text:
\section{Introduction}
This paper is part of a series studying properties of hadron resonances within a unified constituent quark model. The model is based on the Bethe-Salpeter-equation in instantaneous approximation and hence is relativistically covariant by construction. The present work extends the investigations of baryon spectra from previous papers on the non-strange \cite{Loring:2001kx} and the strange \cite{Loring:2001ky} sector to charmed baryons. We have found that our Bethe-Salpeter-model is able to describe not only the Regge-trajectories of light flavor baryon spectra but also their hyperfine structures with only seven free parameters. On this basis we have now calculated mass spectra of single, double and triple charm baryons.

A similar extension from light to heavy flavor ha\-drons within the Bethe-Salpeter-model has already been successfully performed for mesons, see \cite{Merten:2001er} for the results.

In the case of charmless baryons we used a three-quark interaction given by a linearly rising confinement potential equipped with a suitable spinoral structure that has been fixed by light flavored baryons \cite{Loring:2001kx}. This interaction remains completely unchanged in the present work.

As we will show this ansatz also yields a satisfactory description of charmed baryons and hence the confining potential is flavor independent and universal.

The residual interaction we use is a phenomenological extension of 't Hooft's instanton-induced force to charm quarks (see \cite{'tHooft:1976fv,Shifman:1979uw} for details on instantons). Phenomenologically here means that the standard derivation is not justified for massive quarks and applicable only for nearly massless quarks. The couplings that enter into this two-body 't Hooft-con\-tri\-bution are for light quarks fixed phenomenologically by the non-strange and the strange bary\-on spectra and are left unchanged. Two additional couplings enter when including charm quarks and are treat\-ed as free parameters. The constituent quark mass of the charm\-ed quark represents another free parameter as well.

These three new parameters are fixed by fitting the experimental single charm baryon spectrum to which we compare our calculations in detail.

Acting between a heavy and a light quark, the one-gluon-exchange is an alternative residual two-quark interaction and is investigated within the same framework. We thus will compare the influence of the one-gluon-exchange on the charmed baryon spectrum to the effects of the extended instanton force.

With these parameters fixed we go a step further and calculate the masses of the lowest double and triple charm baryons and compare the results to the very sparse experimental data. A number of predictions is added.

The paper is organized as follows: Section \ref{sec:BSM} very briefly recapitulates the Bethe-Salpeter-model and shows the ingredients and basic equations. For more details on the theory of the Bethe-Salpeter-model we refer to \cite{Loring:2001kv}. Section \ref{sec:interaction} shortly explains the interactions we use, especially the phenomenological extension of 't Hooft's force to charm quarks. Section \ref{sec:results} summarizes all model parameters and our results for the spectrum of charmed baryons. These results are discussed in detail in comparison to the experimental data. We display the results for the one-gluon-exchange as a residual interaction as well and discuss calculated mass spectra for double and triple charm baryons before concluding in sec.~\ref{sec:conclusion}.

\vspace{20pt}

\section{Bethe-Salpeter-model} \label{sec:BSM}

\subsection{Bound states}
Our formally covariant constituent quark model is based on the Bethe-Salpeter-amplitude $\chi$ defined via quark field operators $\Psi_{a_i}(x_i)$ by
\begin{eqnarray}
\chi_{\bar P\,a^{}_1 a^{}_2 a^{}_3}(x^{}_1,x^{}_2,x^{}_3)
=\langle0|T\Psi^{}_{a^{}_{1}}(x^{}_{1})\Psi^{}_{a^{}_{2}}(x^{}_{2})
\Psi^{}_{a^{}_{3}}(x^{}_{3})|\bar P\rangle \label{eqn:BSA}
\end{eqnarray}
where $\bar P$ denotes the four-momentum of the on-shell bound state and $T$ the time ordering operator. The Bethe-Sal\-pe\-ter-amplitude describes the baryonic bound states and replaces the usual quantum mechanical wave function in non-relativistic quark models. Due to translational invariance we switch from now on to relative coordinates.

The Fourier-transform of the Bethe-Salpeter-amplitude is determined by the Bethe-Salpeter-equation
\begin{eqnarray}
&& \chi_{\bar P\,a_1 a_2 a_3}(p_\xi,p_\eta)=
S^1_{F\,a_1 a'_1}\left(\textstyle \frac{1}{3}\bar P+p_{\xi}+\frac{1}{2}p_{\eta}\right)\, \nonumber \\
&& \,\times S^2_{F\,a_2 a'_2}\left(\textstyle \frac{1}{3}\bar P-p_{\xi}+\frac{1}{2}p_{\eta}\right)\,
S^3_{F\,a_3 a'_3}\left(\textstyle \frac{1}{3}\bar P-p_{\eta}\right)\nonumber \\
&& \,\times\, (-\mathrm{i})\,
\int\frac{\mathrm{d}^4 p_\xi'}{(2\pi)^4}\,\frac{\mathrm{d}^4 p_\eta'}{(2\pi)^4}\,
K^{}_{\bar P\, a'_1 a'_2 a'_3;\, a''_1 a''_2 a''_3}
(p_\xi,p_\eta;\, p_\xi',p_\eta')\nonumber \\
&& \quad\quad\quad\quad\quad\quad\quad\quad\quad\quad\quad\quad\quad\times\chi_{\bar P\, a''_1 a''_2 a''_3}(p_\xi',p_\eta') \label{eqn:BSE}
\end{eqnarray}
where full quark propagators are denoted by $S^i_F$. $K$ stands for the irreducible interaction kernel that contains both irreducible two- and three-quark interaction kernels, $K^{(2)}$ and $K^{(3)}$ respectively. See fig.~\ref{fig:BSE} for a graphical illustration.
\begin{figure}[t]
\begin{center}
\input{BSE.pstex_t}
\caption{Graphical illustration of the three-quark Bethe-Salpeter-equation (\ref{eqn:BSE}). Thick arrows indicate the full quark propagators $S^i_F$. $K$ contains the three-body confinement and the two-body residual interactions.}
\label{fig:BSE}
\end{center}
\end{figure}

\subsection{Approximations}
Solving eq.~(\ref{eqn:BSE}) rigorously is presently not possible: The propagators and the interaction kernels are sums of an infinite number of Feynman-graphs and are unknown functions within QCD. Furthermore, the dependence of the Bethe-Salpeter-equation (\ref{eqn:BSE}) on the relative energies leads to a complicated analytic pole structure.

We therefore replace the full quark propagator $S^i_F$ by the usual fermion propagator
\begin{eqnarray}
S^{i}_F(p_i)=\frac{\mathrm{i}}{\s p_i-m_i+\mathrm{i}\epsilon}
\end{eqnarray}
where $m_i$ denotes the effective constituent quark mass and enters as a free parameter. As a second approximation we choose the interaction kernels to be instantaneous in the rest frame of the baryon with mass $M$. This means that these do not depend on their relative energies, \textit{i.e.}
\begin{eqnarray}
K^{(3)}_{P}(p^{}_{\xi},p^{}_{\eta},p'_{\xi},p'_{\eta})\Bigg|_{P=(M,\vec
0)}=V^{(3)}(\vec p^{}_{\xi},\vec p^{}_{\eta},\vec
p_{\xi}\,\!\!', \vec p_{\eta}\,\!\!')
\end{eqnarray}
and
\begin{eqnarray}
K^{(2)}_{\frac{2}{3}P+p_{\eta_k}}(p^{}_{\xi_k},p'_{\xi_k})\Bigg|_{P=(M,\vec 0)}=V^{(2)}(\vec p^{}_{\xi_k},\vec p_{\xi_k}\,\!\!')
\end{eqnarray}
where we have suppressed the multi-indices. The index $k$ numbers the three possible quark pairs.

\subsection{Salpeter-equation}
When genuine two-quark kernels $K^{(2)}$ are involved, special difficulties arise if one now wants to eliminate the relative energies in eq.~(\ref{eqn:BSE}). Because of the third non-interacting spectator quark there remains a relative energy dependence.

However, we can eliminate the relative energies in eq. (\ref{eqn:BSE}) by introducing the projected Salpeter-amplitude
\begin{eqnarray}
\Phi^{\Lambda}_M(\vec p^{}_\xi,\vec p^{}_\eta)=\Lambda_+(\vec p^{}_\xi,\vec p^{}_\eta)\,\Phi^{}_M(\vec p^{}_\xi,\vec p^{}_\eta)
\end{eqnarray}
where we use the usual Salpeter-amplitude
\begin{eqnarray}
\Phi^{}_M(\vec p^{}_\xi,\vec p^{}_\eta)=\int
\frac{\mathrm{d}p^0_\xi}{2\pi}\frac{\mathrm{d}p^0_\eta}{2\pi}\,\chi^{}_M(p^{}_\xi,p^{}_\eta).
\label{DefSalAmpl}
\end{eqnarray}
and the Salpeter-projector
\begin{eqnarray}
\Lambda_{\pm}(\vec p_\xi,\vec p_\eta)&=&\Lambda^+(\vec
p_1)\otimes\Lambda^+(\vec p_2)\otimes\Lambda^+(\vec
p_3)\nonumber\\
&&\pm\Lambda^-(\vec p_1)\otimes\Lambda^-(\vec
p_2)\otimes\Lambda^-(\vec p_3).
\end{eqnarray}
The projection operator reads
\begin{eqnarray}
\Lambda^\pm(\vec p_{i})=\sum_f\Lambda^\pm_{m_f}(\vec p_{i})\otimes\mathcal{P}^\mathcal{F}_f
\end{eqnarray}
where
\begin{eqnarray}
\Lambda^\pm_{m_f}(\vec p_{i})=\frac{\omega_{m_f}(\vec p_{i})\pm
H_{m_f}(\vec p_i)}{2\omega_{m_f}(\vec p_{i})}
\end{eqnarray}
and $\mathcal{P}^\mathcal{F}_f=|f\rangle\langle f|$ is the flavor projector  that assigns the correct quark masses with $f$ running over all flavors. The one-quark energy is
\begin{eqnarray}
\omega_{m_f}(\vec p_{i})=\sqrt{\vec p_{i}\;\!\!^{2}+m^2_{f}}
\end{eqnarray}
and the Dirac-Hamilton-operator
\begin{eqnarray}
H_{m_f}(\vec p_i)=\gamma^0(\vec\gamma\cdot\vec p_i+m_{f}).
\end{eqnarray}

