Rparity violation at the LHC
 249 Downloads
 1 Citations
Abstract
We investigate the phenomenology of the MSSM extended by a single Rparityviolating coupling at the unification scale. For all Rparityviolating couplings, we discuss the evolution of the particle spectra through the renormalization group equations and the nature of the lightest supersymmetric particle (LSP) within the CMSSM, as an example of a specific complete supersymmetric model. We use the nature of the LSP to classify the possible signatures. For each possible scenario we present in detail the current LHC bounds on the supersymmetric particle masses, typically obtained using simplified models. From this we determine the present coverage of Rparityviolating models at the LHC. We find several gaps, in particular for a stauLSP, which is easily obtained in Rparityviolating models. Using the program CheckMATE we recast existing LHC searches to set limits on the parameters of all Rparityviolating CMSSMs. We find that virtually all of them are either more strongly constrained or similarly constrained in comparison to the Rparityconserving CMSSM, including the \(\bar{U}\bar{D}\bar{D}\) models. For each Rparityviolating CMSSM we then give the explicit lower mass bounds on all relevant supersymmetric particles.
1 Introduction
Supersymmetry (SUSY) [1, 2, 3, 4] is a unique extension of the external symmetries of the Standard Model of particle physics (SM) [5, 6].^{1} As a solution to the hierarchy problem [8, 9], the supersymmetrybreaking scale should be \(\mathcal {O} ({\mathrm {TeV}})\) and thus testable at the LHC. To date no experimental sign of supersymmetry has been found [10], pushing the lower mass limits for some supersymmetric particles into the TeV range, however, with some clear model dependence; see for example Ref. [11] on discussing the impact of the \(\sqrt{s}=8\,\)TeV data.
The lack of a supersymmetric signal at the LHC puts increasing pressure on the CMSSM in particular with respect to finetuning [18, 19]. Several groups have performed combined frequentist fits of the CMSSM to all the relevant data; see for example [20, 21]. In Ref. [22] it was shown that the CMSSM is experimentally excluded due to tension between the \((g2)_\mu \) measurements and the LHC lower bounds on \(M_0\). Several groups now instead investigate the phenomenological MSSM (pMSSM), which has a more extensive parameter set [23, 24, 25, 26, 27]. In particular the slepton and squark masses at the unification scale are now given by separate parameters, decoupling the \((g2)_\mu \) measurement and the LHC lower mass bound on the squarks [26].
The bounds on the most constrained coupling for each class of RPV operators at \(M_W\). The bounds are given for sfermion masses \(\tilde{m}\) of 1 TeV and typically scale as \(\tilde{m}^{1}\). A complete list with the specific sfermion mass dependence is given in Appendix A. \(^\mathrm{a}\) We disregarded the stringent bound on \(\lambda _{133}\) and \(\lambda _{133}^\prime \) arising from upper bounds on neutrino masses [35]. \(^\mathrm{b}\) \(\lambda ^{\prime \prime }_{112}\) and \(\lambda ^{\prime \prime }_{113}\) can be more strongly constrained under the assumption of a large hadronic scale for double nucleon decay [36]
Couplings  \(\lambda _{ijk}\)  \(\lambda '_{1jk}\)  \(\lambda '_{2jk}\)  \(\lambda '_{3jk}\)  \(\lambda ''_{ijk}\) 

Bound  0.49\(^\mathrm{a}\)  \(0.09^a\)  0.59  1.1  \(0.5^{\mathrm{b}}\) 
Rparity was originally introduced to stabilize the proton. However, this is not a unique choice, with many viable alternatives [37, 38, 39, 40, 41, 42, 43]. There are also a number of simple models which predict a subset of RPV couplings through other discrete gauge symmetries; see for example [33, 44, 45]. In addition phenomenologically RPV supersymmetry models naturally accommodate light massive neutrinos, requiring neither righthanded neutrinos nor an additional heavy Majorana mass scale [32, 46, 47]. As a consequence of Rparity violation, the LSP can decay and is no longer a good dark matter candidate. However, others such as an axion, a sufficiently longlived axino [48, 49, 50, 51] or even the gravitino [52, 53] can account for the measured dark matter relic density [54]. Furthermore, RPV can alleviate part of the light Higgs problem in supersymmetry [55], as it can lead to weaker lower mass bounds on the squark and gluino; see e.g. [56, 57, 58] and the discussion below. We conclude that RPV models are just as well motivated as Rparityconserving (RPC) models.
 1.
In Sect. 2, we analyze the possible LSPs in the \(\Lambda _{\not R_p}\)–CMSSM. This depends on the type and size of the dominant RPV coupling, as well as the CMSSM parameters. We employ the supersymmetric RGEs [59] and go beyond previous work [62], to take into account the recent Higgs boson discovery.
 2.
In Sects. 3 to 5, we review in detail the RPV LHC signatures, summarize the experimental lower mass bounds and thus determine the current LHC coverage of the \(\Lambda _{\not R_p}\)–MSSM. The experiments typically set limits on the parameters of simplified models, with no interpretation in the \(\Lambda _{\not R_p}\)–CMSSM. These searches can, however, be applied to a wide range of RPV models.
 3
In Sect. 6 we investigate the \(\Lambda _{\not R_p}\)–CMSSM as a complete supersymmetric model. We use the program CheckMATE [63, 64, 65] to determine the LHC bounds on the various versions of this model and compare it to the CheckMATE constraints in the RPC–CMSSM.
