text:
Ghostbusters: Afterlife or Stranger Things? Which One Is It?
by Neilan Adams
Fans of the beloved '80s franchise were more than disappointed with 2016's Ghostbusters-Answer the Call. Identity politics and a story which failed to connect with audiences saw the third sequel flounder at the box-office. In order to course-correct, Sony greenlit another reboot, Ghostbusters: Afterlife. Jason Reitman, son of Ivan Reitman (who directed the first two films), was tasked with directing and writing. For angered fans, it seemed like a good sign. Perhaps Sony learnt their lesson and wouldn't alter the Ghostbusters formula too radically again.
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In January, Reitman released a short teaser with the old Ghostbusters Ecto-1 Cadillac, using the original audio files for the sound of the proton pack and the original lettering used in the first films' promotional material. This generated a lot of excitement and seemed to make good on the promise of a return to the essence of what made the originals great. To placate the negative sentiments, Reitman in an interview with Bill Burr said that his upcoming film would "go back to original technique and hand the movie back to the fans…" He goes further in saying, "…This is gonna be a love letter to Ghostbusters… I want to make a movie for my fellow Ghostbusters fans."
Fast forward to a few days ago and the release of another teaser trailer. The reception varied. Some fans are excited, just happy to erase the memory of the disastrous 2016 reboot. However, others, like myself, were a bit underwhelmed and disappointed.
Why was there such a focus on children as stars of the film? What were we missing? Was this a trailer for a new Stranger Things season?
The children referred to are Stranger Things star, Finn Wolfhard and McKenna Grace, who stars in Netflix's Fuller House series. In Ghostbusters: Afterlife, Finn plays Trevor while Grace plays his sister Phoebe. Though Paul Rudd stars in the film as teacher Mr Grooberson, he seems to play second fiddle to the child stars. Perhaps the teaser only makes this appear to be the case and a more adult-centric film will emerge with the next trailer, especially as we know that most of the original cast from the '80s films will appear but, based on the trailer and Reitman's inspiration for the script, an adult-led film seems unlikely, which is a far cry from the original film we all love.
Whichever type of Ghostbusters film we see next year, it's clear that children play a large if not dominant role in the story. In a bizarre twist, the '80s nostalgia wave which inspired and gave impetus to creations like Stranger Things has been channelled back on itself, spawning Ghostbusters: Afterlife.
In an interview at Ghostbusters Fan Fest, Reitman confesses that the inspiration for the new film occurred after a character popped into his head one day, this character was a 12-year-old girl. The director stated, "I don't know who she was or why she popped into my head, but I saw her with a proton pack in her hand." By revealing thoughts such as these, it seems entirely possible that McKenna Grace will be the lead in the film and might be a strong clue to the direction of the film.
Children as leads in a Ghostbusters film might be taking it too far, even further than an all-female Ghostbusters team. Reitman seems to believe that anyone can be a Ghostbuster, even children. If this is the case, then he and his fellow writer Gil Kenan are perhaps making a film closer to Monster House or Goosebumps.
Most of the original cast of the older films are confirmed to appear in the new film. The only question is to what extent they will be part of the story. Will they have cameos or be characters integral to the plot? Let's hope their involvement is more than superficial and let's keep fingers crossed that if this is indeed a child-led film that it will be executed well enough to win fans over. Sony and Reitman might bury the franchise for good if their tinkering with the franchise goes awry or it might become like Predator franchise with their continuing bad sequels incapable of capturing the magic of the original.
Tags: GhostbustersStranger Things
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text:
\section{Introduction}
\label{sec:intro}

The identity of the dark matter (DM) is one of the biggest mysteries in particle physics, astrophysics, and cosmology.
Among various DM candidates, the lightest neutralino in the supersymmetric (SUSY) extension of the Standard Model (SM) is particularly attractive, since it can have the desired thermal relic abundance with a weak scale mass.

The discovery of the 125 GeV Higgs boson~\cite{Aad:2012tfa,Chatrchyan:2012ufa} and no signal of physics beyond the SM so far at the LHC run I might imply that the SUSY particles, in particular the scalar partners of SM fermions (sfermions), are much heavier than ${\cal O}(0.1$--$1)~\text{TeV}$.
After the Higgs discovery, therefore, such heavy sfermion scenarios have attracted attentions (see, e.g., Refs.~\cite{heavy_sfermion}).

If all sfermions are heavy, the lightest neutralino can have the correct thermal relic abundance, $\Omega_\chi\simeq 0.120$~\cite{Agashe:2014kda,1502.01589}, only in limited cases.
For instance, the pure Wino DM with a mass of $\simeq 3$ TeV can explain the DM density~\cite{hep-ph/0610249,0706.4071}, which may be probed by indirect detections~\cite{Wino-indirect}.
The (almost) pure Higgsino with a mass of $\simeq 1$ TeV is also an attractive candidate~\cite{0706.4071,1410.4549}.
In addition, the coannihilations among gauginos~\cite{gaugino-coannihilation-early,hep-ph/0601041} and the well-tempered Bino-Higgsino mixng~\cite{hep-ph/0601041,hep-ph/0004043,hep-ph/0406144} can lead to the desired DM density, with various phenomenological implications (see, e.g., Refs.~\cite{gaugino-coannihilation-recent,1211.4873} and references therein).

In this paper, we study another viable corner of the neutralino parameter space in the heavy sfermion scenario, i.e., the Higgs- and $Z$-reonant neutralino DM.
When the mass of the lightest neutralino is close to the Higgs- or $Z$-resonance, $m_{\chi_1}\sim m_h/2$ or $m_Z/2$, it can have the correct thermal relic abundance while avoiding the constraints from the direct detection and other experiments.
Aspects of such Higgs- and $Z$-resonant neutralino DM have been investigated e.g., in Refs.~\cite{Drees:1992am,hep-ph/0406144,HZpoleDM,1410.5730}.\footnote{
Higgs- and $Z$-resonant DMs are also realized in non-SUSY models with a singlet Majorana fermion and a SU(2)-doublet Dirac fermion. See e.g.,~Refs.~\cite{non-SUSY-HZpole}.
}

We revisit this Higgs- and $Z$-reonant neutralino DM scenario and extend previous studies by comprehensively investigating the current constraints and future prospects.
We include the following constraints and prospects.
\begin{itemize}
\item relic abundance $\Omega_\chi h^2 = 0.120$~\cite{Agashe:2014kda,1502.01589}.
\item DM direct detection.
\begin{itemize}
\item constraints on the spin-independent (SI) scattering cross section from the LUX~\cite{1310.8214},
and on the spin-dependent (SD) scattering cross section from the XENON 100~\cite{1301.6620}.
\item prospects of the XENON 1T for the SI~\cite{1206.6288} and SD~\cite{1211.4573} scattering cross sections.
\end{itemize}
\item Higgs invisible decay.
\begin{itemize}
\item current constraint from global fit~\cite{1303.3570,1306.2941}.
\item expected sensitivity of the HL-LHC~\cite{1309.7925,1310.8361}, and of the ILC~\cite{1310.0763}.
\end{itemize}
\item chargino/neutralino search at the LHC.
\begin{itemize}
\item expected sensitivity at 14 TeV for 300 fb$^{-1}$ and 3000 fb$^{-1}$~\cite{ATL-PHYS-PUB-2014-010}.
\end{itemize}
\end{itemize}
Other constraints are also briefly discussed.
Constraints from the LHC run I~\cite{1402.7029} is discussed in Appendix~\ref{app:8TeV}.
We use a simplified model with only three parameters, the Bino mass $M_1$, the Higgsino mass $\mu$, and $\tan\beta$, assuming that all other supersymmetric particles and heavy Higgs bosons are decoupled.
As we will see, the ``blind spot''~\cite{1211.4873} of the Higgs-neutralino coupling plays an important role also in the Higgs- and $Z$-resonant neutralino DM scenario.
It is shown that there is still a large viable parameter space, and almost all the parameter space of the scenario will be covered complementarily by the future experiments.
Our results are summarized in Figs.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}.

\section{Model}
\label{sec:model}
We assume that all SUSY particles but Bino and Higgsino multiplets are (much) heavier than 1 TeV, and the only scalar particle at the electroweak scale is the SM-like Higgs boson with the mass of 125 GeV.\footnote{
In the minimal SUSY SM, the Higgs mass is raised by the stop-loop~\cite{Higgs-stop} and the 125 GeV Higgs mass implies stop mass of ${\cal O}(1$--$10)~\text{TeV}$ or heavier, depending on $\tan\beta$ and the $A$-term~\cite{Higgs-recent}.
For simplicity, we set the Higgs mass 125GeV and do not consider the heavier particles' effects.
}
In this limit, the Wino component is decoupled, and the neutralino mass matrix becomes a $3\times 3$ matrix,
\begin{align}
M_\chi &=
\begin{pmatrix}
M_1 & -m_Z s_W \cos\beta & m_Z s_W \sin\beta \\
-m_Z s_W \cos\beta & 0 & -\mu \\
m_Z s_W \sin\beta &  -\mu & 0
\end{pmatrix}\,,
\label{eq:massmatrix}
\end{align}
in the basis of Bino and down-type and up-type Higgsinos, $(\widetilde{B}, \widetilde{H}^0_d, \widetilde{H}^0_u)$.
Here, $M_1$ and $\mu$ are the Bino and Higgsino masses, respectively, $\tan\beta=\vev{H_u}/\vev{H_d}$ is the ratio of the vacuum expectation values of the up- and down-type Higgs, and $s_W=\sin\theta_W$.
In this work, we assume that there is no CP-violation in the neutralino sector, and take $M_1>0$ and $\mu=$ real.
Then, the mass matrix is diagonalized by a real orthogonal matrix $O_\chi$ as
\begin{align}
O_\chi  M_\chi O_\chi^T
&=\text{diag}\left(\epsilon_1 m_{\chi_1}, \epsilon_2 m_{\chi_2}, \epsilon_3 m_{\chi_3}\right)\,,
\label{eq:massandmixing}
\end{align}
where $\epsilon_i=\pm 1$ and $0<m_{\chi_1}<m_{\chi_2}<m_{\chi_3}$.
Analytical and approximate formulae for the masses $m_{\chi_i}$ and the mixing matrix $O_\chi$ are given in Appendix~\ref{app:massandmixing}.
The lightest neutralino $\chi_1$ is the DM candidate.
In the chargino sector, there is only one light chargino with mass $m_{\chi^\pm} = |\mu|$.
We assume $m_{\chi^\pm}>100~\text{GeV}$ to avoid the LEP bound~\cite{Agashe:2014kda}.