As shown in detail in \cite{Loring:2001kv}, eq.~(\ref{eqn:BSE}) can now be reduced in Born-approximation to an eigenvalue problem of the form
\begin{equation}
\mathcal{H}\Phi_M^\Lambda = M\,\Phi_M^\Lambda
\label{eqn:SE}
\end{equation}
for the projected Salpeter-amplitude, the eigenvalues being the baryon masses $M$. The Salpeter-Ha\-mil\-tonian $\mathcal{H}$ explicitly reads
\begin{eqnarray}
\label{eqn:SE_Hamiltonian}
&&\left[\mathcal{H}\Phi_M^\Lambda\right]({\bf p_\xi}, {\bf p_\eta})
=
\mathcal{H}_0({\bf p_\xi}, {\bf p_\eta})\;\Phi_M^\Lambda({\bf p_\xi}, {\bf p_\eta})\nonumber\\
&&+\Lambda_{+}(\vec p_\xi,\vec p_\eta)\,\gamma^0\otimes\gamma^0\otimes\gamma^0\;
\int
\frac{\textrm{d}^3 p_\xi'}{(2\pi)^3}\;
\frac{\textrm{d}^3 p_\eta'}{(2\pi)^3}\;\nonumber\\
&&\hspace{80pt} V^{(3)}({\bf p_\xi},{\bf p_\eta};\;{\bf p_\xi'},{\bf
  p_\eta'})\;
\Phi_M^\Lambda({\bf p_\xi'}, {\bf p_\eta'}){\nonumber}\\
&&+\Lambda_{-}(\vec p_\xi,\vec p_\eta)\,\gamma^0\otimes\gamma^0\otimes{\rm 1\kern-.3em I}\;
\int \frac{\textrm{d}^3 p_\xi'}{(2\pi)^3}\;\nonumber\\
&&\hspace{80pt} V^{(2)}({\bf p_\xi},{\bf p_\xi'})\otimes{\rm 1\kern-.3em I}\;
\Phi_M^\Lambda({\bf p_\xi'}, {\bf p_\eta})\\
&&+ \textrm{terms with interacting quark pairs (23) and (31)} \nonumber
\end{eqnarray}
where
\begin{equation}
\mathcal{H}_0=H(\vec p_1)\otimes{\rm 1\kern-.3em I}\otimes{\rm 1\kern-.3em I}+{\rm 1\kern-.3em I}\otimes H(\vec p_2)\otimes{\rm 1\kern-.3em I}+{\rm 1\kern-.3em I}\otimes{\rm 1\kern-.3em I}\otimes H(\vec p_3)
\end{equation}
denotes the free three-quark Hamiltonian with
\begin{eqnarray}
H(\vec p_{i})=\sum_f H_{m_f}(\vec p_{i})\otimes\mathcal{P}^\mathcal{F}_f.
\end{eqnarray}

Equation (\ref{eqn:SE}) is call\-ed the Salpeter-equation and can be solved by standard techniques. This procedure has been already successfully applied to the calculation of light bary\-ons, see \cite{Loring:2001kx,Loring:2001ky,Loring:2001kv}. The spectra describe very well the experimental Regge-trajectories and hyperfine structures with a very limited number of free parameters.

\section{Interactions} \label{sec:interaction}
We will now specify the interaction kernels $V^{(3)}$ and $V^{(2)}$ which enter into eq.~(\ref{eqn:SE_Hamiltonian}). We assume confinement to be a pure three-quark potential; the instanton-induced force and alternatively one-gluon-exchange are given by two-quark residual interactions.

\subsection{Confinement}
Quarks have the property to be asymptotically free at short distances but to interact strongly for large distances in the low and intermediate energy region. It is still not possible to derive their confinement analytically from QCD but it is clearly visible in lattice calculations. We thus parameterize the confinement potential phenomenologically. As shown in \cite{Loring:2001kx,Loring:2001ky} the instantaneous ansatz
\begin{eqnarray}
&& V^{(3)}(x^{}_1,x^{}_2,x^{}_3;x'_1,x'_2,x'_3)
 =V^{(3)}_{\mathrm{conf}}(\vec x^{}_1,\vec x^{}_2,\vec
x^{}_3)\delta(x^0_1-x^0_2) \nonumber \\
&& \times\delta(x^0_2-x^0_3) \delta^{(4)}(x^{}_1-x'_1)\delta^{(4)}(x^{}_2-x'_2)\delta^{(4)}(x^{}_3-x'_3)
\end{eqnarray}
given in coordinate space with
\begin{eqnarray}
&&V^{(3)}_{\mathrm{conf}}(\vec x^{}_1,\vec x^{}_2,\vec x^{}_3)
=3a\cdot\frac{1}{4}(\mbox{{\small 1}\hspace{-0.37em}1}\otimes\mbox{{\small 1}\hspace{-0.37em}1}\otimes\mbox{{\small 1}\hspace{-0.37em}1}+
\gamma^0\otimes\gamma^0\otimes\mbox{{\small 1}\hspace{-0.37em}1} \nonumber\\ &&\hspace{10pt}+\textrm{cycl.~perm.})
+b\sum_{i<j}|\vec x_i-\vec x_j|\cdot\frac{1}{2}(-\mbox{{\small 1}\hspace{-0.37em}1}\otimes\mbox{{\small 1}\hspace{-0.37em}1}\otimes\mbox{{\small 1}\hspace{-0.37em}1}
\nonumber \\ &&\hspace{85pt}+\gamma^0\otimes\gamma^0\otimes\mbox{{\small 1}\hspace{-0.37em}1}+\textrm{cycl.~perm.})
\end{eqnarray}
leads to satisfactory results when describing light flavor baryon spectra. The offset $a$ and the slope $b$ enter as free parameters and are fixed by the light flavor baryon spectrum. We do not change these confinement parameters and thus assume that confinement is indeed universal and flavor independent.

\subsection{'t Hooft's potential}
Instantons are special solutions of the classical Euclidean QCD Yang-Mills-equations. They fall into topologically distinct homotopy classes and can in Minkowski-space be interpreted as tunneling events between distinct vacua.

As shown for the first time in \cite{'tHooft:1976fv} instantons can contribute to quark interactions. In particular, instantons lead to an effective two-quark interaction which for quark pairs in baryons reads
\begin{eqnarray}
&&V^{(2)}(x^{}_1,x^{}_2;x'_1,x'_2)=V^{(2)}_{\mbox{\scriptsize
't Hooft}} (\vec x^{}_1-\vec x^{}_2)
\delta(x^0_1-x^0_2)\nonumber\\
&&\hspace{95pt}\delta^{(4)}(x^{}_1-x'_1)\delta^{(4)}(x^{}_2-x'_2)
\end{eqnarray}
with
\begin{eqnarray}
&&V^{(2)}_{\mbox{\scriptsize 't Hooft}}(\vec x)
=-4v_{\mathrm{reg}}(\vec x)
\times(\mbox{{\small 1}\hspace{-0.37em}1}\otimes\mbox{{\small 1}\hspace{-0.37em}1}+\gamma^5\otimes\gamma^5)
\mathcal{P}^{\mathcal{D}}_{S_{12}=0}\nonumber\\&&
\otimes(g_{nn}\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(nn)
+g_{ns}\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(ns)
+g_{nc}\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(nc)
+g_{sc}\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(sc))\nonumber\\
\label{eqn:instantonact}
\end{eqnarray}
where $\mathcal{P}^{\mathcal{D}}_{S_{12}=0}$ is the projector onto spin-singlet states and $\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(f_1f_2)$ is the projector onto flavor-antisymmetric quark pairs with flavors $f_1$ and $f_2$. The couplings $g_{f_1f_2}$ are given by integrals over instanton densities which are basically unknown. We treat them here as free parameters.

Note that we have ad hoc extended 't Hooft's force to charmed quarks by introducing two additional couplings $g_{nc}$ and $g_{sc}$ and flavor projectors $\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(nc)$ and $\mathcal{P}^{\mathcal{F}}_{\mathcal{A}}(sc)$ that operate exactly like the terms for the light-flavored quarks.

Due to the non-perturbative approach we also had to regularize 't Hooft's potential which is essentially a contact interaction that we have replaced by a Gaussian function of the form
\begin{eqnarray}
v_{\mathrm{reg}}(\vec x)=\frac{1}{\lambda^3\pi^{\frac{3}{2}}} e^{-\frac{|\vec x|^2}{\lambda^2}}.
\end{eqnarray}
The effective range parameter $\lambda$ enters as an additional free parameter assumed to be flavor independent.

\subsection{One-gluon-exchange}
For comparison we investigate also the one-gluon-exchange as a residual interaction by assuming a gluon propagator
\begin{eqnarray}
\gamma^\mu D_{\mu\nu}\gamma^\nu
=
4\pi
\left(\frac{\gamma^0\otimes\gamma^0}{|{\vec q|^2 }}
+
\frac{
\boldsymbol{\gamma}\cdot\!\otimes\;\boldsymbol{\gamma}
-
(\boldsymbol{\gamma}\cdot {\vec{\hat q}})\otimes(\boldsymbol{\gamma}\cdot {\vec{\hat q}})}
{q^2+\mathrm{i}\epsilon}
\right) \nonumber\\
\end{eqnarray}
in Coulomb-gauge with ${\vec{\hat q}}:= {\vec q}/|{\vec q}|$, see \cite{Murota:1982in}. The component $D_{00}(q)$, which describes the Coulomb-potential, is already instantaneous and we make the remaining components also instantaneous by setting $q^2$ equal to $-|\vec{q}|^2$. The two-quark one-gluon-exchange interaction then reads in coordinate space
\begin{eqnarray}
&&V^{(2)}(x^{}_1,x^{}_2;x'_1,x'_2)=V^{(2)}_{\mbox{\scriptsize
OGE}} (\vec x^{}_1-\vec x^{}_2)
\delta(x^0_1-x^0_2)\nonumber\\
&&\hspace{95pt}\delta^{(4)}(x^{}_1-x'_1)\delta^{(4)}(x^{}_2-x'_2)
\end{eqnarray}
with
\begin{eqnarray}
&&V^{(2)}_{\mathrm{OGE}}({\bf x})
=-\frac{2}{3}\frac{\alpha_s}{|{\bf x}|}
\Big(\gamma^0\otimes\gamma^0 -
\frac{1}{2}\boldsymbol{\gamma}\cdot\!\otimes\;\boldsymbol{\gamma}
\nonumber\\&& \hspace{100pt}
- \frac{1}{2}(\boldsymbol{\gamma}\cdot {\bf \hat x})\otimes(\boldsymbol{\gamma}\cdot {\bf \hat x})
\Big) \label{eqn:OGEpot}
\end{eqnarray}
where ${\bf \hat x}:= {\bf x}/|{\bf x}|$, see again \cite{Murota:1982in}. We treat the strong coupling constant $\alpha_s$ as a fit parameter.

\section{Results and discussion} \label{sec:results}

\subsection{General remarks}
The baryons we consider can be classified into flavor SU(4) multiplets. The subsets of defined symmetry under permutation of the three quarks are obtained as irreducible representations of the threefold product of the fundamental representation of SU(4), \textit{i.e.}
\begin{eqnarray}
4\otimes 4\otimes 4=20_{\mathrm{S}}\oplus 20_{\mathrm{M}_\mathrm{S}}\oplus 20_{\mathrm{M}_\mathrm{A}}\oplus\bar{4}_{\mathrm{A}}.
\end{eqnarray}
The indices stand for total and mixed symmetry and anti-symmetry. We are mainly interested in single charm bary\-ons which reduces the flavor space to a symmetric 6-plet containing $\Sigma_c$, $\Xi_c$ and $\Omega_c$, a mixed 6-plet containing $\Sigma_c$, $\Xi'_c$ and $\Omega_c$, a mixed and an anti-symmetric $\bar 3$-plet containing both $\Lambda_c$ and $\Xi_c$. Note that we have exact SU(2)-isospin symmetry in the Bethe-Salpeter-model.

The physical ground states $\Xi'_c\frac{1}{2}^+$ and $\Xi_c\frac{1}{2}^+$ are linear combinations of states belonging to mixed multiplets. This will be further investigated in the discussion.

The Particle Data Group Collaboration lists fourteen detected charmed baryons \cite{Eidelman:2004wy}, thirteen single and only one double charm state. No spin or parity quantum numbers have actually been measured yet but are plausible quark model assignments. There are no quantum numbers known for the one two- and the two one-star states, among them the double charm candidate.