 4
In Sect. 8 we use the results from Sect. 6 to determine absolute lower mass bounds on the supersymmetric particles for a \(\tilde{\chi }^0_1\)LSP and for a \(\tilde{\tau }_1\)LSP scenario in the \(\Lambda _{\not R_p}\)–CMSSM, respectively. We compare our bounds to the simplified models experimental bounds in Sects. 3 to 5.
2 The renormalization group evolution of the RPV–CMSSM
The renormalization group equations (RGEs) of supersymmetric Rparity violation have previously been studied in Refs. [36, 66, 67, 68, 69, 70]. The full twoloop equations are given in Ref. [59]. Through the interface with SARAH [71, 72, 73, 74, 75, 76], they have been implemented in the numerical program SPheno [77, 78], which we employ here.
2.1 General considerations
As is known from the RPC case, large gaugino masses contribute with a negative slope to the RGEs, thus raising the sfermion masses when running from the high to the low scale. This gaugino mass effect typically dominates over the reverse effect due to the softmass contribution \(\Lambda _{\not R_p}^2 \mathcal {F}\) to the RGEs, even for large values of \(M_0(M_X)\). Thus small \(M_0\) is in general favorable for obtaining light sfermions. The most important effect in the RGEs comes from the Aterms, i.e. \(T_\Lambda \), which always decrease the soft masses at the low scale. As is well known for the RPC–CMSSM, large \(A_0\) values of several TeV are needed in order to explain the Higgs mass, if at the same time the stops are at the TeV scale. See, e.g., Ref. [22] for a recent global fit to the RPC–CMSSM. Therefore, in a constrained SUSY model, it is natural to have large trilinear SUSYbreaking terms, leading to sizable effects in the weakscale sfermion masses.
We note that large trilinear sfermion interactions tend to make the electroweak vacuum unstable, i.e. the scalar potential can develop additional minima, which for large \(T_i\) can be deeper than the electroweak minimum [81, 82]. The latter can then tunnel to the energetically preferred configuration at possibly unacceptably large rates. However, using the numerical program Vevacious [83], we have verified that possible new minima induced by the RPV operators \(T_\Lambda \) are, for the scenarios we consider here, never deeper than the vacua we already find in the RPC–CMSSM. Therefore, the findings of Ref. [84] concerning the vacuum stability of the RPC–CMSSM also apply here. All scenarios we present in the following feature an electroweak vacuum which is either stable or metastable longlived, meaning that the tunneling time to the global minimum is longer than the age of the Universe.
2.2 Determining the LSP
The different possible LSP scenarios in the RPV–CMSSM have been first explored in Ref. [62]. This analysis was centered around comparatively small values of \(M_0,M_{1/2}\) and \(A_0\). With the recent measurement of the Higgs mass, these small values of the input parameters lead to an unacceptably light Higgs and are hence excluded. Here we reevaluate the possible LSPs in the RPV–CMSSM, taking into account the new Higgs measurements [79]. Due to the necessity of quite heavy universal softbreaking parameters, the supersymmetric contributions to the muon anomalous magnetic moment can only alleviate but not solve the wellknown discrepancy [85, 86, 87]. We first consider the case of small couplings and then large couplings which can strongly affect the RGEs.
2.2.1 Small \(\Lambda _{\not R_p}\)
2.2.2 Large \(\lambda _{ijk}\)
2.2.3 Large \(\lambda '_{ijk}\)
As the Dterm contributions to the sparticle masses slightly suppress \(m_{\tilde{\nu }}\) w.r.t. \(m_{\tilde{\ell }_L}\), only the sneutrino can become the LSP for large \(\lambda '_{ijk}\) and \(i=1,2\). For \(i=3\) the effect of the left–right mixing in the stau sector reduces the lightest stau mass below \(m_{\tilde{\nu }_\tau }\). The nonneutralino LSP candidates for the large\(\lambda '\) scenario are thus \(\tilde{\nu }_e,\,\tilde{\nu }_\mu \) and \(\tilde{\tau }_1\), where the latter is mainly a \(\tilde{\tau }_L\), unlike \(\tilde{\tau }_R\) within RPC models. However, the parameter space for a \(\lambda '\)induced nonneutralino and nonstau LSP is small because of (i) the already comparably large \(\tilde{m}_L^2\) from the RPC RGEs alone and (ii) the smaller effect of a large \(A_0\) when compared with the \(\lambda _{ijk}\ne 0\) case.
2.2.4 Large \(\lambda ''_{ijk}\)
The constraints on the \(\lambda ''\) couplings, cf. Table 13, are typically rather loose, leaving more room to have strong RPV effects on the RGE squark running. An exception are the strong bounds on \(\lambda ''_{112,113}\), cf. Appendix A. Therefore the only possibility with an operator solely coupled to the first and/or secondgeneration squarks is \(\lambda ''_{212}\). In Fig. 4 we show as an example the LSP nature for the case \(\lambda ''_{212}_ {M_X}=0.5\) in the \(M_0\)–\(\tan \beta \) plane. This is to also show the \(\tan \beta \) dependence of the LSP nature for the case of two particle species (\(\tilde{\chi }^0_1\) and \(\tilde{s}_R,\tilde{d}_R\)) whose mass is largely independent of \(\tan \beta \), and the lightest stau, whose mass depends strongly on \(\tan \beta \). The other parameters are \(M_{1/2}=1~\)TeV and \(A_0=3.3~\)TeV. Here, for large \(M_{1/2}\), the \(\tilde{d}_R\) and \(\tilde{s}_R\) are almost degenerate and can be the joint LSPs. In the figure we have disregarded the \(\tilde{s}_R\)–\(\tilde{d}_R\) mass splitting, as it is below 1 GeV. Throughout the figure the neutralino and gluino masses do not vary significantly as \(M_{1/2}\) is fixed. Their masses are \(m_{\tilde{\chi }_1^0} \simeq 430\,\hbox {GeV}\) and \(m_{\tilde{g}} \simeq 2.2\,\hbox {TeV}\). The remaining labels are as in Fig. 3. In the center of the figure we have a large \(\tilde{s}_R\)/\(\tilde{d}_R\)LSP region. To the right, for large values of \(M_0\), we again have a \(\tilde{\chi }^0_1\)LSP. In the far upperleft corner the white area indicates a tachyonic stau and/or squarks. Just below that is a small region with a \(\tilde{\tau }\)LSP. In the \(\lambda ''_{212}\)scenario the charmsquarks, in turn, cannot become the LSP because of their slightly heavier softbreaking masses at the low scale, \(\tilde{m}_{U}^2\tilde{m}_{D}^2 \simeq 0.05\,M_{1/2}^2\) [89, 90, 92].