In the following, we only consider a light Bino,
\begin{align}
M_1 < 80~\text{GeV}\,.
\end{align}
The lightest neutralino then becomes Bino-like, and its coupling to the SM is given by the following Lagrangian,
\begin{align}
{\cal L}_{\chi_1\text{-SM}}
&=\frac{1}{2}\lambda^h h\;\overline{\psi_1}\psi_1
-\frac{1}{2}\lambda^Z Z_\mu \overline{\psi_1}\gamma^\mu\gamma^5 \psi_1\,,
\label{eq:DM-SM}
\end{align}
where $h$, $Z_\mu$ and $\psi_1$ denote the fields of the the SM-like Higgs boson, $Z$-boson, and the lightest neutralino, respectively.
The DM thermal relic abundance, its direct and indirect detections, and the Higgs and $Z$ invisible decay rates are all determined by the DM mass $m_{\chi_1}$ and the two couplings $\lambda^h$ and $\lambda^Z$.
The couplings are given by\footnote{
We have used the relation in the decoupling limit of the Higgs sector, $\alpha\simeq \beta-\pi/2$ with $\alpha$ being the mixing angle of the Higgs sector.
We also neglect radiative corrections by heavier particles.
}
\begin{align}
\lambda^h &=
g'\epsilon_1 \OiB{1} \left( \OiHd{1}\cos\beta - \OiHu{1}\sin\beta \right)\,,
\label{eq:lambdaH} \\
\lambda^Z &=
\frac{1}{2} g_Z \left( - [\OiHd{1}]^2 + [\OiHu{1}]^2 \right)\,,
\end{align}
where $g_Z=g/\cos\theta_W=g'/\sin\theta_W$, and $g$ and $g'$ are the SU(2)$_L$ and U(1)$_Y$ gauge couplings, respectively.

In terms of ${\cal O}(m_Z s_W / \mu)$ expansion, they are approximately given by (cf. Appendix~\ref{app:massandmixing})
\begin{align}
\lambda^h &\simeq
g' \epsilon_1 \left( \frac{\mu \sin 2\beta + M_1}{\mu^2-M_1^2} m_Z s_W + {\cal O} \lrfp{m_Z s_W}{\mu}{3} \right)\,,
\label{eq:lambdaHapprox}\\
\lambda^Z &\simeq
\frac{1}{2}g_Z \left( \cos 2\beta \frac{m_Z^2 s_W^2}{\mu^2-M_1^2} + {\cal O} \lrfp{m_Z s_W}{\mu}{4} \right)\,.
\label{eq:lambdaZapprox}
\end{align}
From Eq.~\eqref{eq:lambdaHapprox}, the DM-$h$ coupling vanishes when
\begin{align}
M_1\simeq -\mu\sin 2\beta\,.
\label{eq:blindspot}
\end{align}
This leads to a ``blind spot"~\cite{1211.4873}, where the Higgs resonant annihilation, the spin-independent DM scattering for the direct detection, and the Higgs invisible decay are all suppressed.\footnote{
In Ref.~\cite{1211.4873}, all blind spots relevant to spin-independent and spin-dependent scattering are identified, and their phenomenological implications are investigated in the region of heavier neutralino DM, $m_{\chi_1}\gtrsim m_W$.
In the present scenario, only the one of Eq.~\eqref{eq:blindspot} is realized among those blind spots.
In this work, we do not consider the case of $\tan\beta \simeq 1$, which would lead to another blind spot for SD-scattering. The result for $\tan\beta \simeq 1$ will be similar to the case of $\tan\beta=2$ except for the SD scattering.
}
This suppression of $\lambda^h$ results in a parameter region that is not probed by the direct detection nor the Higgs invisible decay, which is important especially for the $Z$-resonant region $m_{\chi_1}\sim m_Z/2$.
On the other hand, the DM-$Z$ coupling is almost independent of $\tan\beta$ for $\tan\beta\gg 1$, and has only a mild dependence on $\tan\beta$ as far as $\tan\beta\gtrsim 2$.

At the LHC, searches for charginos and neutralinos can probe this scenario, which will be discussed in Sec.~\ref{subsec:LHC}.
The relevant interaction Lagrangian is given by
\begin{align}
{\cal L}&= W_\mu^- \sum_{i=1}^3 \overline{\psi_i}\gamma^\mu \left( \lambda_{LiC}^{W} P_L + \lambda_{RiC}^{W} P_R \right) \psi_C +\text{h.c.}
\nonumber\\
&+ Z_\mu \sum_{i<j} \overline{\psi_i}\gamma^\mu \left( \lambda_{Lij}^{Z} P_L + \lambda_{Rij}^{Z} P_R \right) \psi_j
\nonumber\\
&+ h \sum_{i<j} \overline{\psi_i} \left( \lambda_{Lij}^{h} P_L + \lambda_{Rij}^{h} P_R \right) \psi_j\,,
\label{eq:LforLHC}
\end{align}
where
\begin{align}
\lambda_{LiC}^{W}& = -\frac{1}{\sqrt{2}}g\eta_i \OiHu{i} \,,
\label{eq:LiCW} \\
\lambda_{RiC}^{W}&= \text{sign}(\mu) \frac{1}{\sqrt{2}}g\eta_i^* \OiHd{i} \,,
\label{eq:RiCW}\\
\lambda_{Lij}^Z = - (\lambda_{Rij}^Z)^* &= \frac{1}{2}g_Z \eta_i\eta_j^* \left( - \OiHd{i}\OiHd{j} + \OiHu{i}\OiHu{j} \right) \,,
\label{eq:LijZ}\\
\lambda_{Lij}^{h} = (\lambda_{Rij}^{h})^* &= \frac{1}{2}g'\eta_i^*\eta_j^* \left[ \OiB{i} \left( \OiHd{j}\cos\beta - \OiHu{j}\sin\beta \right) + (i\leftrightarrow j) \right]\,.
\label{eq:Rijh}
\end{align}
Here, $\eta_i^2=\epsilon_i$, and $\psi_C$ denotes the chargino field which is defined to have a mass term $-{\cal L}=\text{sign}(\mu)\mu \overline{\psi_C}\psi_C$ and to have a positive charge.

\section{Constraints and Prospects}
Our main results are shown in Figs.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}, where the constraints and prospects listed in Sec.~\ref{sec:intro} are presented in the $(m_{\chi_1}, m_{\chi^\pm})$-planes for $2 \leq \tan\beta \leq 50$.
In the figures, we show only the region with $30~\text{GeV} \le m_{\chi_1}\le 70~\text{GeV}$, because the relic density is always too large outside this region for $m_{\chi^\pm}>100~\text{GeV}$ and $M_1<80~\text{GeV}$.
In the following subsections, we explain each of the constraints and prospects in turn.
We also briefly mention other possible constraints in Sec.~\ref{subsec:others}.

\subsection{Thermal relic abundance}
\label{subsec:relic}
We assume that the present energy density of DM is dominantly given by that of the thermal relic of the lightest neutralino.
In the present scenario, the lightest neutralino can only annihilate into a pair of SM fermions, and the annihilation cross section is given by
\begin{align}
\sigma(\chi_1\chi_1\to f\bar{f}) = \sigma(\chi_1\chi_1\to h^*\to f\bar{f}) + \sigma(\chi_1\chi_1\to Z^*\to f\bar{f})\,,
\end{align}
where
\begin{align}
\sigma(\chi_1\chi_1\to h^*\to f\bar{f})
&\simeq
\frac{1}{2} (\lambda^h)^2 \sqrt{1-\frac{4m_{\chi_1}^2}{s}} \frac{1}{(s-m_h^2)^2+(m_h\Gamma_h)^2} \frac{s}{m_h} \Gamma(h\to f\bar{f})\,,
\label{eq:annihilationH} \\
\sigma(\chi_1\chi_1\to Z^*\to f\bar{f})
&\simeq
(\lambda^Z)^2 \sqrt{1-\frac{4m_{\chi_1}^2}{s}} \frac{1}{(s-m_Z^2)^2+(m_Z\Gamma_Z)^2} \frac{s}{m_Z} \Gamma(Z \to f\bar{f})\,,
\label{eq:annihilationZ}
\end{align}
and there is no interference term.
Here, we have neglected the terms proportional to $(m_f/m_{\chi_1})^2\ll 1$, where $m_f$ is the mass of final state fermion.
The DM abundance is calculated by solving the Boltzmann equation
\begin{align}
\frac{dY_{\chi_1}}{dt} = -n_S\vev{\sigma v_{\text{rel}}}(Y_{\chi_1}^2-Y_{\chi_1,\text{eq}}^2)\,,
\end{align}
where the thermal average of the annihilation cross section times the relative velocity (with Maxwell-Boltzmann equilibrium distribution) is given by~\cite{Gondolo:1990dk},
\begin{align}
\vev{\sigma v_{\text{rel}}}(T)
&= \frac{ \int d^3p_1 d^3p_2\; e^{-E_1/T}e^{-E_2/T} \sigma  v_{\text{rel}} }{ \int d^3p_1 d^3p_2\; e^{-E_1/T}e^{-E_2/T} }
\nonumber\\
&= \frac{1}{8m_{\chi_1}^4T [K_2(m_{\chi_1}/T)]^2} \int^\infty_{4m_{\chi_1}^2}\sigma (s)\sqrt{s}(s-4m_{\chi_1}^2)K_1(\sqrt{s}/T) ds\,.
\end{align}
Here, $Y_{\chi_1\,(\text{eq})}=n_{\chi_1\,(\text{eq})}/n_S$, $n_S=(2\pi^2/45)g_{*S}T^3$ is the entropy density, $n_{\chi_1}$ is the DM number density, $n_{\chi_1,\text{eq}}=2(m_{\chi_1}^2 T/2\pi^2)K_2(m_{\chi_1}/T) \simeq 2(m_{\chi_1}T/2\pi)^{3/2}\exp(-m_{\chi_1}/T)$ is its equilibrium value, and $t$ and $T$ are the cosmic time and the temperature, respecitvely.\footnote{
$T$ and $t$ are related by $dt/dT=(HT)^{-1}[1+(1/3)d(\ln g_{*S})/d(\ln T)]$, where $H=(\pi^2 g_*/90)^{1/2}T^2/M_P$ is the Hubble parameter with the reduced Planck scale $M_P\simeq 2.44\times 10^{18}~\text{GeV}$.
For the effective degrees of freedom $g_*(T)$ and $g_{*S}(T)$, we have used the fitting formula in~\cite{0910.1066}.
}
$K_{1,2}$ are the modified Bessel functions of the first and second kind.
The final relic abundance is given by $\Omega_{\chi_1} = m_{\chi_1}Y_{\chi_1}/(\rho_c/s)_0 $, where $(\rho/s)_0 \simeq 3.65 h^2\times 10^{-9}~\text{GeV}$ is the critical density divided by the entropy density at present, with $h\simeq 0.67$ being the scale factor for Hubble constant~\cite{Agashe:2014kda}.\footnote{
We have also calculated the DM abundance with micrOMEGAs~\cite{micrOMEGAs}, and checked that the results agree within a few \%.
}