The only ground state of the single charm multiplets that has not been detected yet is the $\Omega_c\frac{3}{2}^+$ belonging to the symmetric 6-plet.

\renewcommand{\arraystretch}{1.25}

\subsection{Parameters}
Before we present the results for the spectra of charmed baryons calculated within the Bethe-Salpeter-model, we show in table~\ref{tab:parameters} the explicit numbers of the free parameters that enter into the calculations.
\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|c||c|c|}
\hline
$m_n$ & 330\,MeV & $g_{nn}$ &136.0\,MeV\,fm$^3$ \\
$m_s$ & 670\,MeV & $g_{ns}$ &94.0\,MeV\,fm$^3$  \\
$m_c$ & 1950\,MeV& $g_{nc}$ & 33.3\,MeV\,fm$^3$\\
$a$     & -744\,MeV& $g_{sc}$ & 11.0\,MeV\,fm$^3$\\
$b$ & 470\,MeV\,fm$^{-1}$&$\lambda$&0.4 fm\\
\hline
\end{tabular}
\end{center}
\caption{All ten free parameters of the Bethe-Salpeter-model. In the left columns are the constituent quark masses and the confinement parameters. On the right are the couplings and the effective range of the 't Hooft-interaction.}
\label{tab:parameters}
\end{table}
\end{center}
We have added three additional free parameters, namely the constituent quark mass $m_c$ and the couplings $g_{nc}$ and $g_{sc}$, to the seven parameters already fixed by the light-flavored baryon spectrum.

There is no contribution of 't Hooft's force for flavor symmetric states, such as $\Delta$-resonances. So the off-set parameter $a$ and the slope $b$ of the confinement potential and the non-strange quark mass $m_n$ have been fixed by the well measured positive-parity $\Delta$-Regge trajectory alone. The strange quark mass $m_s$ is fitted to the decuplet hyperons that are not affected by the 't Hooft-potential either. The coupling $g_{nn}$ and the effective range $\lambda$ are fixed by the $\Delta-N$ mass splitting whereas the coupling $g_{ns}$ is determined by the experimental hyperfine splitting between octet and decuplet hyperon ground states.

The new free parameters $m_c$, $g_{nc}$ and $g_{sc}$ are simultaneously fitted to the eleven experimentally known three- and four-star single charm baryons listed in \cite{Eidelman:2004wy}.

\subsection{Spectrum of charmed baryons} \label{sec:Spectra}
Table \ref{tab:singlycharmedmasses} compares the numbers of all thirteen known experimental masses to our computed values for single charm baryonic states using 't Hooft's force between all possible flavor combinations as given in eq.~(\ref{eqn:instantonact}).
\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|c|c|c|c|}
\hline
State&Rating&$J^\pi$&EXP&BSM\\
\hline
$\Lambda_c$      &****&$1/2^+$&$2285\pm 1$&2272\\
$\Lambda_c(2593)$&*** &$1/2^-$&$2594\pm 1$&2594\\
$\Lambda_c(2625)$&*** &$3/2^-$&$2627\pm 1$&2586\\
$\Lambda_c?\Sigma_c?(2765)$&*&$?^?$&$2765\pm 3$&2769\\
$\Lambda_c?\Sigma_c?(2880)$&**&$?^?$&$2881\pm 2$&2874\\
$\Sigma_c(2455)$ &****&$1/2^+$&$2452\pm 1$&2459\\
$\Sigma_c(2520)$ &*** &$3/2^+$&$2518\pm 2$&2539\\
$\Xi_c$          &*** &$1/2^+$&$2469\pm 1$&2469\\
$\Xi'_c$         &*** &$1/2^+$&$2577\pm 3$&2595\\
$\Xi_c(2645)$    &*** &$3/2^+$&$2646\pm 2$&2651\\
$\Xi_c(2790)$    &*** &$1/2^-$&$2790\pm 4$&2769\\
$\Xi_c(2815)$    &*** &$3/2^-$&$2816\pm 2$&2771\\
$\Omega_c$       &*** &$1/2^+$&$2698\pm 3$&2688\\
$\Omega_c$       &    &$3/2^+$&    &2721\\
\hline
\end{tabular}
\end{center}
\caption{Ratings, the experimental (EXP) positions with their error bars (in MeV) together with the quantum numbers total angular momentum $J$ and parity $\pi$ of all known single charm baryons \cite{Eidelman:2004wy}. These masses are compared to the values calculated within the Bethe-Salpeter-model (BSM) using an appropriate three-quark confinement potential and 't Hooft's force between all possible flavor combinations. For the one- and two-star resonances, for which the quantum numbers are not known, we have performed an assignment based on the spectrum given in fig.~\ref{fig:charmed}. We also predict the mass for the $\Omega_c\frac{3}{2}^+$ which has not been measured yet but completes the symmetric 6-plet.}
\label{tab:singlycharmedmasses}
\end{table}
\end{center}
Figure \ref{fig:charmed} displays not only the direct comparison between theoretical and known experimental resonances but also shows many states that we predict which have not been observed so far. We restrict ourselves to the energy region between 2200 and 3120\,MeV and to total spins not exceeding $\frac{3}{2}$.
\begin{figure*}[t]
\begin{center}
\input{Charm_tHooft.pstex_t}
\caption{The single charm baryon spectrum calculated in the Bethe-Salpeter-model using an appropriate three-quark confinement potential and 't Hooft's force between all possible flavor combinations (on the left side of each column) in comparison to the experimental masses from \cite{Eidelman:2004wy} (on the right side where the status is indicated by stars, the mass uncertainty by a shaded box and the lack of knowledge on quantum numbers by a question mark) for various total angular momentum $J$ and parity $\pi$ assignments given by our model.}
\label{fig:charmed}
\end{center}
\end{figure*}

Comparing the gross structure of the experimental and theoretical spectrum we find a very good overall agreement. We have a very clear one-to-one correlation between experimental and theoretical states: There is no experimental state that we do not predict and there is no low lying theoretical resonance that is not observed.

We assign the two-star resonance to the first excited $\Lambda_c\frac{3}{2}^-$-state and the one-star resonance to the first excited $\Lambda_c\frac{1}{2}^+$-state. These assignments are not mandatory. The two-star resonance could have the spin-parity combinations $\frac{3}{2}^+$ or $\frac{1}{2}^-$ just as well. The one-star resonance could also be the lowest lying $\Sigma_c\frac{1}{2}^-$-state.

The discrepancies between the experimental and theoretical masses are relatively small. Eleven of the thirteen predicted states have a relative mass deviation less than one percent. The remaining three baryons have a mass deviation that is approximately one and a half percent.

We underestimate the $\Lambda_c$ by 13\,MeV which is only 0.6\% of its total mass whereas the other four-star state $\Sigma_c(2455)$ is overestimated by 7\,MeV which is only 0.3\% of the total mass.

Both positions of the three-star resonances $\Lambda_c(2593)$ and $\Xi_c$ are exactly reproduced and the masses of the three-star states $\Sigma_c(2520)$, $\Xi_c'$, $\Xi_c(2645)$ and $\Omega_c$ differ only by 0.8\%, 0.7\%, 0.2\% and $-0.4$\% respectively. The first excited $\Omega_c\frac{1}{2}^+$-state is located at 3169\,MeV.

Not accounted for are the experimental splittings of 33\,MeV and 26\,MeV between the spin-parity $\frac{1}{2}^-$ and $\frac{3}{2}^-$ pairs $\Lambda_c(2593)$, $\Lambda_c(2625)$ and $\Xi_c(2790)$, $\Xi_c(2815)$. Where\-as these pairs each are almost degenerate in the Bethe-Sal\-pe\-ter-model they seem to be clearly separated experimentally. The theoretical splitting between the two $\Lambda_c$ resonances even has the wrong sign. This can also be found for the negative parity sector in the light flavor baryon spectra of the $\Delta$, $N$ and $\Lambda$ resonances, see \cite{Loring:2001kx,Loring:2001ky}. Consequently, we underestimate both masses of the $\frac{3}{2}^-$-states $\Lambda_c(2625)$ and $\Xi_c(2815)$ by 1.6\% which is still small but distinctly more than for all the other states. The theoretical splitting between the two $\Xi_c$ resonances has at least the correct sign but the theoretical mass for $\Xi_c(2790)$ is already too small by 0.8\%.

The one- and two-star resonances are again very well described and differ only by 4\,MeV and $-7$\,MeV when compared to the experimental values.

In table \ref{tab:singlycharmedmasses} we have also given our prediction for the $\Omega_c\frac{3}{2}^+$ that is the only baryon of the symmetric 6-plet that has not been detected in experiment yet.

\subsection{Instanton effects between light quarks} \label{sec:instlight}

We shall now study in detail the effects of the 't Hooft-interaction between light quarks and a light and a charm quark by switching off the 't Hooft-couplings $g_{nn}$, $g_{ns}$, $g_{nc}$ and $g_{sc}$ and increasing gradually their strengths one after another to their final values given in table \ref{tab:parameters}. It has been shown in \cite{Loring:2001kx,Loring:2001ky} that instanton-induced effects indeed dominate the fine structure of the whole light-flavored baryon spectrum. We will show that this is also the case for charmed baryons.

\begin{figure*}[p]
\begin{center}
\input{Instanton01.pstex_t}
\caption{The effects of the instanton-induced interaction between the light quarks on the theoretical single charm baryon resonances $\Lambda_c$ and $\Sigma_c$ with total angular momentum $J$ and parity $\pi$. The spectrum on the left in each column is determined by the confinement force alone. The following curves illustrate how the spectrum changes with increasing 't Hooft-coupling $g_{nn}$. The final value is shown in comparison with the experimental value shown on the right in each column.}
\label{fig:charmed_tHooft01}
\end{center}
\end{figure*}

\begin{figure*}[p]
\begin{center}
\input{Instanton02.pstex_t}
\caption{The effects of the instanton-induced interaction between the light quarks on the resonances $\Xi_c$. The spectrum on the left in each column is determined by the confinement force alone. The curves illustrate how the spectrum changes with increasing 't Hooft-coupling $g_{ns}$. The final value is shown in comparison with the experimental value.}
\label{fig:charmed_tHooft02}
\end{center}
\end{figure*}

Figure \ref{fig:charmed_tHooft01} shows on the left in each column the theoretical single charm baryon spectrum of the $\Lambda_c$ and $\Sigma_c$ resonances determined by the confinement interaction alone which means that $g_{nn}=g_{nc}=0$. Due to their quark contents $nnc$ there never is a contribution from the coupling parameterized by $g_{ns}$.

We see that $\Sigma_c(2455)$, $\Sigma_c(2520)$ and $\Lambda_c(2880)$ are actually already well described by the confinement potential alone. It is true that the four-star state $\Sigma_c(2455)$ is with 87\,MeV clearly overestimated but we are now just interested in the gross picture.