For \(\lambda ''_{ij3},\,i\ne 3\), \(\tilde{b}_R\) couples directly to the leading RPV operator, allowing for a sbottom LSP. We always have \(m_{\tilde{b}_R}<m_{\tilde{u}_{i=1,2}},\, m_{\tilde{d}_{i=1,2}}\) due to the larger RPC bottom Yukawa coupling.
2.3 Summarizing the LSP Scenarios
Summary of the various LSP scenarios in the \(\Lambda _{\not R_p}\)–CMSSM as a function of the dominant necessary RPV coupling at \(M_X\).\(^\mathrm{a}\) Note that the couplings on the right column are required but their presence is not necessarily sufficient to get the corresponding LSP. The couplings \(\lambda '_{111},\,\lambda ''_{112}\) and \(\lambda ''_{113}\) are too constrained to produce an LSP of that kind; see Table 13
LSP  Required couplings 

\(\tilde{\chi }^0_1\)  \(\Lambda _{\not R_p}\ll 1\) or large \(M_0\) 
\(\tilde{\tau }_1\)  \(\Lambda _{\not R_p}\ll 1\), small \(M_0\) and large \(M_{1/2}\) 
\(\tilde{\tau }_1\)  \(\lambda _{ij3}\) (dominantly \(\tilde{\tau }_R\)), \(\lambda '_{3jk}\) (\(\tilde{\tau }_L\)) 
\(\tilde{e}_R\)  \(\lambda _{ij1} \) 
\(\tilde{\mu }_R\)  \(\lambda _{ij2}\) 
\(\tilde{\nu }_e\)  \(\lambda '_{1jk},~\{j,k\} \ne \{1,1\}\) \(^\mathrm{a}\) 
\(\tilde{\nu }_\mu \)  \(\lambda '_{2jk}\) 
\(\tilde{s}_R,\tilde{d}_R\)  \(\lambda ''_{212}\) (degenerate LSPs) 
\(\tilde{b}_1\)  \(\lambda ''_{123},\,\lambda ''_{213},\,\lambda ''_{223}\) \(^\mathrm{a}\) (dominantly \(\tilde{b}_R\)) 
\(\tilde{t}_1\)  \(\lambda ''_{3jk}\) (dominantly \(\tilde{t}_R\)) 
2.4 RGEinduced operators
The fourbody decay branching ratio \(\tilde{\tau }\rightarrow \tau \mu +2\,\)jets for \(\lambda '_{211}_{M_X}=0.07\) is shown in the left plot of Fig. 8 as a function of \(\tan \beta \) and \(A_0\). For \(A_0\simeq 2300\,\)GeV, at the finger shaped region, there is a small resonance in the partial width \(\tilde{\tau }\rightarrow \tau \mu +2\)j. This occurs due to a level crossing in the mixing between the left and righthanded gauge eigenstates of the staus as \(A_0\) increases. Subsequently, the largest branching ratio occurs where the stau left–right mixing is maximal as the RPV operator \(L\bar{Q} \bar{D}\) involves only lefthanded sleptons. One should note that this level crossing only appears as a result of the large RPV coupling. For smaller RPV couplings, the right smuon is always the lighter one. However, for the relatively large RPV value used in Fig. 8, the RGE effects of this coupling in conjunction with \(A_0\) drive the lefthanded slepton soft mass towards smaller values, leading to a level crossing at a particular value of \(A_0\). Once again, the gray parameter region is excluded as the Higgs mass is too small.
If we consider the analogous scenario for \(\lambda '_{222}_{M_X}=0.07\) instead of \(\lambda '_{211}\), shown in the left plot of Fig. 8, the partial widths of the fourbody decays do not change. On the other hand the RGEinduced coupling \(\lambda _{233}\) is larger because of the significantly larger strangequark Yukawa coupling \((Y_d)_{22}\gg (Y_d)_{11}\). Therefore the corresponding twobody partial width is much larger. For a more detailed discussion of the effect of the RGEinduced operators and the impact of fourbody decays, we refer to Ref. [93].
3 LHC coverage of RPVinduced neutralino LSP decay scenarios
As we saw in the previous section, throughout wide ranges of the RPV–CMSSM parameter space, and in particular for \(\Lambda _{\not R_p}\ll 1\), the LSP is given by the lightest neutralino. We first discuss in this section the decay lifetime of the neutralino, to see what ranges of parameter space we probe when restricting ourselves to prompt decays. We then discuss in detail the LHC finalstate signatures, depending on the dominant RPV operator at \(M_ X\), and their coverage by existing LHC searches.