In Figs.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}, the contours of relic DM density $\Omega_{\chi_1}h^2=0.120$ is shown in black lines.
The contours have clear peaks at the $Z$- and Higgs-resonances, $m_{\chi_1}\sim m_Z/2\simeq 45~\text{GeV}$ and $m_{\chi_1}\sim m_h/2\simeq 62~\text{GeV}$.
In these regions, the chargino mass $m_{\chi^\pm}$ can take a large value, corresponding to small DM-$Z$ and DM-Higgs couplings, $\lambda^Z$ and $\lambda^h$.

In the $Z$-resonant region, $m_{\chi_1}\sim m_Z/2$, the relic abundance shows a universal behavior for all $\tan\beta\gg 1$.
This is because the DM-$Z$ coupling $\lambda^Z$ is almost independent of $\tan\beta$ for $\tan\beta\gg 1$, as shown in Eq.~\eqref{eq:lambdaZapprox}.
In this region, the chargino mass $m_{\chi^\pm}$ is always bounded from above as $m_{\chi^\pm}\lesssim 450~\text{GeV}$~\cite{1410.5730}, which corresponds to $|\lambda^Z|\gtrsim 0.0034$ [cf. Eq.~\eqref{eq:lambdaZapprox}].\footnote{
This is consistent with the analysis of generic $Z$-portal DM in Ref.~\cite{1411.2985}.
}
This upper bound on the chargino mass is crucial for the LHC search discussed in Sec.~\ref{subsec:LHC}.

In the Higgs resonant region, $m_{\chi_1}\sim m_h/2$, the behavior of the relic abundance in the $(m_{\chi_1},m_{\chi^\pm})$-planes strongly depends on $\tan\beta$ as well as the sign$(\mu)$.
This  is also understood in terms of the DM-Higgs coupling, $\lambda^h$, in Eq.~\eqref{eq:lambdaHapprox}.
As can be seen in Figs.~\ref{fig:main_tb45and6}, for $\mu<0$ and $4\lesssim \tan\beta \lesssim 6$, there are two regions corresponding to $\Omega_{\chi_1}h^2\le 0.12$.
This is because of the blind-spot behavior discussed in Sec.~\ref{sec:model}.
The coupling $\lambda^h$ has  opposite signs in the two separate regions, and it becomes zero in between.
For $\mu<0$ and $\tan\beta\gtrsim 7$, the region of large $m_{\chi^\pm}$ disappears because a sufficiently large $|\lambda^h|$ can no longer be obtained there.
For both $\mu<0$ and $\mu>0$ and for all $\tan\beta$, the maximal chargino mass corresponds to $|\lambda^h| \simeq 0.0052$.
For $\tan\beta\gtrsim 10$, the upper bound on the chargino mass is as small as $m_{\chi^\pm}\lesssim 400~\text{GeV}$ for $\mu<0$ and $m_{\chi^\pm}\lesssim (500$--$800)~\text{GeV}$ for $\mu>0$.
As we shall see in Sec.~\ref{subsec:LHC}, these regions can be probed by the 14 TeV LHC.
For small $\tan\beta$, however, a much larger chargino mass is allowed, e.g., $m_{\chi^\pm}\lesssim 2.5~\text{TeV}$ for $\tan\beta=2$ and $\mu>0$.
Although such a heavy chargino is out of the 14 TeV LHC reach, the direct detection experiments can cover most of the region, as we will see in the next subseciton.

\subsection{Direct detection}
\label{subsec:direct}
In the present scenario, the spin-independent (SI) and spin-dependent (SD) scatterings between the DM and nuclei are induced by Higgs-exchange and $Z$-exchange, respectively.
As we shall see, the former gives a strong bound and high future sensitivity, while the latter plays a complementary role in the blind spot regions for $\mu<0$.

\subsubsection{spin-independent scattering}
The SI scattering cross section of DM per nucleon is given by
\begin{align}
\sigma^{\text{SI}}_N &=
\frac{4}{\pi}\lambda_N^2 m_N^2 \left(1+\frac{m_N}{m_{\chi_1}}\right)^{-2}\,,
\label{eq:SIXS}
\end{align}
where $m_N$ is the nucleon mass and $\lambda_N$ is the effective coupling between the DM and nucleons, ${\cal L}=\sum_{N=p,n} \lambda_N\overline{\psi_1}\psi_1 \overline{N} N$.
In our scenario, the coupling is induced by the Higgs exchange, and given by~(cf.~\cite{Shifman:1978zn,micrOMEGAs})
\begin{align}
\lambda_N &= \frac{\lambda^h}{2m_h^2} \cdot \frac{m_N f_N}{2m_W/g}\,,
\quad
f_N= \sum_{q=u,d,s} f_{T_q}^{(N)} + \frac{2}{9} \left(1-\sum_{q=u,d,s} f_{T_q}^{(N)}\right)\,,
\end{align}
where $f_{T_q}^{(N)}=\vev{N|m_q \bar{q}q|N}/m_N$.
In our analysis, we use the default values of $f_{T_q}^{(p)}$ adopted in micrOMEGAs 4.1~\cite{micrOMEGAs}, $f_{T_u}^{(p)}=0.0153$, $f_{T_d}^{(p)}=0.0191$, and $f_{T_s}^{(p)}=0.0447$, which leads to $f_N = 0.284$.\footnote{
Note that the strange quark mass fraction $f_{T_s}^{(p)}$ is already small.
If we use the result of the lattice calculation in Ref.~\cite{1208.4185}, which is further smaller as $f_{T_s}^{(p)}\simeq 0.009$, it leads to $f_N\simeq 0.256$, resulting in 20\% smaller SI scattering cross sections.
We have also calculated the SI cross section with micrOMEGAs 4.1~\cite{micrOMEGAs}, and checked that the results agree within a few \%.
}
Therefore, the SI scattering cross section is given by
\begin{align}
\sigma^{\text{SI}}_N \simeq
5.2\times 10^{-43}\cdot (\lambda^h)^2 \left(1+\frac{m_N}{m_{\chi_1}}\right)^{-2}~\text{cm}^2\,.
\end{align}

In Fig.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}, we show the constraint obtained by the LUX~\cite{1310.8214} (90\% CL limit) and the future prospects for the XENON 1T~\cite{1206.6288} with blue dashed and solid lines, respectively.
The results are understood in terms of the coupling $\lambda^h$ in Eqs.~\eqref{eq:lambdaH} and \eqref{eq:lambdaHapprox}.
As can be seen in the figures, for $\mu>0$, the region with $m_{\chi^\pm}\lesssim 120$--400~GeV (150--400~GeV) are excluded by the LUX, for $m_{\chi_1}\sim m_Z$ ($m_{\chi_1}\sim m_h$).
The XENON 1T can cover most of the viable parameter space for $\mu>0$, except for the peak of the Higgs-resonance, where $\lambda^h\simeq 0.0052$ (cf. Sec.~\ref{subsec:relic}) and $\sigma^{\text{SI}}_N\simeq 1.4\times 10^{-47}\text{cm}^2$, and the $Z$ peaks for $\tan\beta\gtrsim 30$.
Note that the SI cross sections in these peak regions are just below the sensitivity shown in Ref.~\cite{1206.6288}.
Therefore, it is expected that future experiments with higher sensitivity~\cite{1310.8327} can cover the whole parameter region for $\mu>0$.

For $\mu<0$, because of the cancellation in Eq.~\eqref{eq:lambdaHapprox}, the constraint and  the sensitivity are significantly reduced in terms of the chargino mass $m_{\chi^\pm}$.
In the Higgs-resonant region, the parameter regions with the correct thermal relic abundance, $\Omega_{\chi_1}h^2\simeq 0.12$, will still be mostly covered by the XENON 1T.
This is because both of $\Omega_{\chi_1}h^2$ and $\sigma^{\text{SI}}_N$ are determined by the same coupling, $\lambda^h$.
The correlation is clearly seen, e.g., in Fig.~\ref{fig:main_tb45and6}, for $5\lesssim \tan\beta\lesssim 6$.
In the $Z$-resonant region, however, $\Omega_{\chi_1}h^2$ and $\sigma^{\text{SI}}_N$ are determined by different couplings, $\lambda^Z$ and $\lambda^h$, respectively.
This results in a large parameter region which gives correct $\Omega_{\chi_1}h^2$ but very small $\sigma^{\text{SI}}_N$, as can be seen in Figs.~\ref{fig:main_tb78and9}--\ref{fig:main_tb3040and50}.
Since $\lambda^h$ can be zero at a certain value of $\mu$, there always remains a region which cannot be probed by the SI scattering.
Some of these regions are probed by SD scattering discussed in the next subsection, and the search for the chargino and the neutralino at the LHC, discussed in Sec.~\ref{subsec:LHC}, will be a very sensitive probe in these regions.

