$ with $y_2=y_3$ on the plane of $m_S, m_\psi$ with $m_\psi-m_S>m_t$.}
  \label{fig-BrXs}
\end{figure}

We use FeynRules \cite{Alloul2014} to implement the model in this work to obtain UFO model file
\cite{Degrande:2011ua-UFO}, which is utilized by MadGraph5 \cite{Alwall2014} to generate the parton level events which
is further linked to PYTHIA8 \cite{Sjostrand:2007gs-PYTHIA81,Sjostrand:2014zea-PYTHIA82} to perform the parton shower
and hadronization. Then the PYTHIA event files are processed by CheckMate2 \cite{Kim2015,Drees2013} which utilizes
Delphes3 \cite{Ovyn2009,DeFavereau2014} to simulate the detector response and calculate the cut efficiency $\epsilon$. The number of signal events after selection cuts are calculated as $N_{sig}=\mathcal{L}*\sigma*\epsilon$ where $\mathcal{L}=36 \,{\rm fb^{-1}}$ is the ATLAS integrated luminosity at 13 TeV in \cite{ATLAS-CONF-2017-022,ATLAS-CONF-2017-037} and $\sigma$ is the $pp\to \psi\overline{\psi}$ production cross section at $\sqrt{s} = 13 \,{\rm TeV}$. We use top++2.0 \cite{Czakon:2011xx} to obtain $\sigma(pp\to \psi\overline{\psi})$ at next-to-next-to-leading order (NNLO) which includes the next-to-next-to-leading logarithmic (NNLL) contributions. We vary the renormalization and factorization scale between $(0.5, 2)m_\psi$ to estimate theoretical uncertainty $\Delta \sigma$ at $1\sigma$ level. CheckMate uses $\Delta \sigma$ and the number of simulated events $N_{MC}$ to calculate $\Delta N_{sig}$, which is the total uncertainty of signal event number. Finally, CheckMATE defines the following quantity:
\begin{eqnarray}
r_{CM} \equiv \frac{N_{sig} - 1.96 \Delta N_{sig}}{N^{95}_{obs}}
\end{eqnarray}
where $N_{obs}^{95}$ is the model independent bounds at $95\%$ Confidence Level (C. L.) on the number of new physics signal events provided in the experimental reports. Then a model can be claimed to be excluded at the $95\%$ C. L. if $r_{CM}>1$. Note that $r_{CM}$-limit is usually weaker than the method based on $S/\sqrt{S + B} < 1.96$ since $r_{CM}$-limit utilizes the total uncertainty on the $N_{sig}$ in a more conservative manner. More detail are provided in \cite{Kim2015,Drees2013}.

In fig.\ref{fig-Collider} we show the collider bounds from $jets + \slashed{E}_T$ searches in \cite{ATLAS-CONF-2017-022}
(left panels) and from $1\ell + jets + \slashed{E}_T$ searches in \cite{ATLAS-CONF-2017-037} (right panels), where
different rows correspond to $y_2=0.5,1,3$ with common $y_3=0.5$, respectively. To account for uncertainties in our
simulation, we highlight $r_{CM}=1\pm 20\%$ samples near the exclusion criteria $r_{CM}=1$. Light grey, red, green,
black samples correspond to $r_{CM}$ values within $r_{CM}<0.8$, $[0.8,1)$, $[1,1.2)$ and $r_{CM}>1.2$,
respectively. The left panels show that with increasing $y_2$ the branching fraction $Br(\psi \to Sc)$ get enhanced and
results in generally larger $jets + \slashed{E}_T$ production cross sections and wider exclusion
region. Correspondingly, the smaller $t\bar{t} + \slashed{E}_T$ production cross sections result in lower exclusion
sensitivity. In the first row with $y_3=y_2=0.5$, we found that $m_\psi$ can be excluded up to around 1100 (950) GeV
using $jets + \slashed{E}_T$ ($1\ell + jets + \slashed{E}_T$) signal, and the sensitivity can reach $m_S$ around 200
(400) GeV. On the contrary, in the lowest row with $y_2=3.0$ and dominating $Br(\psi \to Sc)$, the $jets +
\slashed{E}_T$ sensitivity is greatly amplified and $m_\psi$ can be excluded up to 1400 GeV while the $1\ell + jets +
\slashed{E}_T$ search completely loses the sensitivity.