\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|rrrr|rr|}
\hline
Mass &
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{S}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^4 \bar 3_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{A}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{A}}[\mathrm{M}]\!\!$&
$\!\!{}^4 \bar 3_{\mathrm{A}}[\mathrm{A}]\!\!$\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^+$}\\
\hline
     2272 & {\bf \underline{96.9}} &    1.9 &    0.0 &    0.1 &    0.9 &    0.1\\
     2769 & {\bf \underline{75.4}} &   11.7 &    0.1 &    0.2 &    12.1 &    0.4\\
     2935 &   20.6 &   21.3 &    0.0 &    1.6 & {\bf \underline{52.9}} &    3.4\\
     3057 &    1.2 & {\bf \underline{61.1}} &    3.7 &    1.0 &   32.6 &    0.1\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^+$}\\
\hline
     2848 & {\bf \underline{53.2}} &   24.6 &    0.3 &    0.1 &   21.6 &    0.2\\
     3100 &    2.2 &   24.4 &   18.7 &    0.9 & {\bf \underline{50.7}} &    3.0\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^-$}\\
\hline
     2594 &    3.1 &   44.4 &    0.3 &    0.0 & {\bf \underline{52.0}} &    0.0\\
     2853 &    1.7 &   43.8 &    8.0 &    2.2 & {\bf \underline{44.2}} &    0.1\\
     2950 &    0.2 &    7.0 & {\bf \underline{85.4}} &    1.0 &    2.2 &    4.3\\
     3018 &   12.2 &   40.7 &    2.4 &    0.0 & {\bf \underline{44.3}} &    0.1\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^-$}\\
\hline
     2586 &    3.0 &   42.3 &    0.3 &    0.0 & {\bf \underline{54.3}} &    0.1\\
     2874 &    0.1 &   31.9 & {\bf \underline{42.1}} &    1.3 &   23.4 &    1.2\\
     2927 &    1.3 &   23.2 & {\bf \underline{53.9}} &    1.0 &   19.1 &    1.5\\
     3009 &   11.1 &   36.7 &    0.6 &    0.0 & {\bf \underline{51.3}} &    0.0\\
\hline
\end{tabular}
\end{center}
\caption{Masses (in MeV) and SU(8) configuration mixings (in \%) for the $\Lambda_c$ ground state and excited states with total angular momentum $J$ and parity $\pi$. The configuration is denoted by $^{2S+1}\mathrm{Flav}[\mathrm{Sym}]$ where $S$ is the total intrinsic spin of the quarks, Flav is the corresponding flavor multiplet and Sym the symmetry of the total spin-flavor SU(8) configuration. Dominant contributions are bold printed and underlined. We have omitted the generally very small negative energy contributions.}
\label{tab:su8confLamC}
\end{table}
\end{center}

Keeping the coupling $g_{nc}$ at zero, we increase the coupling $g_{nn}$ switching on the 't Hooft-interaction between the light quarks. We observe that in particular the lowest lying states $\Lambda_c\frac{1}{2}^+$, $\Lambda_c\frac{1}{2}^-$, $\Lambda_c\frac{3}{2}^-$ and the second lowest lying state $\Lambda_c\frac{1}{2}^+$ are very strongly influenced by 't Hooft's force. They are each lowered by approximately 230\,MeV. This can be understood when we analyze their spin-flavor SU(8) configurations listed in table \ref{tab:su8confLamC}. The aforesaid states are dominantly $^2 \bar 3_{\mathrm{M}}[\mathrm{S}]$ or $^2 \bar 3_{\mathrm{A}}[\mathrm{M}]$ configurations. The flavor states of the mixed anti-triplet $\bar 3_{\mathrm{M}}$ are anti-symmetric under the interchange of the two light quarks and the flavor states of the anti-symmetric anti-triplet $\bar 3_{\mathrm{A}}$ are anti-symmetric anyway. The total intrinsic spin of these states is not $\frac{3}{2}$ and thus the spin part not totally symmetric. Such spin-flavor combinations allow 't Hooft's force (which projects onto anti-symmetric spin and flavor pairs) to act unhindered which in fact it does.

\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|rrrr|rr|}
\hline
Mass &
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^4      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{A}]\!\!$&
$\!\!{}^4      6_{\mathrm{S}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{S}}[\mathrm{M}]\!\!$\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^+$}\\
\hline
     2459 & {\bf \underline{95.5}} &    3.1 &    0.0 &    0.0 &    0.0 &    1.3\\
     2947 & {\bf \underline{88.9}} &    6.3 &    0.0 &    0.1 &    0.0 &    4.5\\
     3048 &    0.0 &    0.1 &   36.1 &    0.8 & {\bf \underline{62.6}} &    0.1\\
     3088 &    1.6 & {\bf \underline{64.3}} &    0.0 &    0.1 &    0.2 &   33.5\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^+$}\\
\hline
     2539 &    0.0 &    0.0 &    3.0 &    0.0 & {\bf \underline{96.9}} &    0.0\\
     3010 &    0.4 &    0.1 &   19.3 &    0.5 & {\bf \underline{79.4}} &    0.2\\
     3031 & {\bf \underline{66.1}} &    4.3 &    1.2 &    0.7 &    1.9 &   25.7\\
     3049 &    1.3 &    4.3 &   35.9 &    1.4 & {\bf \underline{54.2}} &    2.7\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^-$}\\
\hline
     2769 &    3.4 & {\bf \underline{71.6}} &    2.7 &    0.2 &    0.0 &   22.1\\
     2817 &    0.1 &    3.6 & {\bf \underline{94.2}} &    0.0 &    1.9 &    0.1\\
     2929 &    0.1 &   20.5 &    1.2 &    2.4 &    0.0 & {\bf \underline{75.8}}\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^-$}\\
\hline
     2799 &    1.1 &   10.3 & {\bf \underline{59.0}} &    0.1 &    1.2 &   28.2\\
     2815 &    2.2 & {\bf \underline{50.1}} &   31.9 &    0.1 &    0.6 &   14.9\\
     2919 &    0.2 &   35.7 &    7.1 &    2.5 &    0.1 & {\bf \underline{54.2}}\\
\hline
\end{tabular}
\end{center}
\caption{Masses (in MeV) and SU(8) configuration mixings (in \%) for the $\Sigma_c$ ground state and excited states with total angular momentum $J$ and parity $\pi$.}
\label{tab:su8confSigC}
\end{table}
\end{center}

The opposite is true for the spin-flavor SU(8) configurations of the $\Sigma_c$ in table \ref{tab:su8confSigC}. All states belong to flavor 6-plets. The flavor states of the mixed sextet $6_{\mathrm{M}}$ are symmetric under the interchange of the two light quarks and the flavor states of the symmetric sextet $6_{\mathrm{S}}$ are symmetric anyway. This forbids 't Hooft's force to act between the $nn$-quark pairs which can also be seen in fig.~\ref{fig:charmed_tHooft01}.

Using the confinement potential alone, the calculated ground states $\Lambda_c\frac{1}{2}^+$ and $\Sigma_c\frac{1}{2}^+$ are degenerate. But the 't Hooft-interaction acting only on the $\Lambda_c\frac{1}{2}^+$ leads to their experimentally observed splitting, including the correct sign. This splitting amounts to 167\,MeV in experiment and is 239\,MeV when we only take into account 't Hooft's force acting between the light quarks. The latter number will be modified in the correct direction when we also include the 't Hooft-interaction between a light and the charm quark.

Looking again at fig.~\ref{fig:charmed_tHooft01} and the remaining states, we see that the experimental situation for the $\Lambda_c$ and $\Sigma_c$ resonances is already nicely explained by the confinement and the 't Hooft-potential between light quarks alone: There is a clear one-to-one correlation between experimental and theoretical states. Problematic remain the splittings between the lowest lying pairs $\Lambda_c\frac{1}{2}^-$, $\Lambda_c\frac{3}{2}^-$ and $\Sigma_c\frac{1}{2}^+$, $\Sigma_c\frac{3}{2}^+$. We note that the lowering of the former pair states is almost identically strong which spoils a direct explanation for their splitting. This is no surprise since they possess almost identical spin-flavor configurations, see table \ref{tab:su8confLamC}, but the lowering is too strong for the $\frac{3}{2}^-$-state. Nevertheless, we will see below that the splitting between the second pair can be impressively explained with the extension of 't Hooft's force to charm quarks.

Figure \ref{fig:charmed_tHooft02} shows on the left in each column the theoretical single charm baryon spectrum of the $\Xi_c$ resonances again determined by the confinement interaction alone which means that $g_{ns}=g_{nc}=g_{sc}=0$. Due to the quark flavor content $nsc$ there never is a contribution from the coupling parameterized by $g_{nn}$.

\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|rrrr|rr|}
\hline
Mass &
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^4      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{A}]\!\!$&
$\!\!{}^4      6_{\mathrm{S}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{S}}[\mathrm{M}]\!\!$\\
 &
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{S}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^4 \bar 3_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{M}}[\mathrm{A}]\!\!$&
$\!\!{}^4 \bar 3_{\mathrm{A}}[\mathrm{S}]\!\!$&
$\!\!{}^2 \bar 3_{\mathrm{A}}[\mathrm{M}]\!\!$\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^+$}\\
\hline
     2469 & 1.3 &    0.2 &    0.0 &    0.0 &    0.0 &    0.5\\
          & {\bf \underline{96.9}}&    0.6 &    0.0 &    0.0 &    0.0 &    0.3\\
     2595 & {\bf \underline{94.2}}&    2.5 &    0.0 &    0.0 &    0.0 &    1.0\\
          & 1.2 &    0.5 &    0.0 &    0.0 &    0.0 &    0.5\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^+$}\\
\hline
     2651 & 0.0 &    0.0 &    2.7 &    0.0 &   {\bf \underline{96.0}} &    0.0\\
          & 0.0 &    0.0 &    1.3 &    0.0 &    0.0 &    0.0\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^-$}\\
\hline
     2769 & 0.8 &   40.5 &    0.1 &    1.0 &    0.0 &    0.5\\
          & 0.0 &    0.0 &    0.0 &    0.0 &    0.0 &   {\bf \underline{57.1}}\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^-$}\\
\hline
     2771 & 0.7 &   44.5 &    0.2 &    0.9 &    0.0 &    0.4\\
          & 0.0 &    0.0 &    0.0 &    0.0 &    0.0 &   {\bf \underline{53.2}}\\
\hline
\end{tabular}
\end{center}
\caption{Masses (in MeV) and SU(8) configuration mixings (in \%) for the $\Xi_c$ ground state and excited states with total angular momentum $J$ and parity $\pi$.}
\label{tab:su8confXiC}
\end{table}
\end{center}

We see that only the $\Xi_c(2645)$ is really well described by the confinement potential alone. But increasing just the coupling $g_{ns}$ and keeping the other couplings $g_{nc}$ and $g_{sc}$ at zero, we observe that the two lowest lying degenerate states for the spin-parity combination $\frac{1}{2}^+$ separate creating the experimentally observed splitting between the $\Xi_c$ and the $\Xi'_c$. This behavior can be explained when we investigate their spin-flavor SU(8) configurations listed in table \ref{tab:su8confXiC}: They are almost entirely $^2 \bar 3_{\mathrm{M}}[\mathrm{S}]$ and $^2 6_{\mathrm{M}}[\mathrm{S}]$ configurations respectively. Like before 't Hooft's force acts on the flavor-$\bar 3$-plet but not on the flavor-6-plet members, so it lowers the $\Xi_c$ and leaves the $\Xi_c'$ virtually untouched. Their splitting is 108\,MeV experimentally and 159\,MeV when we consider the 't Hooft-interaction between the light quarks only. This theoretical number will also be improved when we include 't Hooft's interaction between all flavors.

We find the $\Xi(2645)$ as an almost pure $^4 6_{\mathrm{S}}[\mathrm{S}]$ configuration and thus a state entirely consisting of intrinsic spin $\frac{3}{2}$, see table \ref{tab:su8confXiC}. The 't Hooft-interaction will not influence this state, regardless between which quark flavors it acts.