\subsubsection{spin-dependent scattering}
Now let us discuss the SD scattering.
The SD nucleon-DM scattering cross section is given by~\cite{hep-ph/9506380,micrOMEGAs}
\begin{align}
\sigma^{\text{SD}}_N & = \frac{12}{\pi}\xi_N^2m_N^2 \left(1+\frac{m_N}{m_{\chi_1}}\right)^{-2}\,,
\end{align}
where $\xi_N$ is the axial-vector effective coupling between the DM and the nucleons, ${\cal L}=\sum_{N=p,n} \xi_N\overline{\psi_1}\gamma^5\gamma_\mu\psi_1 \overline{N}\gamma^5\gamma^\mu N$.
In our scenario, the coupling is induced by the $Z$ exchange, and given by~(cf.~\cite{hep-ph/9506380,micrOMEGAs})
\begin{align}
\xi_N = \frac{\lambda^Z g_Z}{4m_Z^2}\sum_{q=u,d,s} T^3_q \Delta^{(N)}_q\,,
\end{align}
where $\Delta_q^{(N)}$ parametrize the quark spin content of the nucleon.
In our analysis, we use the default values adopted in micrOMEGAs~\cite{micrOMEGAs}, $\Delta_u^{(p)}=\Delta_d^{(n)}=0.842$, $\Delta_d^{(p)}=\Delta_u^{(n)}=-0.427$, and $\Delta_s^{(p)}=\Delta_s^{(n)}=-0.085$.
Therefore, the SD nucleon-DM scattering cross section is given by\footnote{
We have again checked the SD cross section with micrOMEGAs  and the results agree within 1 \%.
}
\begin{align}
\sigma^{\text{SD}}_{n (p)}
&\simeq 2.3 (3.0)\times 10^{-37}\cdot (\lambda^Z)^2 \left(1+\frac{m_{n (p)}}{m_{\chi_1}}\right)^{-2} \;\text{cm}^2\,.
\label{eq:SDXS}
\end{align}

In Fig.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}, we show the constraint on the SD neutron-DM scattering cross section $\sigma^{\text{SD}}_n$ from the XENON100~\cite{1301.6620} with green dashed lines.\footnote{
See also Ref.~\cite{1211.4573}.
The constraint on proton-DM scattering, $\sigma^{\text{SD}}_p$, is much weaker.
The LUX has not published constraints on SD scattering cross section.
If we adopt the constraint on SD scattering cross section in Ref.~\cite{1407.8257}, which is based on the LUX data, the bound on the chargino mass becomes about 20--30\% stronger.
}
As can be seen from the figures, this bound gives the strongest current constraint in some of the small $m_{\chi^\pm}$ regions for $\mu<0$.
The bound on the chargino is about $m_{\chi^\pm}\lesssim 100$--140~GeV, and it has only very mild dependences on  $\tan\beta$, $m_{\chi_1}$, and sign$(\mu)$.
This can be understood from Eqs.~\eqref{eq:lambdaZapprox} and \eqref{eq:SDXS}, which lead to the following approximate formula,
\begin{align}
\sigma^{\text{SD}}_{n (p)}
&\simeq 1.4 (1.8)\times 10^{-41} \lrfp{300~\text{GeV}}{|\mu|}{4} \left(1-\frac{m_{\chi_1}^2}{\mu^2}\right)^{-2} (\cos 2\beta)^2\;\text{cm}^2\,.
\end{align}
In the figures, we also show the prospect of the XENON 1T for the $\sigma^{\text{SD}}_n$, studied in Ref.~\cite{1211.4573}, with green solid lines.
It reaches the chargino mass of about 280--350~GeV, which will cover a large part of the blind spot for $\mu<0$.

\subsection{Higgs and $Z$ invisible decays}
For $m_{\chi_1}<m_h/2$ the Higgs boson can decay into a pair of DMs, which lead to a invisible decay.
The branching ratio is given by
\begin{align}
\text{Br}(h\to \chi_1\chi_1)
&= \frac{\Gamma(h\to \chi_1\chi_1)} {\Gamma(h\to\text{SM})+\Gamma(h\to \chi_1\chi_1)}\,,
\end{align}
where
\begin{align}
\Gamma(h\to \chi_1\chi_1)=\frac{(\lambda^h)^2}{16\pi} m_h \left(1-\frac{4m_{\chi_1}^2}{m_h^2} \right)^{3/2}\,.
\end{align}
We have used $\Gamma(h\to\text{SM})=4.07~\text{MeV}$~\cite{Agashe:2014kda} in our calculation.
The constraint on the Higgs invisible decay has been obtained by global fits to Higgs data~\cite{1303.3570,1306.2941}.
In our numerical calculation, we adopt the one in Ref.~\cite{1306.2941}, $\text{Br}(h\to \text{invisible})<0.19$ (95\% CL).

As for the future prospects, we consider the high-luminosity (HL) LHC and the ILC.
The sensitivity of the HL-LHC depends on the systematic uncertainties.
Here, we use the estimated future sensitivity of searches for Higgs decaying invisibly using $ZH$ channel in Ref.~\cite{1309.7925}, $\text{Br}(h\to \text{invisible})<0.062$ (95\% CL) for 3000~fb$^{-1}$, adopting the value of the ``realistic scenario'' for the size of systematics.
In the ``conservative scenario'' the estimated sensitivity becomes $\text{Br}(h\to \text{invisible})<0.14$.
In the Higgs working group report of the 2013 Snowmass~\cite{1310.8361}, the 95\% CL limit (for 3000~fb$^{-1}$) is estimated as $\text{Br}(h\to \text{invisible})<0.08$--0.16 for ATLAS and 0.06--0.17 for CMS.
For the ILC, we use the value in~\cite{1310.0763}, $\text{Br}(h\to \text{invisible})<0.004$ (1150 fb$^{-1}$ at $\sqrt{s}=250~\text{GeV}$).

The constraint and prospects for the Higgs invisible decay are shown in Fig.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50} with magenta lines.
As can be seen in the figures, a large parameter space is covered by the Higgs invisible decay search.
For $\mu>0$, the whole $Z$-resonant region will be covered by the ILC.
The blind spots are again clearly seen for $\mu<0$, where Higgs invisible decay becomes very small due to the suppression of $\lambda^h$ in Eq.~\eqref{eq:lambdaHapprox}.

Before closing this subsection, let us comment on the $Z$ invisible decay.
For $m_{\chi_1}<m_Z/2$, the $Z$ can decay into a pair of DMs, with a partial decay rate
\begin{align}
\Gamma(Z\to \chi_1\chi_1)=\frac{(\lambda^Z)^2}{24\pi} m_Z \left(1-\frac{4m_{\chi_1}^2}{m_Z^2} \right)^{3/2}\,.
\end{align}
The bound from the LEP, $\Gamma(Z\to \chi_1\chi_1)<2.0~\text{MeV}$ (95\% CL)~\cite{hep-ex/0509008}, corresponds to $\lambda^Z\lesssim 0.041\times (1-4m_{\chi_1}^2/m_Z^2)^{-3/2}$.
In our setup, this is always weaker than the bound on $\sigma^{\text{SD}}$ from the XENON100~\cite{1301.6620} for $m_{\chi_1}\gtrsim 30~\text{GeV}$~(see Sec.~\ref{subsec:direct}), and hence we do not show the bound in the figures.

\subsection{Search for the chargino and neutralinos at the LHC}
\label{subsec:LHC}
As we have seen in Sec.~\ref{subsec:relic}, the requirement that the thermal relic abundance of $\chi_1$ explains the observed DM density, $\Omega_{\chi_1}h^2\simeq 0.12$, gives upper bounds on the chargino mass $m_{\chi^\pm}$ in the present scenario.
The chargino mass is ${\cal O}(100~\text{GeV})$ except for the Higgs resonance peak for small $\tan\beta$.
In particular, in the $Z$-resonant region, the chargino mass $m_{\chi^\pm}$ is always bounded from above as $m_{\chi^\pm}\lesssim 450~\text{GeV}$~\cite{1410.5730}.
The heavier neutralinos, $\chi_2$ and $\chi_3$, also have masses $m_{\chi_{2,3}}\simeq m_{\chi^\pm}$.
These ${\cal O}(100~\text{GeV})$ chargino and neutralinos are within the reach of the LHC experiments.

In this work, we consider the following production and decay channels at the LHC,
\begin{eqnarray}
pp\to \chi_{2,3}\chi^\pm \to Z\chi_{1}\;W^\pm\chi_{1} \to \ell \ell \chi_{1}\; \ell \nu\chi_{1}\,,
\label{eq:WZatLHC}
\end{eqnarray}
which leads to a signal with three leptons and missing energy, and gives a high sensitivity in the present scenario~\cite{1409.6322}.
The sensitivities of the SUSY searches at 14 TeV, including the high luminosity run of $3000~\text{fb}^{-1}$, is studied by ATLAS~\cite{ATL-PHYS-PUB-2014-010} and CMS~\cite{CMS:2013ega}.
In this work, we reinterpret the ATLAS study~\cite{ATL-PHYS-PUB-2014-010} of the search for the channel in Eq.~\eqref{eq:WZatLHC}, for $300~\text{fb}^{-1}$ and $3000~\text{fb}^{-1}$.
The constraints from the LHC run I~\cite{1402.7029} are discussed in Appendix~\ref{app:8TeV}.

In the ATLAS analysis~\cite{ATL-PHYS-PUB-2014-010}, the results for the three-lepton process~\eqref{eq:WZatLHC} is presented assuming a simplified ``pure Wino" model with three states; two neutralinos $\chi_1$, $\chi_2$ and one chargino $\chi^\pm$, with degenerate masses of the chargino and the heavier neutralino, $m_{\chi^\pm}=m_{\chi_2}$ and 100\% branching ratios of $\text{Br}(\chi_2\to \chi_1 Z)=1$ and $\text{Br}(\chi^\pm \to \chi_1 W)=1$.
There are several differences in the present setup.
(i) There are two heavier neutralinos $\chi_2$ and $\chi_3$.
(ii) The sum of the production cross sections of chargino-neturalino pair is a factor $\simeq 1/2$ smaller, since the ATLAS analysis assumes Wino-like chargino-neutralino pair production.
(iii) The neutralinos $\chi_2$ and $\chi_3$ have generically sizable branching fractions to the Higgs as well, $\text{Br}(\chi_{2,3}\to \chi_1 h)={\cal O}(1)$ and $\text{Br}(\chi_{2,3}\to \chi_1 Z)={\cal O}(1)$.
(iv) Although their masses $m_{\chi_{2,3}}$ are close to the chargino mass $m_{\chi^\pm}$, the difference can have a non-negligible effect on the production cross section, especially for small $|\mu|$.
Therefore, the results in Ref.~\cite{ATL-PHYS-PUB-2014-010} cannot be directly applied to the present model and reinterpretations are necessary.