\begin{figure}[h]
  \centering

  \includegraphics[height=6cm,width=7cm]{fig_jets_y3_05_y2_05.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_lep_y3_05_y2_05.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_jets_y3_05_y2_10.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_lep_y3_05_y2_10.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_jets_y3_05_y2_30.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_lep_y3_05_y2_30.pdf}
  \caption{ATLAS bounds on the model of this work using $36\, fb^{-1}$ data at 13 TeV. {\bf Left}: $jets + \slashed{E}_T$ signal; {\bf Right}: $1 \ell + jets + \slashed{E}_T$ signal. Rows from top to bottom correspond to $y_2=0.5,1,3$ with common $y_3=0.5$. All masses are in unit of GeV.}
  \label{fig-Collider}
\end{figure}

\subsection{Combined results}
\label{section-combined}

In Fig.\ref{fig-combine} we show the combined results from the DM relic abundance requirement $\Omega
h^2\simeq 0.12$, the DM direct/indirect detection constraints, 13 TeV LHC data and the top FCNC predictions for $y_3=0.5$ and $y_2=0.5,1,3$. On the left panels, the pink, cyan, grey and yellow regions are excluded by the current XENON1T,  Fermi dwarf, LHC searches for $jets + \slashed{E}_T$ and $1\ell + jets + \slashed{E}_T$ at 13 TeV, while the black solid lines correspond to the correct DM relic abundance. On the right panels, the solid lines with red, green, blue and orange colors are predictions of $Br(t\to c+\gamma/g/Z/SS)$ when $\Omega_{\rm DM} h^2\simeq 0.12$ is satisfied, respectively. Rows from top to bottom correspond to $y_2=0.5,1,3$ with common $y_3=0.5$.
We found that with increasing $y_2$, the combined results can exclude the whole light DM mass region $m_S\lesssim100$ GeV when scalar $S$ forms the whole DM components, and the top FCNC branching fractions are usually below $10^{-9}$. Note that the current precision of top quark width measurement is around $10\%$ level \cite{CMSCollaboration2014,ATLAS-CONF-2017-056} and the FCNC measurement is around $10^{-4}$ \cite{Bhowmik:2017gzr-FCNCtop}, the benchmark scenarios shown in Fig.\ref{fig-combine} are still safe from top quark width measurements when passing other constraints.
Note that both the experimental measurements ($\sim 10^{-4}$) and the theoretical predictions with $y_2=y_1$ for $Br(t \to u/c+V)$ are close to each other because of $m_u\sim m_c\sim 0$, meanwhile Fig.\ref{fig_DDySQ-Psi} indicates that $y_1$ generally receives stronger limits than $y_2$ from the direct detection (except for a narrow region with destructive cancellation). Therefore, passing the DM direct detection constraints can also make it easy to pass the top FCNC bounds for DM interactions with light quarks through $y_1$ as long as $y_3\sim \mathcal{O}(1)$.

\begin{figure}[h]
  \centering

  \includegraphics[height=6cm,width=7cm]{fig_Collider1.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_BrTopExotic_1.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_Collider2.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_BrTopExotic_2.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_Collider3.pdf}
  \includegraphics[height=6cm,width=7cm]{fig_BrTopExotic_3.pdf}
  \caption{Combined results. {\bf Left: } mass relations required by observed relic abundance confronting the excluded region by direct/indirect detection and 13 TeV LHC data; {\bf Right: } predicted top FCNC branching fractions when satisfying $\Omega_{\rm DM} h^2\simeq 0.12$. Rows from top to bottom correspond to $y_2=0.5,1,3$ with common $y_3=0.5$, respectively.}
  \label{fig-combine}
\end{figure}