The negative parity states $\Xi_c\frac{1}{2}^-$ and $\Xi_c\frac{3}{2}^-$ have practically the same configuration mixings and our residual interaction thus affects both states in nearly the same way, namely each state is lowered by 133\,MeV. But, just like for the $\Lambda_c\frac{3}{2}^-$, the lowering of the $\Xi_c\frac{3}{2}^-$ resonance is too strong to yield the correct experimental splitting between $\Xi_c(2790)$ and $\Xi_c(2815)$.

To sum up the situation for the $\Xi_c$ in fig.~\ref{fig:charmed_tHooft02}, it can be said that the gross experimental structure is already well described by the confinement potential and 't Hooft's residual interaction operating only between light quarks. However, the masses of the the important states $\Xi_c$ and $\Xi'_c$ are too large which will be cured below when including phenomenologically 't Hooft's force also operating between a light and a charm quark.

We finally notice that for the $\Omega_c$ there is no 't Hooft-potential between two light quarks.

\subsection{Instanton effects between a light and a charm quark}

Based on the findings of the preceding section, we will now investigate the effects of the instanton-induced interaction when operating also between a light and the charm quark.

Figure \ref{fig:charmed_tHooft1} shows on the left in each column the theoretical single charm baryon spectrum of the $\Lambda_c$ and $\Sigma_c$ resonances determined by the confinement interaction and the residual interaction operating only between light quarks, \textit{i.e.}~for $g_{nn}=136\,\mathrm{MeV\,fm}^3$ and $g_{nc}=0$.

\begin{figure*}[p]
\begin{center}
\input{Instanton1.pstex_t}
\caption{The effects of the instanton-induced interaction between a light and the charm quark on the theoretical single charm baryon resonances $\Lambda_c$ and $\Sigma_c$ with total angular momentum $J$ and parity $\pi$. The spectrum on the left in each column is determined by the confinement and the residual interaction operating only between the light quarks. The following curves illustrate how the spectrum changes with increasing 't Hooft-coupling $g_{nc}$. The final value is shown in comparison with the experimental value shown on the right in each column.}
\label{fig:charmed_tHooft1}
\end{center}
\end{figure*}

\begin{figure*}[p]
\begin{center}
\input{Instanton2.pstex_t}
\caption{The effects of the instanton-induced interaction between a light and the charm quark on the resonances $\Xi_c$ and $\Omega_c$. The spectrum on the left in each column is determined by the confinement and the residual interaction operating only between the light quarks. The curves illustrate how the spectrum changes with firstly increasing the 't Hooft-coupling $g_{nc}$ and secondly $g_{sc}$. The final value is shown in comparison with the experimental value.}
\label{fig:charmed_tHooft2}
\end{center}
\end{figure*}

Increasing the coupling $g_{nc}$, we mainly observe the lowering of the lowest lying $\Sigma_c\frac{1}{2}^+$-state by 80\,MeV describing excellently the mass of the $\Sigma(2455)$. This selective lowering also explains the mass splitting between the lowest lying $\Sigma_c\frac{1}{2}^+$- and $\Sigma_c\frac{3}{2}^+$-states which is again 80\,MeV and almost coincides with the experimental value of 66\,MeV. This behavior of 't Hooft's interaction can again be understood from inspection of the SU(8) spin-flavor configurations in table \ref{tab:su8confSigC}. The $\Sigma_c\frac{1}{2}^+$ is a $^26_{\mathrm{M}}[\mathrm{S}]$-state which can be influenced by the extended 't Hooft-force we are considering whereas the $\Sigma_c\frac{3}{2}^+$ is a $^46_{\mathrm{S}}[\mathrm{S}]$-state and thus not changed by 't Hooft's interaction for reasons we have already explained.

We also see a lowering of the $\Lambda_c$ whose position was 16\,MeV too high before and 13\,MeV too low after we have increased $g_{nc}$ to its maximum value. So we achieve also here a slight improvement by the additional residual interaction.

Due to this strong lowering of the theoretical $\Sigma_c(2455)$ we get in the Bethe-Salpeter-model a final splitting between the $\Lambda_c$ and the $\Sigma_c(2455)$ of 187\,MeV which is very close to the already mentioned experimental value of 167 MeV. The extended version of 't Hooft's force improves this number by 52\,MeV.

The experimental splitting between the lowest lying $\Lambda_c\frac{1}{2}^-$- and $\Lambda_c\frac{3}{2}^-$-states remains unexplained. Although the additional 't Hooft-force acts only very weakly on these two states, the effect of it even has the wrong sign. Nevertheless, the mass of the $\Lambda_c(2593)$ is perfectly described.

We shall not forget the two- and one-star resonance whose masses are lowered by 17\,MeV and 21\,MeV respectively, almost matching their alleged experimental counterparts.

Figure \ref{fig:charmed_tHooft2} shows on the left in each column the theoretical single charm baryon spectrum of the $\Xi_c$ and $\Omega_c$ resonances determined by the confinement interaction and the residual interaction operating only between the light quarks which means that $g_{ns}=94.0\,\mathrm{MeV\,fm}^3$ and $g_{nc}=g_{sc}=0$. We first increase the coupling $g_{nc}$ to its maximum value and depict the resulting masses in the middle of each column. We then increase the coupling $g_{sc}$ and compare the final theoretical values to the experimental ones.

We especially observe further lowerings of the theoretical $\Xi_c$ and $\Xi'_c$ which can now happen to both states, see again table \ref{tab:su8confXiC}, in contrast to the situation in the previous section where we have considered 't Hooft's interaction only between light flavor quarks. 't Hooft's interaction between the quark pairs $nc$ and $sc$ lowers the position of the $\Xi_c$ successively by 19 and 3\,MeV and the position of the $\Xi_c'$ by 35 and 20\,MeV reproducing perfectly the mass of the $\Xi_c$ and describing very well the mass of the $\Xi_c'$. Moreover, the $\Xi_c'$ is lowered by a larger amount than the $\Xi_c$ which improves the theoretical mass splitting between them by 16\,MeV to 143\,MeV. This is closer to the experimental splitting of 108\,MeV than without the phenomenological extension of 't Hooft's force to charm quarks.

Due to its SU(8) spin-flavor configuration the theoretical mass for the $\Xi_c(2645)$ is never influenced by 't Hooft's force which is reasonable because it is already well described by the confinement potential alone.

Concerning the theoretical states for the $\Xi_c(2790)$ and the $\Xi_c(2815)$, the circumstances are very similar to the case of the $\Lambda_c(2593)$ and the $\Lambda_c(2625)$. These pairs not only have exactly the same quantum numbers but show the same behavior when compared to the experimental states. The effects of 't Hooft's force are almost negligible but again have the wrong sign and cannot explain the experimentally observed splitting.

\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|rrrr|rr|}
\hline
Mass &
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^4      6_{\mathrm{M}}[\mathrm{M}]\!\!$&
$\!\!{}^2      6_{\mathrm{M}}[\mathrm{A}]\!\!$&
$\!\!{}^4      6_{\mathrm{S}}[\mathrm{S}]\!\!$&
$\!\!{}^2      6_{\mathrm{S}}[\mathrm{M}]\!\!$\\
\hline
\multicolumn{7}{|c|}{$J^\pi=1/2^+$}\\
\hline
     2688 & {\bf \underline{97.5}} &    1.5 &    0.0 &    0.0 &    0.0 &    1.0\\
     3169 & {\bf \underline{84.3}} &    8.1 &    0.0 &    0.1 &    0.0 &    7.3\\
\hline
\multicolumn{7}{|c|}{$J^\pi=3/2^+$}\\
\hline
     2721 &    0.0 &    0.0 &    2.0 &    0.0 & {\bf \underline{97.9}} &    0.0\\
\hline
\end{tabular}
\end{center}
\caption{Masses (in MeV) and SU(8) configuration mixings (in \%) for the $\Omega_c$ ground state and excited states with total angular momentum $J$ and parity $\pi$.}
\label{tab:su8confOmeC}
\end{table}
\end{center}

When we only take the confinement potential into account the theoretical position of the $\Omega_c$ is overestimated by 23\,MeV compared to the experiment. There is trivially no contribution from 't Hooft's force between a quark pair $nc$ but a strong lowering by 33\,MeV due to 't Hooft's force between the quark pair $sc$ although the value for $g_{sc}$ is relatively small. Looking at its SU(8) spin-flavor configuration in table \ref{tab:su8confOmeC}, the $\Omega_c$ is indeed almost a pure $^26_{\mathrm{M}}[\mathrm{S}]$-state. With 't Hooft's force switched on we underestimate the mass of the $\Omega_c$ only by 10\,MeV which is much better than without 't Hooft's force.

Summarizing the effects of the 't Hooft-interaction, we can clearly state that the phenomenological extension of the instanton-in\-duced force to charmed baryons leads to visible improvements in the whole fine structure of the single charm baryon spectrum.

\subsection{One-gluon-exchange}

For comparison we replace within the same framework our phenomenologically motivated extension of 't Hooft's force by the instantaneous version of the one-gluon-exchange as a residual interaction operating between a light and a heavy quark. The interaction between the light quarks, which is taken to be the instanton-induced force, remains the same and forms the basis for all calculations presented in this work. Its effects on the single charm baryon spectrum is already discussed in sec.~\ref{sec:instlight}.

In the previous section we have seen that the extended 't Hooft-interaction is able to improve the fine structure of the spectrum which is based on the confinement potential and the 't Hooft-interaction between the light quarks. We now show that the one-gluon-exchange between a light and a heavy quark is capable of explaining some features of the experimental spectrum but cannot account for some distinguished properties concerning the fine structure.

The confinement potential used here primarily models the long-ranged confining forces. But due to its negative offset it also induces a short-ranged attractive influence which acts between all quark pairs, regardless if they are light or heavy. When we want to consider the one-gluon-exchange and 't Hooft's interaction combined in the same framework but acting between differently flavored quark pairs, we have to take into account that the one-gluon-exchange also has a short-ranged attractive effect. Avoiding a double-counting of these attractive short-ranged forces, we therefore neglect the Coulomb-term $D_{00}$ in eq.~(\ref{eqn:OGEpot}).

\begin{figure*}[t]
\begin{center}
\input{Charm_OGE.pstex_t}
\caption{The single charm baryon spectrum calculated in the Bethe-Salpeter-model using an appropriate three-quark confinement potential, the one-gluon-exchange between a light and a heavy quark and 't Hooft's force between the remaining flavor combinations (on the left side of each column) in comparison to the experimental situation taken from \cite{Eidelman:2004wy} (on the right side where the status is indicated by stars, the mass uncertainty by a shaded box and the lack of knowledge of quantum numbers by a question mark) in dependence of the total angular momentum $J$ and the parity $\pi$.}
\label{fig:onegluonexchange}
\end{center}
\end{figure*}

Figure \ref{fig:onegluonexchange} shows the resulting single charm baryon spectrum where the new parameters, namely the strong coupling constant $\alpha_s=1.0$ and the charm constituent quark mass $m_c=1940\,$MeV, are fitted simultaneously to the three- and four-star charmed baryons.

We find that there is a clear one-to-one correlation between experimental and theoretical states and in particular the masses of the lowest lying states $\Lambda_c\frac{1}{2}^+$, $\Sigma_c\frac{3}{2}^+$, $\Xi_c\frac{1}{2}^+$, $\Xi_c\frac{3}{2}^+$ and $\Omega_c\frac{1}{2}^+$ are very well described.