In the ATLAS analysis, several signal regions (SRs) are defined by various kinematical cuts.
We calculate the expected number of events in each signal region (SR) X as
\begin{align}
N_{\text{SR-X}}
= \sum_{j=2,3} & \sum_{\chi^\pm} \sigma^{\text{NLO}}(pp\to \chi^\pm\chi_j) \cdot \text{Br}(\chi^\pm\to\chi_{1}W^{(*)}\to\chi_1\ell\nu) \cdot \text{Br}(\chi_j\to\chi_{1}Z^{(*)}\to\chi_1\ell\ell)
\nonumber\\
&\times A_{\text{SR-X}} \cdot \int {\cal L}dt\,,
\label{eq:NSRX}
\end{align}
where $\int {\cal L}dt$ denote the integrated luminosity, and $A_{\text{SR-X}}$ is defined by
\begin{align}
A_{\text{SR-X}}
&= \frac{ \text{\# of events which pass the cuts of SR-X} }{ \text{\# of generated events in } pp\to \chi^\pm \chi_i\to W^{(*)} \chi_1 Z^{(*)}  \chi_1 \to \ell\nu \chi_1 \ell\ell \chi_1 }\,,
\label{eq:acc2}
\end{align}
where $\ell$ denotes $e$, $\mu$ and $\tau$.
For simplicity, we discard the hadronic decays of $W$ and $Z$.
The branching fractions are given by (when kinematically allowed)
\begin{align}
\text{Br}(\chi^\pm\to\chi_{1}W\to\ell\nu)
&= \text{Br}(\chi^\pm\to\chi_{1}W) \cdot \text{Br}(W\to\ell\nu)\,, \\
\text{Br}(\chi_j\to\chi_{1}Z\to\ell\ell)
&= \text{Br}(\chi_j\to\chi_{1}Z) \cdot \text{Br}(Z\to\ell\ell)\,,
\end{align}
where the chargino branching fraction is $\text{Br}(\chi^\pm\to\chi_{1}W)=1$, while the neutralino branching is given by
\begin{align}
\text{Br}(\chi_{j}\to Z\chi_1)
&= \frac{\Gamma_{j}(\chi_{j}\to Z\chi)}{\Gamma_{j}(\chi_{j}\to Z\chi)+\Gamma_{j}(\chi_{j}\to h\chi)}\,,
\label{eq:BranchingRatio}\\
\Gamma(\chi_{j}\to Z\chi)
&= \frac{1}{16\pi}m_{\chi_j}|\lambda_{L1j}^Z|^2 \left( 1+6\epsilon_1\epsilon_{j} r^1_{j}+(r^1_{j})^2-2(r^z_{j})^2+\frac{(1-(r^1_{j})^2)^2}{(r^z_{j})^2} \right)
\nonumber\\
&\times \left(1-(r^1_{j}-r^z_{j})^2\right)^{1/2}\left(1-(r^1_{j}+r^z_{j})^2\right)^{1/2}\,, \\
\Gamma(\chi_{j}\to h\chi)
&= \frac{1}{16\pi}m_{\chi_j}|\lambda_{L1j}^h|^2 \left( 1+2\epsilon_1 \epsilon_{j} r^1_{j} +(r^1_j)^2-(r^h_{j})^2 \right)
\nonumber\\
&\times  \left(1-(r^1_{j}-r^h_{j})^2\right)^{1/2}\left(1-(r^1_{j}+r^h_{j})^2\right)^{1/2}\,,
\end{align}
where $r^1_{j}=m_{\chi_1}/m_{\chi_j}, r^z_{j}=m_Z/m_{\chi_j}, r^h_{j}=m_h/m_{\chi_j}$, and the couplings are given by Eqs.~\eqref{eq:LijZ} and \eqref{eq:Rijh}.

In the numerical calculations, we generate the events with MadGraph5\_aMC@NLO 2.2.3~\cite{1405.0301} in combination with PYTHIA 6.4~\cite{hep-ph/0603175}.
We generate the events at LO and rescale the acceptance with the NLO cross section calculated by Prospino 2.1~\cite{prospino} with CTEQ6L1~\cite{hep-ph/0201195} parton distribution functions (PDFs).
Delphes 3~\cite{1307.6346} is used with the ATLAS parameters card\footnote{
$b$-tagging efficiencies and lepton isolation criteria are arranged as used in the ATLAS analysis~\cite{ATL-PHYS-PUB-2014-010}.
}
given in the MadGraph5\_aMC@NLO package for the fast detector simulation.

There are three (four) SRs considered to probe the signal of Eq.~\eqref{eq:WZatLHC} for 300 (3000) fb$^{-1}$, denoted as SRA--SRC (SRA--SRD).
In our analysis, after electrons, muons, and jets are selected following the ATLAS analysis~\cite{ATL-PHYS-PUB-2014-010},
the following cuts are applied.
\begin{itemize}
\item There should be exactly three leptons in each event, and at least one SFOS lepton pair is required to have invariant mass $|m_{\text{SFOS}}-M_Z|<10~\text{GeV}$.
\item Events with $b$-tagged jets are discarded.
\item The three lepton $p_T$ should be larger than 50 GeV.
\item Then, the events are divided into SRs depending on the missing transverse energy $E_T^{\text{miss}}$ and the transverse mass $m_T$, where $m_T$ is calculated with the missing transverse energy and the lepton which does not form the SFOS lepton pair whose mass is closest to the $Z$-boson mass.
\end{itemize}

As a validation of our analysis, we have calculated the expected numbers of events in each SRs for the ``pure Wino" model points with $(m_{\chi_2}, m_{\chi_1})=$ (400,0), (600,0), (800,0), (1000,0)~GeV, which are studied in the ATLAS analysis~\cite{ATL-PHYS-PUB-2014-010}.
They are in good agreement with the ATLAS analysis.

In the ATLAS analysis~\cite{ATL-PHYS-PUB-2014-010}, the expected 95\% exclusion limit is shown in the $(m_{\chi_2},m_{\chi_1})$-plane by combining disjoint versions of SRs.
We have analyzed the same parameter space as a validation.
In our analysis, the expected exclusion line is obtained as follows.
 (i) For each SR, the expected upper limit on the number of beyond-the-SM events is calculated from the number of the background events given in~\cite{ATL-PHYS-PUB-2014-010}, using $Z_N$=1.64~\cite{Linnemann:2003vw} for 95\% CL exclusion.
(ii) At each model point, the expected number of signal events in each SR is calculated.
(iii) The model point is excluded if and only if it is excluded in at least one of the SRs.
The obtained expected exclusion line, $m_{\chi^\pm} \lesssim$ 800 (1100) GeV for $m_{\chi_1} \lesssim 100~\text{GeV}$ at 300 (3000) fb$^{-1}$, agrees with the ATLAS result within an error of $\delta m_{\chi^\pm}\simeq 20~\text{GeV}$.

Next, we apply the same analysis in the present scenario, i.e., the Higgs- and $Z$-resonant neutralino DM.
The cross sections and acceptances are calculated as follows.
For $|\mu|=100, 110, \cdots 1000~\text{GeV}$, the cross sections are calculated at $(\tan\beta,\text{sign}(\mu),M_1)=(2,+,80~\text{GeV})$ and $(50,+,30~\text{GeV})$.\footnote{These points are chosen since they give the smallest, medium and largest value of $m_{\chi_j}-m_{\chi^\pm}$.}
Then, the cross sections normalized by the coupling, $\sigma^{\text{NLO}}/(|\OiHu{j}|^2+|\OiHd{j}|^2)$, are interpolated.
The acceptances are calculated by varying $m_{\chi^\pm}$, $m_{\chi_j}-m_{\chi^\pm}$ and $m_{\chi_1}$ by $(\Delta m_{\chi^\pm}, \Delta (m_{\chi_j}-m_{\chi^\pm}), \Delta m_{\chi_1})=(20,10,5)~\text{GeV}$ for 100$~\text{GeV}<|\mu|<300~\text{GeV}$ and $(\Delta m_{\chi^\pm}, \Delta (m_{\chi_j}-m_{\chi^\pm}), \Delta m_{\chi_1})=(20,10,10)~\text{GeV}$ for  300$~\text{GeV}<|\mu|<1000~\text{GeV}$, while the couplings are fixed as the ones of $\tan\beta=5, M_1=50~\text{GeV}$, $\mu=200~\text{GeV}$, for simplicity.\footnote{
We have checked that the acceptance does not depend much on these parameters.
}
Here, we do not consider the region of $m_{\chi_j}-m_{\chi_1} < m_Z$, for simplicity.

Results are shown in Fig.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50} with red lines.
The expected exclusion region at 300~fb$^{-1}$ is shown in light orange region with red dotted lines.
One can see that the $Z$-peaks in the whole parameter space, including the blind spot, will be probed at 300~fb$^{-1}$.
For $\tan\beta\geq 30$, the Higgs peaks can also be covered.
The small $m_{\chi^\pm}$ region is not covered because of the small mass differences between $\chi_{2,3}, \chi^\pm$ and $\chi_1$.

At 3000 fb$^{-1}$, much larger parameter space will be probed, up to $m_{\chi^\pm}\sim 800~\text{GeV}$.
The Higgs-resonant regions are covered for $\tan\beta\gtrsim 15$ ($\tan\beta\geq$ 6) for $\mu > 0$ ($\mu < 0$).
Though the small $m_{\chi^\pm}$ region can not be covered even at 3000 fb$^{-1}$, combination with other experiments such as the direct searches can probe almost all the parameter region of the present scenario.