\section{Conclusion}
\label{section-conclusion}

The absence of confirmed signal in dark matter direct detection may suggest weak interaction strengths between DM and the abundant constituents inside nucleon, i.e. gluons and valence light quarks. In this work we consider a real scalar dark matter $S$ interacting only with $SU(2)_L$ singlet Up-type quarks $U_i=u_R,c_R,t_R$ via a vector-like fermion $\psi$ which has the same quantum number as $U_i$. The DM-nucleon scattering can proceed through both $h$-mediated Higgs portal (HP) and $\psi$-mediated vector-like portal (VLP), in which HP can receive sizable radiative corrections through the new fermions.

We first study the separate constraints on the new Yukawa couplings $y_i$ and find that the constraints of XENON1T results are strong on $y_1$ from VLP scattering and on $y_3$ from its radiative contributions to HP scattering. Since both DM-light quark interactions and HP have been well studied in the existing literature, we move forward to focus on DM-heavy quark interactions. Since there is no valence $c,t$ quark inside nucleons at $\mu_{\rm had}\sim 1$ GeV, $y_2,y_3$ interactions are manifested in DM-gluon scattering at loop level.
We find that renormalization group equation and heavy quark threshold effects are important if one calculates the DM-nucleon scattering rate $\sigma^{\rm SI}_{p}$ at $\mu_{\rm had}\sim 1\, {\rm GeV}$ while constructing the effective theory at $\mu_{\rm EFT}\sim m_Z$. For the benchmarks $y_3=0.5, y_2=0.5, 1, 3$, combined results from $\Omega_{\rm DM} h^2\simeq 0.12$, XENON1T, Fermi-LAT, 13 TeV LHC data have almost excluded $m_S<m_t/2$ when only DM-$\{c,t\}$ interactions are considered. FCNC of top quark can be generated at both tree level $t\to \psi^{(*)}S \to cSS$ and loop level $t\to c+\gamma/g/Z$, of which the branching fractions are typically below $10^{-9}$ after passing the other constraints, which are still safe from the current top quark width measurements.

\section*{Acknowledgement}
Peiwen Wu would like to thank Junji Hisano, Natsumi Nagata, Kei Yagyu, Junjie Cao, Yi-Lei Tang, Yangle He, Liangliang
Shang for helpful discussions. We thank the Korea Institute for Advanced Study for providing computing resources (KIAS
Center for Advanced Computation Abacus System) for this work. This research was partly supported by the Munich Institute
for Astro- and Particle Physics (MIAPP) of the DFG cluster of excellence "Origin and Structure of the Universe".
This work is also supported in part by National Research Foundation of Korea (NRF) Research Grant
NRF-2015R1A2A1A05001869 (SB, PK).