But the one-gluon-exchange fails completely in explaining the positions of the $\Sigma_c(2455)$ and the $\Xi'_c$, missing them by 65\,MeV and 52\,MeV. Remember that both states could be accurately described by 't Hooft's force.

The one-gluon-exchange also underestimates the resonance masses of the problematic pairs $\Lambda_c(2593)$, $\Lambda_c(2625)$ and $\Xi_c(2790)$, $\Xi_c(2815)$, on average by 20\,MeV more for each state when compared to 't Hooft's force. But the one-gluon-exchange can allusively account for their splittings: They are 7\,MeV and 10\,MeV respectively and are approximately by factors five and three too small when compared to the experimental splittings of 33\,MeV and 26\,MeV.

We can summarize that the one-gluon-exchange is able to describe the single charm baryon spectrum but the splittings it generates (in spite of the embarrassing large coupling $\alpha_s$) are too small to explain the fine structure satisfactorily. We obtain overall better results using the extended version of 't Hooft's interaction.

\subsection{Double and triple charm baryons}

\begin{center}
\begin{table}[t]
\begin{center}
\begin{tabular}{|c|c|c|c|}
\hline
State&$J^\pi$&BSM&BSM\\
&&conf.&full\\
\hline
$\Xi_{cc}$    &$1/2^+$&3723&3642\\
$\Xi_{cc}$    &$3/2^+$&3723&3723\\
$\Xi_{cc}$    &$1/2^-$&3941&3920\\
$\Xi_{cc}$    &$3/2^-$&3941&3920\\
$\Omega_{cc}$ &$1/2^+$&3765&3732\\
$\Omega_{cc}$ &$3/2^+$&3765&3765\\
$\Omega_{cc}$ &$1/2^-$&3994&3986\\
$\Omega_{cc}$ &$3/2^-$&3994&3986\\
$\Omega_{ccc}$&$1/2^+$&5216&5216\\
$\Omega_{ccc}$&$3/2^+$&4773&4773\\
$\Omega_{ccc}$&$1/2^-$&5019&5019\\
$\Omega_{ccc}$&$3/2^-$&5014&5014\\
\hline
\end{tabular}
\end{center}
\caption{The theoretical mass positions (in MeV) of the lowest lying double and triple charm baryons in dependence of the total angular momentum $J$ and parity $\pi$ calculated within the Bethe-Salpeter-model (BSM) using an appropriate three-quark confinement potential and 't Hooft's force between a light and a charm quark (labeled with ``full''). We also show the computed masses using the confinement potential only (labeled with ``conf.'').}
\label{tab:doublytriplycharmedmasses}
\end{table}
\end{center}

Without introducing additional free parameters we also compute the masses of double and triple charm baryons. Encouraged by its success in describing the fine structure of the single charm baryon spectrum, we again use the extended 't Hooft-force between a light and a charm quark which now is the only residual interaction since there is only one light quark. Table \ref{tab:doublytriplycharmedmasses} shows the results for the lowest lying states of the $\Xi_{cc}$, $\Omega_{cc}$ and $\Omega_{ccc}$. To demonstrate the effects of 't Hooft's force we also present the results using the confinement potential alone.

The only experimental clue for a double charm bary\-on, namely the $\Xi_{cc}$, stems from the SELEX Collaboration at Fermilab \cite{Mattson:2002vu}. The mass of the $\Xi_{cc}$ is supposedly $(3519\pm 1)\,$MeV, its quantum numbers are unknown. The Particle Data Group Collaboration does not consider this state as established and gives it only one star \cite{Eidelman:2004wy}.

Taking this experimental finding seriously, all our theoretical predictions given in table \ref{tab:doublytriplycharmedmasses} are too high. The most obvious identification is with the lowest lying state which has the spin-parity combination $\frac{1}{2}^+$ and mass 3642\,MeV. The deviation to the experiment is thus 3.5\% which is still not too bad.

Various other models predict the lowest lying $\Xi_{cc}\frac{1}{2}^+$-state also too high and very comparable to our result: In the framework of a relativistic quasi potential quark model \cite{Ebert:1996ec} one obtains 3660\,MeV and in a similar and more recent calculation by the same authors \cite{Ebert:2002ig} 3620\,MeV, in the framework of a simple potential model \cite{Richard:1994ae} one obtains 3630\,MeV, by exploiting regularities in the hadron interaction energies to obtain sum rules \cite{Lichtenberg:1995kg} one obtains 3676\,MeV and in quenched lattice calculations \cite{Woloshyn:2000fe,Mathur:2002ce} one obtains approximately 3600\,MeV.

Looking again at table \ref{tab:doublytriplycharmedmasses}, we see that 't Hooft's force lowers the $\Xi_{cc}\frac{1}{2}^+$-state by the relatively large amount of 81\,MeV but it seems that its strength is somehow not strong enough to explain the small experimental value.

\vspace{20pt}

\section{Conclusion} \label{sec:conclusion}

In the framework of a relativistically covariant constituent quark model we calculated on the basis of the Bethe-Salpeter-equation in its instantaneous approximation mass spectra of single, double and triple charm baryons. We used a linearly rising three-body confinement potential and as a residual interaction between light quarks always 't Hooft's instanton-induced force. Both interactions were already fixed by the non-strange and strange baryon spectrum and were left unchanged in the calculations of this work. The Bethe-Salpeter-model was therewith able to give an adequate description of the whole non-strange and strange baryon spectra.

In this paper we have extended phenomenologically 't Hooft's force to include also charm quarks introducing two additional new free coupling parameters. However, the foundations of such a force are by no means clear because strictly it exist only for nearly massless quarks. Our investigation is therefore to be considered to be of exploratory nature. Nevertheless, we keep for simplicity the name instanton-induced force. We showed the computed spectrum of the single charm baryons, compared it to the experiment and discussed in detail the instanton-induced effects. We have found not only a very good overall agreement between theory and experiment but also the hyperfine splittings were accurately described. The extended version of 't Hooft's force can in particular correctly explain the experimental mass splittings between the $\Lambda_c$ and the $\Sigma_c(2455)$ and the $\Xi_c$ and the $\Xi_c'$.

Within the same framework we have considered an alternative effective interaction by replacing the 't Hooft-interaction between a light and the charm quark by the one-gluon-exchange. We reject the one-gluon-exchange phenomenologically due to its insufficiency to reproduce the correct fine structure of the charmed baryon spectrum.

Finally we have given many predictions for masses of double and triple charm baryons to be confronted with results from further experimental analyses.

Our theoretical investigations are still incomplete because so far we have not computed electroweak couplings and strong decays of charmed baryons. This is work in progress. Indeed, a number of experimental results are now available which could be used to further test our specific model. With the exception of $\Lambda_c$, semileptonic decays have been computed so far only for light flavored baryons \cite{Migura:2006en}.

\section*{Acknowledgments}
Financial support from the Deutsche Forschungsgemeinschaft by the
SFB/Transregio 16 ``Subnuclear Structure of Matter'' and from the European
Community-Research Infrastructure Activity under the FP6 "Structuring the
European Research Area" programme (HadronPhysics, contract number
RII3-CT-2004-506078) is gratefully acknowledged.

redpajama_set_name: RedPajamaArXiv


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text:
Hedge fund firm Man AHL cuts through the noise around machine learning
Man AHL chief scientist Anthony Ledford is a speaker at Newsweek's forthcoming Data Science Capital Markets event.
By Ian Allison
Updated February 9, 2017 12:02 GMT
Much of the current machine learning revolution originated around applications like computer vision that have nothing to do with finance. It's an interesting question, the extent to which the latest artificial intelligence and deep learning techniques can crossover into finance.
Financial data modelling is beset by a low signal to noise ratio, whereas data used to teach a computer to identify a picture of a cat, for example, is unambiguous. The financial universe is a non-stationary environment with variable patterns of correlation between stocks, bonds and other instruments. Not least, the task in hand is essentially about predicting things that haven't happened yet.
More from IBTimes UK
Two Sigma's Wes McKinney talks about big data interoperability
Hedge fund firm Man AHL says open sourcing software helps attract best developer talent
The 'fintech' approach to data science and machine learning
For nearly 30 years now, UK hedge fund manager Man AHL has been trawling through enormous historical datasets trying to understand what is predictable and what's just noise. Today the company employs a team from a diverse range of scientific backgrounds and uses a combination of data science and machine learning techniques to manage significant amounts of client money.
Anthony Ledford, Man AHL's chief scientist, emphasises the importance of diversity in all things and knows never to have too much faith in any one prediction model.
Ledford said: "In that kind of data environment, where the 'grand truth' is changing, you can't just have a model that expects some static view, like learning the characteristics of a cat.
"Also, it's easy to look at historical data and to decompose it and say this is related to that, and do that retrospectively. But if you've actually got to make a snapshot based on the data you've got today and understand what predictability there is about what happens next, that is a much harder problem."
Ledford believes it's important to convey to people that this is not some magic black box, where nobody can understand what's going on inside.
"I think that's an unacceptable way of thinking of these things. You absolutely have to be able to understand how the algorithms are behaving; what they are actually learning and articulate that in ways that people can have some feeling for. In the past, there wasn't very much work done on this, whereas a lot of academic work now is looking specifically at the interpretability of these machine learning systems and I think that's a very good thing," he said.
Some financial data scientists also see as problematic the complexity of certain types of financial models constructed in the past couple of decades. Indeed, quantitative finance can be broadly divided into two categories: those who like writing down equations for how the world should behave, and those who look at data and try to detect patterns of behaviour. The former can come up with elegant formulas for pricing things, but the assumptions behind these equations may not hold robustly in the real world.
Anthony Ledford, chief scientist, Man AHL
"You can end up with market shocks that are much larger than the assumptions of the underlying probability distributions," said Ledford. "It's because of having too much faith in your model. We realise this at Man AHL. If you ask me how much faith do I have in any particular model or being able to predict an individual price of a financial instrument, well I have very little faith in it. It's slightly more than a coin flip.
"The point is to understand the uncertainty and build the portfolio in a rational way. You would never bet the house on a very concentrated trade; you have to diversify by trading hundreds of individual instruments and you have to trade them for a long period of time.
"In every single instrument that we are trading, for every single trade that we put on, we seek to have a tiny edge that is working in our favour. The way you can turn that into something that makes sense from an investment point of view, is to distil those tiny statistical edges down into something that, at the portfolio level, makes sense as an investment product," he said.
A toolbox for learning
Starting out with data rather than a hypothesis is how machine learning works. It could be described as a toolbox for learning from data where you don't pre-define the shape of the model. The software has to try and fathom that out by itself, so an order of magnitude more intelligence or complexity is involved compared to standard models.
This approach is enhanced by Bayesian methodology, which is based on a theorem that updates belief probabilities as more evidence or information becomes available. Man AHL uses a branch of machine learning called Bayesian nonparametrics: methods that can determine an appropriate model complexity directly from data.
Ledford said: "In the Bayesian area, your numerical belief in the parameters updates as the data comes in. Now, there is a very genuine question as to whether you need to update those beliefs as every single data point comes in, or can you update them for example on a calendar period basis – once a month or once a year, or whatever. I could follow the full Bayesian prescription and have all these things update as every single new data point comes in, but it's important to assess whether the impact of updating at that rate is sufficient to be worth that computational effort?"
Summing up, Ledford said understanding the uncertainly of what you are doing is probably even more important than understanding the alpha generation side. Risk has to be managed every day; you can take too little risk or you take too much risk.
"It's tempting to have an idealised view of the world where everything is stationary and you get your risk right and then your returns per day come in steadily. But in reality you tend to see that the results of strategies arise in bursts; they don't come in uniformly. There are periods when it works and there are periods when it doesn't work.
"You can try and build systems that turn on when they see these things happening and turn off when they don't. That's quite hard to do. What you find easier to do is actually build systems that are capable of performing well when the effect they are trying to capture is there; and when it's not there, they are very lean in terms of the risk they take and amount of trading they do.
"It's really about understanding the nature of this effect we are trying to capture. And that goes right the way through from the first things we have been doing in the traditional momentum and carry space to our machine learning systems."
Anthony Ledford is a speaker at Newsweek's forthcoming Data Science and Capital Markets event in London 1,2 March.
Related topics : Artificial Intelligence

redpajama_set_name: RedPajamaCommonCrawl


text:
Northport is a community in the Canadian province of Nova Scotia, located on the Northumberland Strait in  Cumberland County at the mouth of the Shinimicas River.