3.1 Neutralino lifetime
Besides being proportional to the Rparityviolating coupling squared, as we see in Eq. (21), the decay width scales as the fifth power of the neutralino mass, which is strongly connected to \(M_{1/2}\), and the fourth inverse power of the scalar fermion propagator mass, which is not only strongly correlated with \(M_0\), but also depends on \(M_{1/2}\) via the RGEs. When comparing the top left with the lower panels in Fig. 9, we see that the lifetime can be quite different for equally sized \(\Lambda _{\not R_p}\) at the unification scale. \(\lambda ''\) is almost a factor three larger than \(\lambda \) at the low scale, due to the RGEs. However for \(\lambda \) the slepton masses in the propagator are much lighter than the squark masses for \(\lambda ''\). The propagator effect dominates, as \(\Gamma \propto (\lambda ^{(\prime \prime )})^2M^{4}_{\tilde{f}}\). Furthermore, when comparing the \(M_{1/2}\) dependence for \(\Lambda _{\not R_p}=\lambda \) versus \(\Lambda _{\not R_p}=\lambda ''\), there is a much slower decrease of \(c\tau \) with increasing \(M_ {1/2}\) for the \(\lambda ''\) scenarios, since the squark masses also increase with \(M_{1/2}\). In the lower panels we also see a marked difference in the decay lengths for the two couplings, \(\lambda ''_{323}\) and \(\lambda ''_{122}\). For \(\lambda ''_{323}\) the finalstate top quark leads to a phasespace suppression if \(M_{1/2}\) and therefore \(m_{\tilde{\chi }^0_1}\) is small. For larger \(M_{1/2}\) as well as large \(M_{0}\) (corresponding to large \(A_0\) in the setup at hand), a lighter virtual top squark, in comparison to the first/secondgeneration squarks, is more important. This lighter top squark leads to smaller decay lengths for \(\lambda ''_{323}\) compared to \(\lambda ''_{121}\) in these regions.
When \(m_{\tilde{\chi }_1^0}<m_t+m_b\) (not shown in the figure), the three body decay \(\tilde{\chi }^0_1\rightarrow tbs\) is kinematically forbidden. We found that the corresponding fourbody decay \(\tilde{\chi }^0_1 \rightarrow t^{(*)} bs \rightarrow (b W^+) b s\) can be prompt for \( \lambda _{323}''\sim \mathcal {O}(0.1)\) and \(m_t+m_bm_{\tilde{\chi }_1^0} \lesssim 10~\)GeV. In addition, CKMeffects in the RGE evolution of the \(\lambda ''\) coupling lead to a nonzero \(\lambda ^{\prime \prime }_{223}\) at the scale of the decaying particle and therefore open the alternate threebody decay \(\tilde{\chi }^0_1 \rightarrow cbs\). However, we found the 4body decay remains the dominant channel.^{4} The RGEgenerated coupling is loop suppressed and thus if only the decay \(\tilde{\chi }^0_1\rightarrow cbs\) is kinematically accessible, the lightest neutralino is stable on detector scales. Note, however, that a neutralino as light as a top quark or lighter requires \(M_{1/2} \lesssim 410~\)GeV, resulting in \(m_{\tilde{g}} \lesssim 1~\)TeV, and can therefore be safely regarded as excluded, as we see below.
3.2 Neutralino LSP decay via an \(LL\bar{E}\) operator
Possible charged lepton final states in the \(LL\bar{E}\) case for a pair of LSP neutralinos resulting from the cascade decays of pair/associated produced SUSY particles. In each case, if distinct, the charged conjugate final state is also possible. The various charge combinations have equal branching ratios, due to the Majorana nature of the \(\tilde{\chi }^0_1\)LSP. All final states are accompanied by (at least) two neutrinos typically leading to some missing transverse momentum
Scenario  Charged lepton signatures  RPV operators 

Ia  \(e^+ e^ e^+ e^\)  \(\lambda _{121,131}\) 
Ib  \(\mu ^+ \mu ^ \mu ^+ \mu ^\)  \(\lambda _{122,232}\) 
Ic  \(\tau ^+ \tau ^ \tau ^+ \tau ^\)  \(\lambda _{133,233}\) 
Id  \(e^+ e^e^\pm \mu ^\mp \)  \(\lambda _{121}\) 
Ie  \(e^+ e^e^\pm \tau ^\mp \)  \(\lambda _{131}\) 
If  \(\mu ^+ \mu ^\mu ^\pm e^\mp \)  \(\lambda _{122}\) 
Ig  \(\mu ^+ \mu ^\mu ^\pm \tau ^\mp \)  \(\lambda _{232}\) 
Ih  \(\tau ^+ \tau ^\tau ^\pm e^\mp \)  \(\lambda _{133}\) 
Ii  \(\tau ^+ \tau ^\tau ^\pm \mu ^\mp \)  \(\lambda _{233}\) 
Ij  \(e^+\mu ^e^\pm \mu ^\mp \)  \(\lambda _{121,231,122,132}\) 
Ik  \(e^+\tau ^e^\pm \tau ^\mp \)  \(\lambda _{131,231,123,133}\) 
I\(\ell \)  \(\mu ^+\tau ^\mu ^\pm \tau ^\mp \)  \(\lambda _{132,232,123,233}\) 
Im  \(e^\tau ^+\mu ^\pm \tau ^\mp \)  \(\lambda _{123}\) 
In  \(e^\mu ^+\tau ^\pm \mu ^\mp \)  \(\lambda _{132}\) 
Io  \(e^\mu ^+e^\pm \tau ^\mp \)  \(\lambda _{231}\) 
As can be seen, for each dominant RPV operator, we can get SFOS (same flavor, opposite sign) lepton pairs. In all cases we can also get SFSS (same flavor, same sign) lepton pairs. This includes the somewhat exotic signatures (Im) \(\tau ^\tau ^e^+\mu ^+\), (In) \(\mu ^\mu ^e^+\tau ^+\), and (Io) \(e^e^\mu ^+\tau ^+\). For each dominant operator one should check which final state leads to the optimal experimental sensitivity. In the case of a discovery, we see that each operator has two alternate channels, with definitive branching ratios, which should give a good experimental cross check.