\bibliographystyle{JHEP}

redpajama_set_name: RedPajamaArXiv


text:
Danny Katch
Duncan's twisted vision for our schools
Arne Duncan is showing more backbone in taking on teachers and kids than the Obama administration did with insurance companies, says Danny Katch.
HURRICANE KATRINA was "the best thing that happened to the education system in New Orleans," according to Arne Duncan, who unfortunately is not a diabolical Hollywood villain but the U.S. Secretary of Education.
Four years after the hurricane, there are barely half as many children in New Orleans public schools. That's 30,000 mostly African American boys and girls who have been driven from their homes, possibly permanently.
Besides this easily being the Cheneyist thing anybody in the Obama administration has said, Duncan's words are so striking because he claims he thinks of his education policies as "the civil rights issue of our generation."
But then, Duncan has always been a strange type of civil rights leader. On the day he claimed to speak out for this generation's civil rights issue, he was the "CEO" of the Chicago public schools, which is the fourth most segregated school system in the country, according to a study sponsored by the Center for Urban Research and Learning at Loyola University.
Education Secretary Arne Duncan
It's not just Chicago. All around the country, schools are about as racially integrated today as they were before Brown v. the Board of Education. Schools named after Martin Luther King are often exclusively non-white. This is like having a George Washington High with Redcoats for teachers.
There are no rewards in Duncan's "Race to the Top" program for states that desegregate their schools. To Duncan, however, segregation is so "last generation." If he were to give a Martin Luther King-style civil rights speech, it might go like this:
I have a dream that one day this nation will rise up and give an equal opportunity to some of its children.
I have a dream that from the streets of Harlem to the hills of Georgia, children will be judged not by the color of their skin but by their score on a test...which is based on the school they attend...which is based on the color of their skin.
I have a dream that our nation will have enough educated workers to meet our needs without having to actually educate everybody.
And finally, I have a dream that our most disadvantaged children--the little girl coming to school from a homeless shelter, the boy from Mexico just learning our language, the children struggling with autism--I have a dream that these children will just go away because they are very poor investments.
THE NEW York Times reports that Duncan is planning "sweeping change" to the notorious No Child Left Behind law pushed through by George W. Bush.
Like many things in the age of Obama, this sounds great until you keep reading. "Right now, most federal money goes out in formulas, so schools know how much they'll get, and then use it to provide services for poor children," Jack Jennings, president of the Center on Education Policy, who attended a recent meeting with administration officials, told the Times. "The department thinks that's become too much of an entitlement."
Amen. If you've ever spent time in a school with poor children, the first thing that strikes you is their sense of entitlement: "I don't want to climb five flights of stairs to get to my 7th period class," they whine about the constantly broken elevator. "I have asthma."
As Jennings continued: "[Obama and Duncan] want to upend that scheme by making states and districts pledge to take actions the administration considers reform, before they get the money."
Darn right, kids! You don't get any money from this administration before you change your failing ways. Just ask AIG. Or the Pentagon. Or Joe Biden.
But we all know that children aren't the real villains. Teachers are. The thing that's really insidious about teachers is they're so committed to ruining education that, unlike many school administrators, they devote their whole lives to the field. Well, no more. The heart of the Duncan's plan is to "evaluate teachers based on student test data."
Using test scores to pay and fire teachers is such a good idea that this method should be applied to all professions. Firefighters should be judged according to the percentage of the fire that they put out--tough luck if you were driving the truck. Doctors should be paid based on survival rates, so when EMTs rush a multiple gunshot wound victim into the emergency groom, they'll find everyone out for a coffee break.
The other main element of the new policy is charter schools--lots of them. Teachers don't like charter schools in the same stubborn, unreasoning way that 11-year-olds don't like sweatshops.
But charters have proven teachers wrong on one thing. At least some teachers in their darkest moments have looked at their students and thought that if they could just get rid of that one kid--usually the one running a dice game in the corner of the classroom--the entire class would be better off. But charter schools have shown that you can exclude all those kids, as well as the ones with special learning needs or parents who can't spend 20 hours a week at the school, and still not show improvement over regular schools.
Hopefully, Duncan's education "reforms" will go the way of the health care bill. But don't count on it. The Obama administration might show more backbone taking on teachers and kids than it did with the insurance companies. After all, this is a question of civil rights.
Jamie Partridge
A Portland school fights for its activist teacher
Protests by students and parents stopped the wrongful suspension of a Portland middle school teacher and critic of the district.
NYC students get metal detectors expelled
A December walkout at New York City high schools was the culmination of a successful student-led fightback against police searches.
Gabriel Paez
Protesting another TIF theft in Chicago
New hot dog stands and a taller Ferris wheel at Navy Pier are more important to Chicago Mayor Rahm Emanuel than our schools.
Peter Lamphere
CPE1's principal has to go
Parents have banded together with teachers at a public elementary school in New York City to demand a change in leadership.
Mike Shea
Will school end early in Chicago?
Chicago school officials blame Illinois' Republican governor for their threat to end school early, but they helped make the mess.
How can we win trans and queer liberation?
Confronting anti-trans bigotry is not merely a matter of political principle for socialists, but a life-and-death struggle against erasure.
The Democratic Party doesn't deserve your vote
Millions of people will vote against the Republicans in the midterm elections, but it's not because the Democrats are offering an alternative.
Interview: Phil Gasper
Is there still time to stop climate catastrophe?
The predictions in the UN report on climate change are dire and demand drastic change. But the obstacle isn't technical — it's political.
Julian Guerrero and Carlos Perez
Driven to death by Uber
Worsening conditions for New York City taxi drivers have led to five suicides in the past year — but drivers are standing up to demand change.
Elizabeth Schulte and Dorian Bon
The uprising of the airports
In city after city and airport after airport, people are turning out to protest Trump's executive orders declaring war on immigrants and refugees.