Parks
Amherst Shore Provincial Park
Northport Beach Provincial Park

References

 Northport on Destination Nova Scotia

Communities in Cumberland County, Nova Scotia

redpajama_set_name: RedPajamaWikipedia


This week's topic is "Top Ten Most Anticipated Releases For The Second Half Of The Year." There are some amazing books coming out during the second half of 2016 and I seriously cannot wait for them. It was SO hard to narrow my list down to only ten this week. Here they are in order by release date, titles link to Amazon (affiliate).
The Gilded Cage by Lucinda Gray (Release Aug. 2): I came across this one on another blogger's Waiting on Wednesday post and knew that I had to read it! I have a soft spot for historical fiction and this one has the promise of a little mystery as well.
Empire of Storms by Sarah J Maas (Release Sept. 6): Do I even need to explain why I'm anticipating this book?! I'm dying to get my hands on it. I need more Aelin and Rowan. I need to know what will happen to Dorian. AND I must know how the kingdoms of Erilea are handling the return of magic.
Stalking Jack the Ripper by Kerri Maniscalco (Release Sept. 20): This one started popping up all over the blogging community during BEA and I have to admit I had some book envy for those lucky enough to get an advanced copy. As I've already said, I have a soft spot for historical fiction and Jack the Ripper has always fascinated me. I can't wait to read it!
Bright Smoke, Cold Fire by Rosamund Hodge (Release Sept. 27): I'm lucky in that I received an eARC of this one from the publisher and even so I still can't wait to share it with all of you. I love Hodge! She writes some amazingly dark, gothic romances and this one is a retelling of Romeo and Juliet….yes, please!!
The Midnight Star by Marie Lu (Release Oct. 11): The conclusion to The Young Elites series and I cannot explain how excited I am for it. Adelina is one of those characters that even though it's been a year since I read the last book, I find myself thinking about her and wondering just how dark her character will become. She's already pretty dark, but I have a feeling Lu is going to give us nightmares. Bring it on!!!
A Million Worlds with You by Claudia Gray (Release Nov. 1): I adore this series and I'm so sad to see it come to an end. BUT at the same time I need it. The last book's cliffhanger destroyed me. I had a book hangover for days and I need to know what will happen to Meg, Paul, and Theo.
Turbo Twenty-Three by Janet Evanovich (Release Nov. 15): By the Numbers has been a favorite series of mine since high school. Stephanie Plum and her antics never get old, in my opinion and I can't wait to read each one.
I learned two things making this post: I like authors with the last name Gray/Grey and all the good books are coming out in September!
Stalking Jack the Ripper looks interesting!
If you like stories about him, The Giftsnatcher by Charity Bishop is awesome!
Oh I haven't heard of that one! I'll have to check it out 🙂 Thanks for sharing!
There are so many wonderful releases coming out soon. I'm looking forward to reading more from Laini Taylor! I hope you enjoy them.
The Midnight Star has an awesome cover and I should read The Young elites, several bloggers have recommended it but I just haven't picked them up yet. Bright Smoke, Cold Fire looks interesting too.
Oh The Young Elites and The Rose Society are amazing! It's an antihero/villain story. I'm not sure how Lu can make Adelina darker and honestly I'm a little terrified to find out. You should definitely put them high on your TBR ?
I was always interested in Jack the Ripper and used to read a lot of non fiction about the murders so I'm now keen to read these fiction books about it. Looking forward to this one!
You have nice books in your list. I really want Bright Smoke….
A lot of these made my list too, and while it's not on there, I am looking forward to the latest Stephanie Plum!
I would suggest reading Snow Like Ashes in a binge read 😉 The second book while good suffers from second book syndrome and reads a lot like a bridge book instead. Hope you enjoy them!
I have a Shadow Hour ARC that I should definitely start soon, and I also plan on reading Bright Smoke, Cold Fire! I haven't read anything by Rosamund Hodge yet but I am in love with this cover.
I've been eyeing Stalking Jack the Ripper lately! So many awesome picks here! So many trilogies ending this year! I still need to read all the book 2's from these it seems! I heard about so many cliffhangers, I decided to avoid all of them until book 3 came out! Hopefully book 2 will have a good recap since I read them all so long ago! LOL!
I didn't realize there was a new Laini Taylor book coming out. How wonderful!
Stalking Jack the Ripper and Bright Smoke are two that I can't wait for! I just LOVE that cover for Strange the Dreamer!!


Sure it's free. But is it easy?
Nothing lonlier than a comfy chair at the curb … hopefully onto a wonderful new life. Is there a chair rescue association out there?
So simple, a great pic! Back in my younger days I got some of my favourite things this way.
many items in our basement came from a sign like that!
A discerning eye can do quite well!
Is it from watching too many cartoons as a kid that I expect this chair to just lift itself up and wander down the street in search of a new owner? :) I do wonder, when I see older furniture, what things it's seen and what stories it could tell. Carin, you have such a great eye for the extraordinary in the ordinary. Great WW.
p.s. How many cartoons did you watch???


Club members Cruise to Redding, CA for our annual "Over-the Weekend" road tour and entry into the Cool April Nights Car Show put on by the City of Redding. This week long event is something you should attend once in your life. Over 1000 cars cruise through the streets Fri. night prior to the show on Sat. For more info, please stop by the club.
An all Chevy Day event!
BBQ burgers and hot dogs.
Any and all Chevys invited. No fee to enter. Trophies for "Oldest Chevy", Coolest Chevy Car" & "Coolest Chevy Truck".
Club open to non-members to view.
Sorry. no pets in club house.
San Rafael hosts the 31st "May Madness Car Show". We have attended this event for the past five years as a club. Our 10x10 "pop-up" tent offers plenty of cover from the heat and also a great place to hang out during the event. Always plenty of refreshments. We have picked up the "Club Participation Award" for the past 3 years. Lets do in once again!! Applications now available at The 51-50 Foundation.
Ask about the dance and BBQ following the annual Fr. night Cruise!
Our annual Summer BBQ for 51-50 Members and the Marin Coupe and Roadster Club. Lots of burgers, dogs and chicken. Good time for the ENTIRE family, but leave the pets at home. Games, prizes and raffles.
Our 4th Annual "Clean out the Garage Event" Bring whatever you have you no longer need. This combo event is auto parts and house hold items. Call to reserve your FREE space while they last. Food and beverages available!
An all Ford Day event!
Any and all Fords invited. No fee to enter. Trophies for "Oldest Ford", Coolest Ford Car" & "Coolest Ford Truck".
Our FAVORITE go-to event each year! Put on by the San Rafael Lions Club. This great event benefits local community services. On the lawn with plenty of shade, great food and folks! Meet up at The Foundation and cruise over as a group. Time TBD.
NOTE: Date subject to change.
A good time for members to get together on the 3rd Friday of each month. Pot luck gathering starting at 6pm and going until 9pm. Light appetizers provided by The Foundation. No-Host bar.
Always on the 1st Monday of each month. Starts at 6pm and goes until 10pm. Entertainment at no cost to listen. You can participate if you like or enjoy the voices of some talented singers. $5. appetizers available from 6pm to 8pm.
Back Again. Wednesday Night Drag Races!
Time to see who has the quickest ride through the qtr. mile. Club members compete at Sonoma Raceway. Fastest time slip of the season gives you bragging rights and a trophy awarded by The Foundation at the end of the season. Check with The Foundation for race dates.
Great old Sci-Fi films from the 50's. Shown on our big screen TV every Sat. afternoon.. Popcorn and "theater" candy (beverages too!).


text:
ID cards for children not a priority
Jakarta Post - February 15, 2016
Nani Afrida, Jakarta – The government's plan to issue ID cards for children has met with criticism from activists who call for better public service with regard to birth certificate issuance and child protection.
"Before issuing child ID, how about accomplishing the law's mandate on providing birth certificate to all Indonesian children first?" Seto Mulyadi, a children's rights activist, told The Jakarta Post on Sunday.
The Home Ministry announced last week that it would issue Kartu Identitas Anak (KIA), or child ID cards, beginning next month. The KIA will be the official form of identification for unmarried Indonesians below the age of 18.
Seto said that many parents still meet with difficulties when trying to obtain birth certificates for their children. They have to deal with a long bureaucratic process, which lower-income people often find difficult to access. "Make the process [to issue a birth certificate] fast and free, this is more important than providing KIA for children," he said.
Seto said many children, especially street children, had no birth certificates. As a consequence, they could not access their right to education, health and or even a place in a cemetery. "We also believe that the KIA will create discrimination and as a result the gap between the rich and poor will get even bigger," Seto added.
The Indonesian Child Protection Commission (KPAI) share similar thoughts, saying that there were many other problems related to children that should be prioritized. "The birth certificate is one of issues, the others are child protection as well as family strengthening and self defense programs for children, for them to say 'no' to people who will exploit them," said Erlinda, a spokesperson for KPAI.
In the past four years KPAI received more than 16,000 reports of violence against children across the country's 34 provinces and 179 cities. "Child ID might be important, but it is not a priority," she added.
Erlinda said KIA will help children who get into trouble due to laws or regulation. Based on the government's plan, there will be two categories of KIA, one for children under 5 years of age and one for those between 5 and 17. For newborns, the civil registration agency will issue a KIA along with their birth certificate.
However, it is not clear whether the two identification cards will have to be renewed after a period of time. The government also plans to finance the program using the state budget.
Home Ministry spokesperson Dodi Riyadmadji told the Post that 50 cities throughout Indonesia had initiated the program a year ago. "We expect that all cities will follow the program gradually," Dodi said.
He said that the 50 cities that had already started to implement KIA included Malang in East Java, Balikpapan in East Kalimantan, Bantul regency in Yogyakarta, Makassar in South Sulawesi and Depok in West Java.
Previously the minister of home affairs Tjahjo Kumolo said that KIA would be very important for the children as, for instance, the children can open bank accounts or get an Indonesia health card and an Indonesian smart card using their own names.
"If they go abroad, they don't need to have a passport attached to their parent," Tjahjo said. He also said the ministry would form a special team in each village or sub-village to input child data.
Source: http://www.thejakartapost.com/news/2016/02/15/id-cards-children-not-a-priority.html
Population & Migration
Parents told to place at least two words in naming children
Jakarta Globe - May 23, 2022
National economy declines while population rate increases: BKKBN
Antara News - April 22, 2022
Indonesia needs 3 mln jobs annually to prepare for demographic bonus
Indonesia's demographic dividend reaches peak in 2021
Jakarta Globe - January 22, 2021
Social Affairs Minister Risma to create ID cards for marginal people
Indonesia erupts in protests for third day over controversial labour laws
ABC News - October 9, 2020
Deputy police chief named deputy head of national COVID-19 recovery committee
Jakarta Post - August 13, 2020
Low participation, COVID-19 pandemic push BPS to extend national online census
Online census begins despite glitches
Government to kick off 2020 population census in February
Lack of birth certificates exposes children to risks of exploitation
Bali gov't steps up campaign to increase birth rate on the island, says more Balinese should have four children
Coconuts Bali - June 27, 2019
Jakarta to document post-Idul Fitri newcomers
Jakarta Post - June 11, 2019
Jakarta puts end to annual civil registration raids
Jakarta Post - May 31, 2019
Bogor overwhelmed by thousands of illegal immigrants
Palm oil association urges gov't to revive Suharto-era transmigration program
Suspicious citizens can use e-KTP to catch polygamous spouses
Indonesia's big cities face more migration woes
Jakarta Globe - August 3, 2014
Population control key to economic development: BKKBN
Indonesia's population boom unsustainable, experts warn
Jakarta Globe - March 27, 2014
Agency head slams Indonesia population management
Jakarta Post - July 17, 2013
Religious leaders in North Sulawesi deem E-ID project heretical
Jakarta Post - May 2, 2012
Aging population 'will place burden on Indonesia'
Jakarta Post - April 7, 2012
Booming population spawns urbanization problems: UN
Jakarta Post - October 27, 2011
Government 'fails' to control population growth