Best limits in RPV searches using simplified models and \(LL\bar{E}\) operators. The pairproduced SUSY particles are assumed to decay down to the neutralino LSP which itself always decays as \(\tilde{\chi }^0_1\rightarrow \ell _i^\pm \ell _k^\mp \nu _j,\, \ell _j^\pm \ell _k^\mp \nu _i,\,\) for \(L_iL_j\bar{E}_k\). The bounds are only estimates, as they have been read off the relevant plots. The first column shows the particle on which a bound is set. The second and third columns show the lower mass bounds and the \(LL\bar{E}\) operator which has been assumed for the respective scenario
Particle  Lower bound (GeV)  \(LL\bar{E}\) coupling  Simplified model  Comment  References 

\(\tilde{\chi }_1^0\)  900 (740)  \(\lambda _{122}\) (\(\lambda _{123,233}\))  \(m_{\tilde{\chi }^\pm _1} = m_{\tilde{\chi }^0_1}+1\,\)GeV  Wino production  [110] 
\(\tilde{\chi }_1^0\)  900 (560) [260]  \(\lambda _{122}\) (\(\lambda _{123}\)) [\(\lambda _{233}\)]  \(m_{\tilde{\chi }^\pm _1} = m_{\tilde{\chi }^0_1}+1\,\)GeV  Higgsino production  [110] 
\(\tilde{\chi }^\pm _1\)  Up to 750 (470)  \(\lambda _{121}\) (\(\lambda _{133}\))  \(\tilde{\chi }^\pm _1\rightarrow W^\pm \tilde{\chi }^0_1\)  \(\tilde{W}^ \tilde{W}^+\) production  [56] 
\(\tilde{\chi }_1^\pm \)  Up to 1100  \(\lambda _{121,122}\)  \(\tilde{\chi }^\pm _1\rightarrow W^\pm \tilde{\chi }^0_1\)  13 TeV update of [56]  [113] 
\(\tilde{\ell }^\pm _L\)  500 (425)  \(\lambda _{121,122}\) (\(\lambda _{133,233}\))  \(\tilde{\ell }^\pm \rightarrow \ell ^\pm \tilde{\chi }^0_1\)  \(N_\ell \ge 4\), 8 TeV  [107] 
\(\tilde{\ell }^\pm _R\)  425 (325)  \(\lambda _{121,122}\) (\(\lambda _{133,233}\))  \(\tilde{\ell }^\pm \rightarrow \ell ^\pm \tilde{\chi }^0_1\)  \(N_\ell \ge 4\), 8 TeV  [107] 
\(\tilde{\nu }_L\)  450  \(\lambda _{121,122}\)  \(\tilde{\nu } \rightarrow \nu \tilde{\chi }^0_1\)  \(N_\ell \ge 4\), 8 TeV  [107] 
\(\tilde{q}\)  1850 (1750) [1600]  \(\lambda _{122}\) (\(\lambda _{123}\)) [\(\lambda _{233}\)]  \(\tilde{q}\rightarrow q\tilde{\chi }^0_1\)  \(N_\ell \ge 3\), 8 TeV  [112] 
\(\tilde{t}_R\)  950 (900) [900]  \(\lambda _{122}\) (\(\lambda _{123}\)) [\(\lambda _{233}\)]  \(\tilde{t}_R\rightarrow t\tilde{\chi }^0_1\)  \(m_{\tilde{\chi }^0_1}=300\,\)GeV in [112]  
\(\tilde{g}\)  1450 (1270) [1200] {1050}  \(\lambda _{121,122}\) (\(\lambda _{123}\)) [\(\lambda _{233}\)] {\(\lambda _{133}\)}  \(\tilde{g}\rightarrow q\bar{q}\tilde{\chi }^0_1\)  \(N_\ell \ge 3\), 8 TeV 
Employing Refs. [107, 112] we see that at least at the level of simplified models the neutralino LSP model with a dominant \(L_iL_j\bar{E}_k\) operator has been tested at the LHC, setting lower mass bounds. When going to the full \(\Lambda _{\not R_p}\)–CMSSM we expect these bounds to be weaker, as the rates will be degraded through additional decay modes. However, several distinct decay chains will contribute to a signal rate, possibly compensating the above degradation. In Sect. 6 we set bounds on the \(\Lambda _{\not R_p}\)–CMSSM parameter space using the LHC searches which have been implemented in the program CheckMATE [63, 64], but which have not necessarily been designed for RPV searches. We note that a fit similar to [22] could possibly exclude these models, even though the dark matter constraint does not apply. In Sect. 8 we give the explicit resulting lower mass bounds for the individual supersymmetric particles.
ATLAS has performed an RPV–mSUGRA/CMSSM search [107] for the fixed parameters \(M_0=A_0=0\) excluding \(M_{1/2}<800\,\)GeV. This corresponds roughly to a gluino mass of 1.8 TeV. CMS has also performed an RPV–CMSSM analysis using 9.2 fb\(^{1}\) of data at \(\sqrt{8}\,\)TeV for the specific coupling \(\lambda _{122}\) [112]. They obtain a lower bound of \(M_{1/2}\gtrsim 1200\,\)GeV for \(M_0=1000\,\)GeV, \(\tan \beta =40\) and \(A_0=0\). This corresponds roughly to a lower gluino mass bound of 2.6 TeV, and a lower squarkmass bound of 1.9 TeV, at this benchmark point.