As can be seen in the figures, the expected reach for the chargino mass, $m_{\chi^\pm}\sim 800~\text{GeV}$, is almost independent of $\tan\beta$ and $m_{\chi_1}$.
This can be understood as follows.
In the large $m_{\chi^\pm}$ region, the cross section is mainly determined by $|\mu|$ because the masses are almost degenerate as $m_{\chi_2}\simeq m_{\chi_3}\simeq m_{\chi^\pm}=|\mu|$, and the coupling with $W$-boson is universal for $|\mu|\gg m_Z$ [cf. Eqs.~\eqref{eq:LiCW},~\eqref{eq:RiCW},~\eqref{eq:OuHdj},~\eqref{eq:OuHuj}].
In addition, because of the large mass hierarchy $m_{\chi_{2,3}}\simeq m_{\chi^\pm}\gg m_{\chi_1}, m_Z, m_W$, the acceptance is determined almost only by $|\mu|$.
Thus, from Eq.~\eqref{eq:NSRX}, $N_{\text{SR-X}}$ becomes
\begin{eqnarray}
N_{\text{SR-X}}
\simeq &
\sum_{\chi^\pm} \sigma^{\text{NLO}}(pp\to \chi^\pm\chi_2) \cdot \text{Br}(\chi^\pm\to\chi_{1}W\to\chi_1\ell\nu)\cdot A_{\text{SR-X}} \cdot \int {\cal L}dt\nonumber
\\
&\times \sum_{j=2,3} \text{Br}(\chi_j\to\chi_{1}Z\to\chi_1\ell\ell)\,,
\end{eqnarray}
The first line of this equation is determined almost only by $|\mu|$.
The second line can be expanded in terms of ${\cal O}(m_Z s_W / \mu)$ as
\begin{eqnarray}
\text{Br}(\chi_2\to Z\chi_1)&=&\frac{1}{2}(1+\sin{2\beta})+\frac{M_1}{\mu}(1-\sin^2{2\beta})+{\cal O}\left(\frac{m_Z s_W}{\mu}\right)^2\,,\\
\text{Br}(\chi_3\to Z\chi_1)&=&\frac{1}{2}(1-\sin{2\beta})-\frac{M_1}{\mu}(1-\sin^2{2\beta})+{\cal O}\left(\frac{m_Z s_W}{\mu}\right)^2\,.
\end{eqnarray}
From this expression, $\text{Br}(\chi_2\to Z\chi_1)+\text{Br}(\chi_3\to Z\chi_1)\simeq 1$ for $|\mu|\gg m_Z$, and it is almost
independent of $\tan\beta$ and $m_{\chi_1}$.

\subsection{Other constraints}
\label{subsec:others}
Let us briefly comment on other possible constraints on the present scenario.
\begin{itemize}
\item
indirect search.
\\
The DM annihilation in the present Universe can lead to cosmic rays such as photons, positrons, and anti-protons.
In the present model, however, the annihilation cross section in the present Universe is suppressed by the velocity, $\vev{\sigma v_{\text{rel}}}_0 \sim v^2$, as shown in Eqs.~\eqref{eq:annihilationH} and \eqref{eq:annihilationZ}.
In the limit of $v\to 0$, the leading term in the amplitude comes from the $Z$-exchange diagram and is proportional to the mass of the final state fermion.
The annihilation cross section is approximately given by
\begin{align}
\vev{\sigma v_{\text{rel}}}_0 &\simeq \vev{\sigma v_{\text{rel}} (\chi_1\chi_1\to b\bar{b},\tau\bar{\tau})}_0
\nonumber \\
&\simeq \frac{g_Z^2}{32\pi} (\lambda^Z)^2 \frac{3m_b^2+m_\tau^2}{m_Z^4} \quad (v_{\text{rel}}\to 0)
\nonumber \\
&\sim 2.8\times 10^{-26}\cdot (\lambda^Z)^2 \;\text{cm}^3/s
\nonumber \\
&\sim 1.8\times 10^{-30} \lrfp{300~\text{GeV}}{|\mu|}{4} \left(1-\frac{m_{\chi_1}^2}{\mu^2}\right)^{-2} (\cos 2\beta)^2\;  \text{cm}^3/\text{s}\,,
\end{align}
where we have used running bottom quark mass $\overline{m}_b^{\overline{\text{MS}}}(100~\text{GeV})\simeq 3~\text{GeV}$ (cf.~\cite{hep-ph/0503172}) in the third line, and Eq.~\eqref{eq:lambdaZapprox} in the last line.\footnote{
We omitted the correction around the region of $|m_{\chi_1}-m_Z/2|\lesssim \Gamma_Z$, where $\vev{\sigma v_{\text{rel}}}_0$ is further suppressed.
We have also checked by the micrOMEGAs, and the results agree within ${\cal O}(1\%)$ and ${\cal O}(10\%)$ for $\tau\bar{\tau}$ and $b\bar{b}$ modes, respectively.
}
Therefore, it is at most ${\cal O}(10^{-28})~\text{cm}^3\text{s}^{-1}$, and much smaller in most of the parameter space, which is smaller than the constraints such as the Fermi-LAT bounds in Ref.~\cite{Ackermann:2015zua}.

\item neutrinos from DM annihilation in the Sun.

Pair annihilations of DMs which are captured in the Sun generate neutrinos, and there have been searches for such neutrinos.
In the present scenario, the DM annihilation rate in the Sun is proportional to the following effective SD scattering cross section~\cite{hep-ph/0404175}
\begin{align}
\sigma_p^{\text{SD(eff)}} = \sigma_{p}^{\text{SD}} \tanh^2
\left( \sqrt{\Gamma_{\text{cap.}} \Gamma_{\text{ann.}}}\cdot t_\odot \right)\,,
\end{align}
where $\sigma_p^{\text{SD}}$ is given in Eq.~\eqref{eq:SDXS}, $\Gamma_{\text{cap.}}$ and $\Gamma_{\text{ann.}}$ are the capture and annihilation rates of DM in the Sun, respectively, and $t_\odot$ is the age of the solar system.
If $\sqrt{\Gamma_{\text{cap.}} \Gamma_{\text{ann.}}}\cdot t_\odot\gg 1$, the capture and annihilation rates are in equilibrium.
In the present case, it is given by~\cite{hep-ph/0404175}
\begin{align}
\sqrt{\Gamma_{\text{cap.}} \Gamma_{\text{ann.}}}\cdot t_\odot
\simeq
1.3\times \lrfp{\sigma_p^{\text{SD}}}{10^{-40}\text{cm}^2}{1/2} \lrfp{\vev{\sigma v_{\text{rel}}}_0}{10^{-29}\text{cm}^3/\text{s}}{1/2} \lrfp{50~\text{GeV}}{m_{\chi_1}}{1/4}\,,
\end{align}
and hence the annihilation rate is not completely saturated by the scattering rate.
Currently the Super-Kamiokande gives the strongest bound~\cite{1503.04858} in the mass range of our interest, which are given by $\sigma_p^{\text{SD(eff)}}\cdot \text{Br}(\tau\bar{\tau})\lesssim (1 \text{--} 2)\times 10^{-40}\text{cm}^2$ and $\sigma_p^{\text{SD(eff)}}\cdot \text{Br}(b\bar{b})\lesssim (2 \text{--} 3)\times 10^{-39}\text{cm}^2$ for $m_{\chi_1}\simeq (30 \text{--} 70)~\text{GeV}$.
Here, $\text{Br}(X)=\vev{\sigma v_{\text{rel}}(\chi\chi\to X)}_0/\vev{\sigma v_{\text{rel}}(\chi\chi\to \text{all})}_0$ represent the branching fractions of the annihilation channels, which are given by $\text{Br}(\tau\bar{\tau})\sim m_\tau^2/(3m_b^2+m_\tau^2)$ and $\text{Br}(b\bar{b})\sim 3m_b^2/(3m_b^2+m_\tau^2)$ in the present scenario.
As a result, we found that the bound from the neutrinos are weaker than the bound on $\sigma^{\text{SD}}$ from the XENON~100~\cite{1301.6620} for $m_{\chi_1}\simeq (30 \text{--} 70)~\text{GeV}$ (see Sec.~\ref{subsec:direct}).

\item
mono-jet and mono-photon.
\\
Mono-photon events could be produced at the LEP, via $e^+e^-\to \gamma Z^*\to \gamma \chi_1\chi_1$, but it is expected that the constraint is very weak (see, e.g., Ref.~\cite{hep-ex/0406019}).
The mono-jet events at the LHC, $pp\to j Z^*\to j \chi_1\chi_1$, also gives only a weak constraint on $\lambda^Z$ compared to the other bounds. (See, e.g., a study on generic vector mediator in Ref.~\cite{1407.8257}.)

\end{itemize}

\section{Summary}
\label{sec:discussion}
In this paper, we have investigated the Higgs- and $Z$-resonant neutralino DM scenario.
The phenomenology of this scenario is determined only by three parameters, Bino mass $M_1$, Higgsino mass $\mu$ and $\tan\beta$ when all other SUSY particles and heavy Higgs bosons are decoupled.
In this scenario, the Bino-like neutralino DM can have the correct thermal relic abundance via the Higgs- and $Z$-resonant annihilations.
We have investigated the current constraints and future prospects comprehensively for essentially all the parameter space.

As constraints, we have included: (i) relic abundance $\Omega_\chi h^2 = 0.120$~\cite{Agashe:2014kda,1502.01589}, (ii) direct detection constraints on the spin-independent (SI) scattering cross section from the LUX~\cite{1310.8214} and on the spin-dependent (SD) scattering cross section from the XENON 100~\cite{1301.6620},
(iii)  constraint on the Higgs invisible decay from global fit~\cite{1303.3570,1306.2941}.
For future prospects, we have investigated: (i) prospects of the XENON 1T for the SI~\cite{1206.6288} and SD~\cite{1211.4573} scattering cross sections of DM direct detection, (ii) expected sensitivity of the HL-LHC~\cite{1309.7925,1310.8361} and of the ILC~\cite{1310.0763} for the Higgs invisible decay, and (iii) expected sensitivity of the LHC chargino/neutralino search at 14 TeV for 300 fb$^{-1}$ and 3000 fb$^{-1}$~\cite{ATL-PHYS-PUB-2014-010}.