text:
Q: PayPal webhook created vs activated I'm finishing up my paypal webhook, when dealing with subscriptions is their a difference between BILLING.SUBSCRIPTION.ACTIVATED and BILLING.SUBSCRIPTION.CREATED?
When looking at the the developer docs it doesn't say anything except A subscription is activated and A subscription is created which doesn't say if there is a difference.

A: A subscription must necessarily be created before it can be activated, you can't activate a nonexistent subscription.
A subscription can be suspended and activated again.

redpajama_set_name: RedPajamaStackExchange


The office of Councilmember David Grosso focuses on many different issues with one main goal always in the forefront of our minds: making D.C. a better city. All of our efforts are based in ensuring and promoting the human rights of the people of the District. David is the chairman of the Committee on Education and sits on four other committees: Judiciary and Public Safety; Health; Human Services; and the Committee of the Whole.


I have recently received this purple Flexishield Skin for HTC Desire HD to review. I absolutely love this . It fits perfectly and I had no problems putting it on my phone. It fits well around the camera,flash, microphone and such. It covers the top button and the volume buttons but it is alright because it doesn't affect the use of those buttons at all. I have already dropped my phone once since I have this skin on it and it did protect it. I have to say I love my new HTC Desire HD skin.
Mobile Fun also has cases for newer phones like Galaxy Note 2 case and so many others that there is a case or skin for everyone. I will be buying case or skin from Mobile Fun when I get a new phone or when I have a new e-Reader in the future.
Hello, this looks like a cool mobile cover. But it doesn't look that strong. I think HTC should make mobile cases too, as its mobiles are very costly.


Where do I start? Dropbox was my go-to backup service when it came to my digital pictures and videos. I still remember discussing cloud storage with my coworker "its a simple, automatic backup for your phone's pictures". I was instantly sold. Before that, my digital pictures were scattered across my flash drive, portable hard disk and PC. What made the deal even sweeter was my Samsung Galaxy Note II promo which added free extra space.
Paying 10 bucks a month while not allot in it self-do add up. A couple years back Dropbox was the only solid gamer in town. Not so much in 2018. In fact, Dropbox offers you the least when compared to other services namely One Drive and Google Photos.
For $7 a month, One Drive offers 1 TB of storage as well as Office 365. Not bad considering your getting free full office apps along with your subscription. On top, your also getting files on demand feature that cost extra in Dropbox. Add the amazing Android support and you have a must buy deal. What extras does Dropbox offer? Non.
Speaking of Google Photos. The service offers FREE unlimited pictures and videos backup! FREE. Yes, the files are slightly compressed but the average consumer can't tell the difference and you can't beat FREE! A paid subscription gives you full resolution backups as well 1 TB of space for any other file. But what really makes Google Photos stand out is the ability to backup live photos. Seeing photos come to life is amazing plus you can download the short video on itself. As a bonus, all of your pictures are categorized and are searchable. What extras does Dropbox offer? Nothing again.
Which above subscription should you choose can be subjective but one thing is for sure: It's not Dropbox. I supported the company for the few years I used it and even advocated the service to friends and family but I can no longer recommend it to anyone. The company is clearly trying to milk its users base with the new personal plans while offering nothing in exchange. I can't even create albums for my pictures let alone organize and search them like Google Photos. So in essence, I'm paying more for less.
I know this may sound strange coming for a guy who loved Dropbox as he switched away and switched back before but one can love a product so much before realizing that the company is no longer interested in offering the best it can. I'm currently using my Galaxy S8's promotion space on One Drive while also uploading to Google Photos. I'm leaning toward dropping all and just upload to Google Photos via a paid subscription. But then again, going over the limit and not paying will stop my Gmail from working but that's fair I guess.
My final word for Dropbox: Get your act together. Taking your customers for granted while living off your legacy can work so much before they start looking for another lover. I wish Dropbox the best. As for me, I'm off to Google Photos.


text:
Prince Consort Road is a street in London, United Kingdom. It is named after Prince Albert, consort to Queen Victoria. It is located between Queen's Gate to the west and Exhibition Road to the east, running parallel to Kensington Gore.

Several landmark buildings have entrances on Prince Consort Road, including:
Royal Albert Hall, which has steps leading down from its south side to Prince Consort Road;
Royal School of Mines (now part of Imperial College London);
Holy Trinity Church;
Royal College of Music;
Beit Hall;
Jamaican High Commission.

The average residential property price on the street in 2013 was around £5.5 million, making Prince Consort Road the fourth-most expensive street in the UK.

London Buses route 360 starts and ends on Prince Consort Road. The nearest London Underground stations are South Kensington, Gloucester Road, Knightsbridge and High Street Kensington.

See also
 List of eponymous roads in London

References

Streets in the City of Westminster


text:
Project Tasks
=============

Documentation Publishing
------------------------
* Ensure you have configured a [GitHub SSH Key](https://help.github.com/articles/generating-ssh-keys)
* Pull the latest [betfair/gh-pages-publishing](https://github.com/betfair/gh-pages-publishing)
* Run ```./publish betfair/disco <github-user> <branch-name>```

Snapshot Deploy
---------------
* Ensure you have configured [servers in settings.xml](https://docs.sonatype.org/display/Repository/Sonatype+OSS+Maven+Repository+Usage+Guide)
* Pull down latest code
* Run ```mvn:deploy```

Releasing
---------
* Branch disco to create the release branch
* Perform maven release - see below
* Branch the disco-documentation to create the release documentation
* Publish the new disco-documentation branch
* Change the master disco-documentation to reference the new release
* Branch the archetypes to create the release branch
* Update the archetypes on the release branch to reference the release version
* [Perform maven release](https://docs.sonatype.org/display/Repository/Sonatype+OSS+Maven+Repository+Usage+Guide) on the archetype branch
* Update the archetypes on the master branch to reference the new SNAPSHOT
* Send notification email to betfair-disco@googlegroups.com

Perform maven release
---------------------
```
# make sure up to date
git pull
# create branch
git checkout -b 3.2
mvn versions:set -DnewVersion=3.2.0-SNAPSHOT
find . -name "pom.xml.versionsBackup" -exec rm {} \;
git commit -m "Update versions on branch"
# fix master
git checkout master
mvn versions:set -DnewVersion=3.3-SNAPSHOT
find . -name "pom.xml.versionsBackup" -exec rm {} \;
git commit -m "Update versions post branch"
# prepare release
git checkout -b 3.2
mvn release:prepare
# enter gpg passphrase when prompted
# some fudgery required now, namely to install release binaries in local repo
cd target
git clone --branch 3.2.0 git@github.com:betfair/disco checkout
cd checkout
git fetch git@github.com:betfair/disco
git checkout 3.2.0
mvn install -Dmaven.test.skip=true
cd ..
rm -rf checkout
cd ..
# now do the actual release
mvn release:perform
```

If you have any errors at the release:perform stage, then reset using the following:
```
find . -name "*.releaseBackup" -exec rm {} \;
rm release.properties
mvn versions:set -DnewVersion=3.2.0-SNAPSHOT
find . -name "pom.xml.versionsBackup" -exec rm {} \;
git add .
git commit -m "Reset versions following failed release"
git push
mvn release:clean
```

If you have errors due to a pre-existing tag, then delete using:
```
# remote
git push origin :disco-master-pom-3.2.0
# local
git tag -d disco-master-pom-3.2.0
```

redpajama_set_name: RedPajamaGithub


text:
Pseudomyrmosa minuta is een vliesvleugelig insect uit de familie van de mierwespen (Mutillidae). De wetenschappelijke naam van de soort is voor het eerst geldig gepubliceerd in 1894 door F. Morawitz.

Mierwespen


Judy Wallman, a professional genealogy researcher in southern California, was doing some personal work on her own family tree. She discovered that Senator Harry Reid's great-great uncle, Remus Reid, was hanged for horse stealing and train robbery in Montana in 1889. Both Judy and Harry Reid share this common ancestor.
Now THAT's how it's done folks! That's REAL political spin at its finest. It just doesn't get much better than that!.
CEO Magazine, which issued its annual "Best/Worst States for Business" this week, ranked Washington state below average at 34th. The more telling statistic from this survey of over 500 CEOs is that our state has fallen 18 spots over the last five years. In 2006 Washington ranked an above average 16th.
The best states: Texas, North Carolina, Florida, Tennessee and Georgia.
The worst states: California (50th), New York, Illinois, New Jersey and Michigan.
Seems like pretty good advice. It's unfortunate that enough CEOs nationwide think Washington is doing a fair-to-poor job at implementing such policies. But at least we're not Illinois, which has fallen 40 spots in five years. Ouch.
A new Elway Poll says Democratic Senator Maria Cantwell, who is facing re-election next year, could very well be vulnerable — IF, and it's a BIG IF — the Republicans can actually field a credible candidate. But then that's their usual problem with getting anyone elected. The R's can't seem to find moderates — folks whose message resonates with the majority of Washington voters — usually opting instead for right-wing extremists.
The Elway Poll supposed an imaginary open primary against a group of Republicans that included former KIRO TV anchor Susan Hutchison; Tea Party activist Clint Didier; U.S. Rep. Dave Reichert; and Seattle Port Commissioner Bill Bryant. Cantwell won big over all of them with 47 percent of the vote.
You may recall that in 2006 Cantwell beat former Safeco Insurance CEO Mike McGavick, 56.9 percent to 39.9 percent. Obviously, she could do a lot worse in 2012 and still win since the Republicans don't have anyone waiting in the wings that according to Elway would be a credible challenger. On the other hand, I've heard it said more than once that Elway couldn't poll himself accurately.
Cantwell's approval numbers in the most recent poll aren't higher now because there are fewer undecided voters. Clearly, between a May 2005 poll, until the Summer of 2009 Cantwell's approval numbers have been more volatile, with six of the last 20 polls showing her with higher disapproval, ratings but with an overall trend of a higher approval. So given the absence of any strong Republican opponent on the horizon, it doesn't appear to me at least that she has too much to worry about.