We also note that in scenarios IjI\(\ell \), we have with equal rates the special signatures SFSSSF’SS, i.e. for two distinct lepton flavors they have same flavor, same sign. These should have an even lower background and in particular for the \(\tau \) scenarios could lead to improved bounds.
The special case of stop pair production followed by the cascade decay to neutralinos which then decay via \(LL\bar{E}\) operators was also investigated in Ref. [112]. In Refs. [110, 115], simplified models of squark, gluino and stop pair production have been considered. For not too light neutralino masses they obtain lower mass bounds of about 1750 GeV for the squarks, 1500 GeV for the gluinos and 950 GeV for the top squark, when considering \(\lambda _{122}\). Similar results are obtained for \(\lambda _{121}\).
In addition to colored production, the electroweak production of wino or higgsinolike neutralinos has been considered [110]. In Ref. [56], the pair production of winolike charginos is considered. It is assumed that the charginos are the NLSPs which decay to \(W\tilde{\chi }^0_1\). The results are then interpreted in terms of \(\lambda _{121}\ne 0\) and \(\lambda _{133}\ne 0\), with the chargino mass bound depending on the mass difference \(m_{ \tilde{\chi }^\pm _1}m_{\tilde{\chi }^0_1}\). The respective lower mass bounds range up to 750 GeV for \(\lambda _{121}\) and 470 GeV for \(\lambda _{133}\). The analysis has been updated using 13 TeV data [113], yielding bounds which range up to 1140 GeV, when considering \(\lambda _{12a}\ne 0,~a=1,2\).
However, in a realistic model, a winolike chargino is accompanied by a winolike neutralino. The relevant associated production \(pp\rightarrow \tilde{\chi }^0 \tilde{\chi }^\pm \) has a larger cross section than neutralino or chargino pair production alone. This has been taken into account in Ref. [110] where wino and higgsinoLSPs are treated separately in a simplified model setup, always assuming the associated chargino state to be \(\sim 1~\)GeV heavier. However, it is further assumed that the bino is sufficiently heavy to play no role in either the production or decay. The resulting bounds for winos are about 900 GeV for \(\lambda _{122}\) and 740 GeV for both \(\lambda _{123}\) and \(\lambda _{233}\). For a higgsinolike neutralino LSP the range of the bounds is much larger, constraining the mass below 260–900 GeV, where the strongest bound is obtained for \(\lambda _{122}\) and the weakest one for \(\lambda _{233}\).
In Sect. 6.2.2, we will compare this simplified model approach directly with the bounds we obtain for a more realistic scenario with CMSSM boundary conditions. In that case, the bino is always the LSP whose production cross section is, however, negligible compared to the wino and higgsino production. This results in a decay of the produced wino/higgino state down to a bino first which then itself decays via the \(LL\bar{E}\)operatorinduced threebody decay. We will see that, using the search strategy of Ref. [56] and also taking into account electroweak neutralinochargino production, we can improve upon the bounds which have been obtained in Ref. [56].
3.3 Neutralino LSP decay via an \(LQ\bar{D}\) operator
Possible final states in the \(LQ\bar{D}\) case for a pair of neutralino LSPs decaying via the same operator. There will be further accompanying particles from the cascade decay of the originally produced particles, e.g. for squark pair production \(\tilde{q}\tilde{q}^*\rightarrow q\bar{q}\tilde{\chi }^0_1\tilde{\chi }^0_1\), giving two extra jets. \(\ell \) denotes a charged lepton and the indices \(a,b,c = 1,2\) denote leptons or quarks from the first or second generation. We have separated out the signatures with no charged lepton. For each listed signature there is a corresponding chargeconjugate signature due to the Majorana nature of the neutralino
Scenario  Signature  \(LQ\bar{D}\) operator 

IIa  \([\ell ^+_a\ell ^\pm _a,\,\ell ^+_a \not \!\!E_T]\; 4j\)  \(\lambda ^\prime _{abc}\) 
IIb  \([\ell ^+_a\ell ^\pm _a,\,\ell ^+_a \not \!\!E_T]\; 2b\,2j\)  \(\lambda ^\prime _{ab3}\) 
IIc  \([\ell ^+_a\ell ^\pm _a \bar{t} {\mathop {t}\limits ^{{\tiny ()}}},\,\ell ^+_a\bar{t}\,b\not \!\!E_T]\;2j\)  \(\lambda ^\prime _{a3c}\) 
IId  \([\ell ^+_a\ell ^\pm _a \bar{t} {\mathop {t}\limits ^{{\tiny ()}}},\,\ell ^+_a\bar{t}\,b\not \!\!E_T]\;2b\)  \(\lambda ^\prime _{a33}\) 
IIe  \([\tau ^+\tau ^\pm ,\,\tau ^+ \not \!\!E_T]\;4j\)  \(\lambda ^\prime _{3bc}\) 
IIf  \([\tau ^+\tau ^\pm ,\,\tau ^+ \not \!\!E_T]\; 2b\,2j\)  \(\lambda ^\prime _{3b3}\) 
IIg  \([\tau ^+\tau ^\pm \bar{t} {\mathop {t}\limits ^{{\tiny ()}}},\,\tau ^+\bar{t}\,b\not \!\!E_T]\;2j\)  \(\lambda ^\prime _{33c}\) 