The results are summarized in Figs.~\ref{fig:main_tb2and3}--\ref{fig:main_tb3040and50}.
It was shown that there is still a large viable parameter space, and almost all the parameter space of the scenario will be covered complementarily by the LHC search, the direct detection experiments, and the Higgs invisible decay search.

In the $Z$-resonant region, the thermal relic abundance leads to an universal upper bound on the chargino mass, $m_{\chi^\pm}\lesssim 450~\text{GeV}$, independently of the $\text{sign}(\mu)$ and $\tan\beta$.
This region will be covered by the chargino/neutralino searches at the LHC at 300 fb$^{-1}$ except for light chargino region $m_{\chi^\pm}\lesssim 200~\text{GeV}$.
For $\mu>0$, almost all the $Z$-resonant region is probed by both of the XENON 1T and the Higgs invisible decay search at the ILC.
For $\mu<0$, due to the blind spot, there are parameter regions which are not covered by the XENON 1T and/or the Higgs invisible decay search at the ILC, depending on $\tan\beta$ and $m_{\chi_1}$.

In the Higgs resonant region, the upper bound on the chargino mass depends on $\text{sign}(\mu)$ and $\tan\beta$.
For $\mu>0$, larger $\tan\beta$ leads to smaller upper bound on $m_{\chi^\pm}$, e.g., $m_{\chi^\pm}\lesssim 2.5~\text{TeV}$ for $\tan\beta=2$ and $m_{\chi^\pm}\lesssim 500~\text{GeV}$ for $\tan\beta=50$.
The XENON 1T will cover almost all the region, and the ILC can probe the Higgs invisible decay for the small $m_{\chi^\pm}$ region.
For $\mu<0$, the allowed region has a nontrivial behavior due to the blind spot.
In both of the two cases $\mu>0$ and $\mu<0$, the LHC at 3000 fb$^{-1}$ can cover the region of $m_{\chi^\pm}\lesssim 800~\text{GeV}$.

It is interesting that, depending on $\tan\beta$, $\text{sign}(\mu)$,  $m_{\chi_1}$,  and $m_{\chi_\pm}$, different combinations of positive and negative signals from different experiments may appear.
It is also encouraging that, in the mass range of Higgs- and $Z$-resonant DM, the direct detection experiments may be able to determine the DM mass within certain uncertainty~\cite{1206.6288,Kavanagh:2013wba,Feldstein:2014gza}.

In this paper, we considered a simplified model where all SUSY multiplets except for the Bino and Higgsino multiplets are decoupled and CP is conserved.
It is interesting to construct a SUSY breaking model to realize such a spectrum and to study the effects of heavier particles and possible CP violation, which are left for future work.

\begin{figure}[t]
\begin{center}
\includegraphics[width=6.5cm]{figures/tb2-mum.eps}
\includegraphics[width=6.5cm]{figures/tb2m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb3-mum.eps}
\includegraphics[width=6.5cm]{figures/tb3m.eps}
\caption{
Constraints and future sensitivity of the Higgs- and $Z$-resonant neutralino DM, for $\tan\beta = 2$ and 3, and for $\mu<0$ (left) and $\mu>0$ (right).
The black lines show abundance $\Omega_{\chi_1}\simeq 0.120$.
The gray shaded region is excluded by current constraints;
the LUX bound on $\sigma^{\text{SI}}_N$ (blue dashed), the XENON100 bound on $\sigma^{\text{SD}}_n$ (green dashed), and the Higgs invisible decay (magenta dashed).
The light yellow region will be probed by future experiments;
the XENON 1T via SI-scattering (blue solid) and SD-scattering (green solid), the Higgs invisible decay at the HL-LHC (magenta dot-dashed) and at the ILC (magenta solid), and the search for the chargino and neutralinos at the 14 TeV LHC, at $3000~\text{fb}^{-1}$ (red solid).
In particular, the light orange region within the red dotted lines will be reached at $300~\text{fb}^{-1}$.
See text for details.
}
\label{fig:main_tb2and3}
\end{center}
\end{figure}

\begin{figure}[t]
\begin{center}
\includegraphics[width=6.5cm]{figures/tb4-mum.eps}
\includegraphics[width=6.5cm]{figures/tb4m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb5-mum.eps}
\includegraphics[width=6.5cm]{figures/tb5m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb6-mum.eps}
\includegraphics[width=6.5cm]{figures/tb6m.eps}
\caption{The same as Fig.~\ref{fig:main_tb2and3}, but for $\tan\beta=4$, 5 and 6.
For $\mu<0$, the blind spot $\lambda^h=0$ is shown with a brown dotted line.}
\label{fig:main_tb45and6}
\end{center}
\end{figure}

\begin{figure}[t]
\begin{center}
\includegraphics[width=6.5cm]{figures/tb7-mum.eps}
\includegraphics[width=6.5cm]{figures/tb7m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb8-mum.eps}
\includegraphics[width=6.5cm]{figures/tb8m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb9-mum.eps}
\includegraphics[width=6.5cm]{figures/tb9m.eps}
\caption{The same as Fig.~\ref{fig:main_tb2and3}, but for $\tan\beta=7$, 8 and 9.}
\label{fig:main_tb78and9}
\end{center}
\end{figure}

\begin{figure}[t]
\begin{center}
\includegraphics[width=6.5cm]{figures/tb10-mum.eps}
\includegraphics[width=6.5cm]{figures/tb10m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb15-mum.eps}
\includegraphics[width=6.5cm]{figures/tb15m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb20-mum.eps}
\includegraphics[width=6.5cm]{figures/tb20m.eps}
\caption{The same as Fig.~\ref{fig:main_tb2and3}, but for $\tan\beta=10$, 15 and 20.}
\label{fig:main_tb1015and20}
\end{center}
\end{figure}

\begin{figure}[t]
\begin{center}
\includegraphics[width=6.5cm]{figures/tb30-mum.eps}
\includegraphics[width=6.5cm]{figures/tb30m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb40-mum.eps}
\includegraphics[width=6.5cm]{figures/tb40m.eps}
\\
\includegraphics[width=6.5cm]{figures/tb50-mum.eps}
\includegraphics[width=6.5cm]{figures/tb50m.eps}
\caption{The same as Fig.~\ref{fig:main_tb2and3}, but for $\tan\beta=30$, 40 and 50.}
\label{fig:main_tb3040and50}
\end{center}
\end{figure}

\section*{Acknowledgement}
The authors are thankful to
Motoi Endo, Jonas Lindert, Takeo Moroi, Mihoko Nojiri, Toshihiko Ota, Satoshi Shirai, Yoshitaro Takaesu, and Takashi Yamanaka for helpful comments and discussions.
K.H. thanks DESY theory group and the Munich Institute for Astro- and Particle Physics (MIAPP), where part of this work was done.
This work was supported by Grant-in-Aid for Scientific research No. 26104001, 26104009, 26247038, 26800123, by World Premier International Research Center Initiative (WPI Initiative), MEXT, Japan, by the MIAPP of the DFG cluster of excellence "Origin and Structure of the Universe", and by the Collaborative Research Center (SFB) 676 at Hamburg University, ``Particles, Strings, and the Early Universe.''

\clearpage

redpajama_set_name: RedPajamaArXiv


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text:
Credit card breach hits 2 E.L. Jimmy John's restaurants
Lindsay VanHulle
lvanhulle@lsj.com
EAST LANSING – Customers who swiped credit or debit cards at two East Lansing Jimmy John's restaurants may have had personal data stolen as part of a security breach, the company said today.
The Champaign, Illinois-based sandwich chain said stores at 4790 S. Hagadorn Road and 143 N. Harrison Road in East Lansing are among 216 nationwide — and 18 in Michigan — that were affected.
Customers who visited the Hagadorn Road store between June 16 and Aug. 2 and the Harrison Road store between June 16 and July 30 were affected.
Jimmy John's said it learned of the breach in July. An investigation found that someone stole login information for a Jimmy John's vendor and used it to hack into cashier systems at multiple stores between June 16 and Sept. 5, the company said.
The company said the breach has been fixed and cards are able to be used at its stores. Only cards swiped at the cash register appear to have been compromised, Jimmy John's said. Online users should not be affected.
Jimmy John's says it will offer identity protection services to affected customers. For that and more information, visit www.jimmyjohns.com.

redpajama_set_name: RedPajamaCommonCrawl


We are proud to say that we have offered this unique and fun league for over 20 years. We hope to keep our youth engaged and participating in our program for many more years to come. Youth In-Line Hockey is offered to boy and girls for those ages 5-18 years old as of September 1st of current school year. This league is offered in the spring and fall, each year. Located at Oatesdale Park, Martinsburg, WV. We offer 3 divisions: Squirt (ages 5-8), Minor (ages 9-12) and Major (ages 13-through High School) and games are typically played on Wednesdays & Sundays. The Adult league is separate from this league.
Registration opens for the fall league in the beginning of July and the spring league opens at the beginning of January. Be sure not to miss the deadlines!
Player Fee & Jersey info- New players must submit a copy of their birth certificate with the application. Player fee is $55.00. In addition, players should have a home and away jersey. New players will need to purchase jerseys for $40. Each player can reuse their jerseys each season so long as they fit and remain in the same division. Each division has a different colored jersey. Payment for Jerseys will need to be received in order to complete your order and before the registration deadline. *Reminder any player needing Jerseys after the registration deadline will be subject to an additional $15 shipping/processing fee and orders submitted will take up to 10-15 business days to complete from date of order placed.
Equipment- The required equipment needed to play: helmet with properly attached face guard, chin guard, chin strap, mouth guard, elbow pads, gloves, shin/knee pads, skates with proper amount of wheels and a stick with a 2″ minimum blade width. Click HERE for complete list of required equipment and stores for used/new equipment.
Anyone interested in coaching or sponsoring, please contact hockey@mbcparks-rec.org / 304-264-4842.
Make-up Games- Cancelations due to weather/rink conditions: Sundays games will be played on the following Saturday at 2pm, 3pm & 4pm. Wednesdays games will be played on the following Monday after at the same time games were scheduled. Please note this is subject to change based on holidays and other circumstances. Always check our website & social media for game cancellations. Parents who are able to volunteer to sweep off the rinks, help vacuum any water etc. are always welcome! Any help is appreciated!
For your spirit wear needs we recommend using Pink Ribbon Designs www.pinkribbondesigns.com.
The Capitals were awarded funding from an Industry Growth Fund last year, and the organization has used those funds to upgrade inline hockey rinks in Hagerstown, Maryland and three locations in Loudoun County, Virginia. Plans were also put in place to upgrade Oatesdale Park in Martinsburg, WV. The Washington Capitals will be renovating our In-Line Hockey rinks at Oatesdale Park in the summer of 2019.
**Please be aware that this renovation is scheduled to begin in summer 2019. Demolition and construction may be taking place prior to and during the current hockey season. Schedules may be delayed, changed and/or cancelled. Thank you for your cooperation and patience. We hope that this renovation helps spark others interest and improve our leagues.