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Everything sold in the COH° Shop will go directly to the Thomas Fire Fund in partnership with the United Way of Ventura. You can make any purchase in the shop and 100% of the donations will go to these families or if you would rather donate directly, here is compiled list of options to give back.
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text:
/ News / Is ME/CFS caused by hydrogen sulphide? And is there is a new diagnostic test for ME/CFS? | 2 June 2009
Is ME/CFS caused by hydrogen sulphide? And is there is a new diagnostic test for ME/CFS? | 2 June 2009
Following a press conference at the Ritz Hotel in London on Thursday May 28, a number of UK newspapers have been reporting that the cause of ME/CFS has been identified and that a simple home-based diagnostic test for ME/CFS is now commercially available to the general public.
Daily Telegraph story
The urine test is based on the new scientific hypothesis that people with ME/CFS are producing excessive amounts of a chemical called hydrogen sulphide (H2S) and that this abnormality can be measured by a specific urine test (that measures an H2S metabolite/by-product).
The production of excessive amounts of this chemical (which can act as a mitochondrial* poison) is claimed to be due to overgrowth of certain lactate-producing bacteria in the gut (including species of enterococcus and streptococcus). This is coupled with the presence of metal (such as mercury and nickel) intoxication in the body. It is also claimed that the problem in the gut can be successfully treated through changes in diet, probiotics ('healthy bacteria') and antibiotics.
The urine test kit, designed for use at home, is only available privately at a basic cost of around £13. This test kit has to be ordered from abroad and is not available on the NHS.
Dr Charles Shepherd, Medical Adviser to the MEA, adds a note of caution:
"I have looked at the scientific information upon which this test is based and heard the presentations from Professor Kenny De Meirleir and Dr Chris Roelant – two of the people involved with this research – at the Invest in ME Conference on Friday 29 May. My conclusion is that while this is an interesting hypothesis, the test itself cannot yet be regarded as a scientifically proven diagnostic test for ME/CFS.
"The urine test needs to be further validated using significant numbers of ME/CFS patients with all degrees of severity and from various other referral centres. The results need to be compared to significant numbers of healthy matched controls, people with other conditions that involve fatigue, and people who are bed-bound or severely affected by other disabling conditions that may affect their nutritional status. The latter point, which was made at the conference on Friday, is particularly important because much of the work so far appears to relate to a severely affected sub-group of ME/CFS patients. The MEA's Ramsay Research Fund would be willing to consider any such research proposal. The results then need to be published in peer-reviewed medical journals.
"There are also going to be problems if people start using this test and then expect their family doctors/GPs to interpret the results and recommend/prescribe specific treatment, including antibiotics, based on the results. This is because the UK medical profession has not yet received any information about either the underlying hypothesis, or the test, or the treatment recommendations, in their scientific journals.
"Until we have further results from several good quality independent studies, it would be premature to conclude that a significant factor in the causation of ME/CFS has been discovered and that a simple urine test is now available for diagnosing ME/CFS. The recommendations regarding treatment are speculative and need to be subjected to equally rigorous clinical trials before any firm conclusions can be drawn about their general efficacy in ME/CFS."
If anyone has made use of this test please let us know how much it cost in total and what happened when the results were discussed with a doctor.
* Mitochondria are the energy-producing parts of cells.


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