IIh  \([\tau ^+\tau ^\pm \bar{t} {\mathop {t}\limits ^{{\tiny ()}}},\tau ^+\bar{t}\,b\not \!\!E_T]\;2b\)  \(\lambda ^\prime _{333}\) 
IIi  \(4j\not \!\!E_T\)  \(\lambda ^\prime _{abc},\lambda '_{3bc}\) 
IIj  \(2b\,2j\not \!\!E_T\)  \(\lambda ^\prime _{ab3},\lambda '_{3b3},\lambda '_{a3c},\lambda '_{33c}\) 
IIk  \(4b\not \!\!E_T\)  \(\lambda '_{a33},\lambda ^\prime _{333}\) 
Best limits in RPV searches using simplified models and \(LQ\bar{D}\) operators. Here \(a,b,c\in \{1,2\}\). The pairproduced supersymmetric particles are assumed to decay directly to the neutralino LSP. The bounds are only estimates, as they have been read off the relevant plots. The neutralino LSP always decays as \(\tilde{\chi }^0_1\rightarrow ( \ell _i^u_j\bar{d}_k,\;\nu _i d_j\bar{d}_k)+c.c.\) for \(L_iQ_j\bar{D}_k\). We have included the analysis of a scalar top decaying via a chargino and not the neutralino LSP in the last line, since it is similar. The neutralino decay can be blocked due to the heavy top quark. The search in Ref. [112] allows for simultaneously nondecoupled squarks and gluinos. In Ref. [114] CMS was able for a given \(m_{\tilde{\chi }^0_1} \in [200,800]\)GeV to exclude a range of scalar top squark masses, without a fixed lower bound. We have included the bound on the smuon mass, which is from a search for resonant production, since it also decays via the neutralino LSP. Each mass bound is for a fixed value of the RPV coupling
Particle  Lower bound (GeV)  \(LQ\bar{D}\) coupling  Simplified model  Comment  References 

\(\tilde{\chi }^0_1 \)  720 (620) [660] {500}  \(\lambda '_{131}\) (\(\lambda '_{131}\)) [\(\lambda '_{233}\)] {\(\lambda '_{233}\)}  \(\tan \beta = 2\) (40) [2] {40}  [110]  
\(\tilde{\mu }\)  440 (825) [1290]  \(\lambda '_{211}=\)0.003 (0.01) [0.04]  \(\tilde{\mu }\rightarrow \mu \tilde{\chi }^0_1\)  Res. \(\tilde{\mu }\) prod, \(m_{\tilde{\chi }^0_1}=200\,\)GeV  [117] 
\(\tilde{q}\)  1160 (1090) [1065]  \(\lambda '_{abc,ab3}\) (\(\lambda '_{3bc}\)) [\(\lambda '_{3b3}\)]  \(\tilde{q}\rightarrow q\tilde{\chi }^0_1\)  \(m_{\tilde{\chi }^0_1}=0.5\, m_{\tilde{q}}\)  [116] 
1315 (1360) [1225] {1215}  \(\lambda '_{abc}\) (\(\lambda '_{ab3}\)) [\(\lambda '_{3bc}\)] {\(\lambda '_{3b3}\)}  \(m_{\tilde{\chi }^0_1}=0.9\, m_{\tilde{q}}\)  [116]  
1310 (1400) [2000]  \(\lambda '_{23c,233}\)  \(m_{\tilde{g}}\lesssim 2000\, (1500) \,[1000]\,\)GeV  [112]  
\(\tilde{g}\)  1010 (970) [1070] {1050}  \(\lambda '_{abc}\) (\(\lambda '_{ab3}\)) [\(\lambda '_{3bc}\)] {\(\lambda '_{3b3}\)}  \(\tilde{g}\rightarrow q\bar{q}\tilde{\chi }^0_1\)  \(m_{\tilde{\chi }^0_1}=0.1\, m_{\tilde{g}}\)  [116] 
1135 (1085) [1220]  \(\lambda '_{abc}\) (\(\lambda '_{ab3}\)) [\(\lambda '_{3bc,3b3}\)]  \(m_{\tilde{\chi }^0_1}=0.5\, m_{\tilde{g}}\)  [116]  
1285 (1260) [1200]  \(\lambda '_{abc}\) (\(\lambda '_{ab3}\)) [\(\lambda '_{3bc,3b3}\)]  \(m_{\tilde{\chi }^0_1}=0.9\, m_{\tilde{g}}\)  [116]  
2000 (1500) [1000]  \(\lambda '_{23c,233}\)  \(m_{\tilde{q}}\lesssim 1310\, (1400) \,[2000]\,\)GeV  [112]  
1520 (1770) [1820]  \(\lambda '_{abc}\)  \(m_{\tilde{\chi }^0_1}= 100\,(500)\,[890]\,\)GeV  
\(\tilde{t}\)  890 (1000)  \(\lambda '_{1bc}\) (\(\lambda '_{2bc}\))  \(\tilde{t}\rightarrow b(\ell ^+2j)_{\tilde{\chi }^+_1}\)  \(m_{\tilde{\chi }^+_1}=100\,\)GeV  [120] 
580  \(\lambda '_{3bc}\)  \(\tilde{t}\rightarrow b(\tau ^+2j)_{\tilde{\chi }^+_1}\)  \(m_{\tilde{\chi }^+_1}=100\,\)GeV  [121]  
710 (860)  \(\lambda '_{132}\) (\(\lambda '_{232}\))  \(\tilde{t}\rightarrow 2b(\ell ^+j)_{\tilde{\chi }^+_1}\)  \(m_{\tilde{\chi }^+_1}=m_{\tilde{t}}(100\,\)GeV)  [122] 
A summary of the experimental lower mass bounds on the supersymmetric particles for a given dominant coupling is given in Table 6. For squarks we have lower mass bounds ranging from about 1 to 1.4 TeV. In one special case for light gluino masses there is a bound of 2 TeV. The lower gluino mass bounds are similar, ranging from about 1 to 1.3 TeV, with some stricter bounds achieved in special scenarios with light squarks or heavy neutralino LSPs. The pairproduction cross section of neutralino LSPs is typically much smaller than for squarks and gluinos resulting in the correspondingly weaker lower limits ranging from 500 to 720 GeV. Unlike the \(LL\bar{E}\) case there are no RPV searches for charginos or sleptons here. In the last three lines of the table we have included lower limits on the top squark, which in the case at hand decays via a chargino instead of a neutralino while the chargino decays to a charged lepton and two jets. The lower mass bounds range from 580 to 1100 GeV depending on the flavor of the charged lepton and the jets. This decay is not in the spirit of this section, but we considered it similar enough to include, as the neutralino decay could be blocked by the heavy top quark.