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One of our core values is a commitment to stay Informed 'To keep ahead of the game and always be understanding of latest policy and technical advice'.
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The roads at the centre of debate include State Highway 2, between Te Puna and Omokoroa, and SH2 between Waihi and Omokoroa.
Community concerns about the priority given to a number of significant roading projects in the Bay of Plenty are being heard across regional and national levels.
At a local level community concern has helped to re-prioritise the projects, following public submissions being made to the Bay of Plenty Regional Council Transport Committee.
The result of the submissions means the committee has made improving State Highway 2 between Te Puna and Omokoroa its main priority, and the corridor between Waihi and Omokoroa its third.
Regional Transport Committee chairman, Cr Stuart Crosby says a number of the stories told during the plan's hearing process resonated with the committee and the public submissions helped the committee produce a better balanced, multi-modal document.
"When we released the draft we knew it would evolve and we now believe we have balanced all the transport challenges, needs and priorities across the region," Cr Crosby says.
In addition to these results a petition with close to 3500 signatures has also been delivered to Parliament today by Bay of Plenty MP Todd Muller and Coromandel MP Scott Simpson.
The petition calls on the Government to commit a continuous four-lane highway stretching from Tauranga to Katikati.
"Close to 3500 locals signing a petition gives a fair indication of the huge level of support the project has in our community – and its being put under threat as a direct result of the Government's bad transport policies.
"They expect the Bay to pay more at the pump and see less for it. Our community isn't crying out for a tram in Auckland – what we need is this expressway including a grade separated connection from Omokoroa onto SH2 and the Katikati bypass.
"The Government have also failed to give our community certainty on the TNL. National signed off on it and construction was due to start this year but it has been put on ice until they can work out if it fits with their new transport strategy. It's simply not good enough".
With the traffic congestion from Omokoroa and Tepuke maybe we could utilise the rail corridors. Adapt buses to travel along the rails adding tracks to existing sleepers. This could travel from Apita to Tepuke with stops and stations along the way . Get on at Omokora read the paper on the way in get off at the strand walk to work. Shuttle Buses could feed the stations, also park an ride for those from further a field.
It is disappointing that the new government have scrapped this excellent proposal. It is also a crying shame that WBOPDC did not adequately consult their community, before disrupting the sensible traffic flow in Te Puke, and creating a similar bottleneck to that shown. Some so-called traffic planners should remain at home with their tooys, as they are not welcome in the real world.


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Discussion in 'Sold/Expired' started by murphystrut1, Jun 15, 2018.
I have a Elite 1600 RF for sale. Great Condition. Looking to get $300 shipped in the CONUS. Thanks for looking!
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3 Rising Cybersecurity Issues, According to a Columbia University Expert
From Russian election meddling to North Korean computer warfare, cybersecurity issues have roared into the center of daily life at tremendous speed. At times, these developments can seem overwhelming. So we checked in with an expert to find out what, exactly, we need to know.
Dr. Corey Hirsch is a globally recognized cybersecurity leader. When he's not presenting papers at places like the International Cyber Warfare Conference in Washington, D.C, he serves as Chief Information Security Officer of Teledyne Technologies. He also teaches in the Technology Management and Enterprise Risk Management master's programs at Columbia University's School of Professional Studies. Below, he tells us about crucial new developments in his field.
Collateral Damage from Cyberwarfare
Collateral damage is as old as war itself. But we may not think about the unintended victims of cyberweapons—yet. According to Dr. Hirsch, cyberweapons, which can victimize any device connected to the internet, are even more prone to causing collateral damage than bullets or bombs. That is bad news for their victims. "Unlike casualties in the four 'native' theaters of human conflict (land, sea, air, space), there are no treaties, no conventions, no Red Cross, and barely any meaningful law enforcement mechanisms to protect non-combatants," he says. The NotPetya attack, for example, which Russia launched at Ukrainians, resulted in a $200 million loss for a Danish shipping company. Boeing became collateral damage of the North Korean WannaCry attack nearly one year after it began. "With billions of vulnerable devices connected on the internet, the risks of extreme disruption…are rising exponentially," Dr. Hirsch cautioned. Serious conversations are needed at the organizational, national, and international level to make progress on this issue before we witness further, potentially cataclysmic, damage.
A useful catchphrase from Dr. Hirsch: "Technology leaps, the law creeps." In other words, technology blazes ahead while "regulation tries to keep up—and often makes a bigger mess." The latest example: the European Union recently implemented new data protection laws known as GDPR, some of the strictest in the world. They give consumers more control over how their data is collected and shared, and violations are met with fines of up to four percent of global revenue. (That would be roughly five billion dollars for Google.) According to Dr. Hirsch, the upside is that GDPR "should reduce wasteful and dangerous SPAM, and reduce the toxicity of the 'email information bloodstream.'" The downside: criminal entities will inevitably find a way to exploit new privacy to hide their behavior. "Often, when you go toward anonymity—which is pro-privacy—you're reducing security," said Dr. Hirsch. GDPR will "block clues that cyberdefenders use…so it's harder to discover who may be setting up to attack you." As citizens of the digital age, we all need to consider how we negotiate the trade-offs between privacy and security.
Blockchain technology is an attempt to solve a fundamental problem of the internet: how can we trust people we interact with online? Blockchain is an open-source, online database that "ensures if 51% of the community is trustworthy, you can't cheat." The technology "could play a tremendous positive role in the future of humanity online," said Dr. Hirsch. It has the potential to restore the viability of teamwork—which Dr. Hirsch calls "the core competency of our species"—in the finance, health care, and insurance industries, just to name the most obvious applications. "However," Dr. Hirsch warns, "it has so far been misused." Most people have encountered blockchain only through Bitcoin and other virtual currencies where the blockchain, far from boosting accountability has "massively inflated the threat from ransomware, DDoS, and other extortionate threats." Individuals, companies, and governments should to monitor this emerging trend carefully to ensure it improves—and doesn't plague—global technology.
Dr. Hirsch teaches Enterprise Info Security: Threats and Defense for the Technology Management and Enterprise Risk Management master's programs at Columbia University's School of Professional Studies. The Executive Technology Management program offers a Cybersecuity area of focus.
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Talk Show News: Bigger than Bruce?
OK, so the biggest talk show "get" of the week won't be Bruce Springsteen on "The Daily Show" after all, but President Barack Obama, who will be talking about his economic plan to Jay Leno that same night (Thursday the 19th). Garth Brooks will appear on the same show as the musical guest. It's the president's first sit-down talk show appearance in studio in front of a live audience since becoming elected, according to NBC.


text:
World renowned horsemen head to QSEC
Clinton Anderson and Ian Francis will join forces to host a weekend of innovation, inspiration and motivation for equine enthusiasts
On the 14th and 15th of December at the Queensland State Equestrian Centre in Caboolture, two Australian legends, Clinton Anderson and Ian Francis, will join forces to host a weekend of innovation, inspiration and motivation for equine enthusiasts. Working with several horses to demonstrate their techniques, the weekend's program will include topics such as colt starting techniques, the importance of ground work and gaining a horse's respect, as well as ridden work including exercises and manoeuvres to prepare horses and riders for cattle related events.
The chance to witness these world renowned horsemen working together is a rare opportunity; Clinton lives in America, while Queensland based Ian is regularly on the road hosting horsemanship and cattle working clinics all around Australia. Their deep respect for each other reflects a friendship that spans several decades, and was sparked when Clinton, then 17 years old, spent a year working for Ian. At the time, neither thought that this year would lead to an enduring friendship beyond Mentor and Student. Clinton moved to the USA after his year with Ian, and later wrote "I always knew Ian was a great horseman, but in my youthful ignorance, I thought surely in my quest for knowledge I would find a horseman even more talented than Ian. It wasn't until several years later, with age, experience and travels halfway around the world that I came to realise I would never find anyone better than Ian Francis."
The combination will work with several horses to demonstrate their skills and techniques
Born in Maryborough Queensland, Ian did not come from a family who understood or encouraged his obsession with horses and cattle. His determination to learn and be successful eventually led to an extremely successful career with horses, however it wasn't until he was in his thirties that he became successful and well known. In 2010, after a competitive career that spanned over 30 years, Ian ceased training and showing horses to focus on teaching and his other passion; Brangus cattle. During his career, Ian was a multiple National and State Champion across all Australian Quarter Horse Association disciplines including Reining, Cutting, Reined Cowhorse, Trail, and Western Riding. He has won two Cloncurry Stockans' Challenge Cups, and was awarded the Medal of the Order of Australia (OAM) for Services to the Equine Industry as a trainer, breeder and competitor.
Like Ian, Clinton was not born into a family that shared his drive to succeed in the equine world. He was born and raised in the city, and it was during weekends spent at his grandparents' farm where he learned basic skills from his grandmother aboard her old thoroughbred mare.
Ian Francis showing his talent, technique and style
Since making the United States his home in 1997 and officially launching Downunder Horsemanship and his Downunder Horsemanship Method, Clinton has become recognised as a world leader in the equine industry, helping thousands of horses and riders to form a safe, enjoyable partnership via his step by step program. The Method is based on mutual respect and understanding between horse and rider, and provides people with the knowledge and skills required to train their horses to be safe, consistent and willing partners.
This year, Clinton has reduced the number of tours and clinics he offers in order to concentrate on training and showing his own horses. As such, the opportunity to witness him in Australia sharing his knowledge alongside fellow Australian Legend Ian Francis is both exciting and rare.
Click here to learn more about this event and purchase tickets
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