Regarding the coverage of the \(LQ\bar{D}\) Rparityviolating signatures by explicit Rparityviolating searches, we see that Ref. [116] focused on leptons plus jets signatures. They consider \(\tau \) leptons and b quarks, but explicitly omit top quarks, and they also do not consider the neutrino \(\not \!\!\!E_T\) cases. They thus cover the signatures IIa, IIb, IIe, and IIf. Comparing with Table 6 we see that [116] covers a wide range of possible \(LQ\bar{D}\) couplings, however it omits the couplings \(\lambda '_{i3k}\), since in that case the charged leptons are accompanied by a heavy top quark. Reference [112] explicitly looked for the cases \(\lambda '_{23k}\), partially covering the signatures IIc and IId. The signatures IIgIIk either involve charged leptons with top quarks, or have \(\not \!\!\!E_T\) signatures instead of the charged leptons. None of these have been covered by explicit RPV searches at the LHC. In particular the couplings \(\lambda '_{33k}\) have not been looked for. Note that in Ref. [110], CMS did search for electroweak gaugino production decaying via \(\lambda '_{331,333}\), however the sensitivity was insufficient to lead to any bound. In Ref. [116] explicitly looked for \(\tau \)’s and bquarks, giving specific sensitivity to the cases \(\lambda '_{3jk}\) and \(\lambda '_{ij3}\). However, they did not look for top quarks together with charged leptons and therefore explicitly omitted \(\lambda '_{i3j}\).
We make a special mention of Ref. [117], where CMS analyzed resonant smuon production, with the smuon decaying via the neutralino LSP [123, 124]. The production cross section is proportional to the RPV coupling squared and in order to get an appreciable rate requires \(\lambda '_{211}\gtrsim 0.003\). The lower mass limit then depends strongly on the assumed coupling value, as can be found in Table 6. We point out that CMS have interpreted this search also in terms of the RPV–CMSSM.
We note that some of the signatures listed in Table 5 are also covered by RPC searches. The first two scenarios, IIa and IIb, involve only light leptons and jets, possibly bjets. They always include an option also with \(\not \!\!E_T\). The scenarios IIi–k involve multijet events with missing transverse momentum, but zero leptons. These correspond to the standard RPC supersymmetry searches; see for instance Refs. [125, 126, 127] for isolated leptons and jets, and [128, 129, 130, 131, 132] for zero leptons and multijets accompanied by missing energy. The signatures including a \(\tau \) lepton, i.e. IIeIIh are in principle also covered by these multijet analyses. In this case, a hadronically decaying tau lepton is not explicitly tagged but handled as a hadronic object. Thus the couplings \(\lambda '_{i3k}\) have in principle been probed via the scenarios IIi–k. The exact sensitivity will only be known once the corresponding experimental searches have been interpreted in terms of these RPV models.
All of the searches listed in Table 6, except the smuon search, employ minimal or nexttominimal simplified models, with an intermediate neutralino LSP state in the decay chain. Thus it is difficult to see how the bounds in Table 6 are modified in the case of realistic cascadedecay branching ratios, as for example in the CMSSM. In Sect. 6, we shall use LHC analyses implemented in the computer program CheckMATE to obtain realistic limits on the parameters of the RPV–CMSSM models. We shall see for example that the search of Ref. [128] which looks for RPC as well as \(\bar{U} \bar{D} \bar{D}\)RPV is very sensitive to the signatures arising from the \(\lambda '_{3ij}\) operators within CMSSM boundary conditions. In Ref. [116], a reinterpretation of the RPC searches of Refs. [125, 128, 129, 133] in terms of leptonnumberviolating SUSY is presented. The resulting lower bounds on the gluino mass are around 1 TeV, for some cases similar limits are obtained for squark masses. In a CMSSM context, the bounds are even stronger as we see in Sect. 8.
3.4 Neutralino LSP decay via an \(\bar{U} \bar{D}\bar{D}\) operator
Possible final states in the \(\bar{U}\bar{D}\bar{D}\) case for a pair of neutralino LSPs decaying via the same operator. There will be further accompanying particles from the cascade decay of the originally produced particles, e.g. for squark pair production \(\tilde{q}\tilde{q}^*\rightarrow q\bar{q}\tilde{\chi }^0_1\tilde{\chi }^0_1\), giving two extra jets. The indices \(a,b = 1,2\) denote quarks from the first or second generation
Scenario  Signature  \(\bar{U}\bar{D}\bar{D}\) operator 

IIIa  6j  \(\lambda ^{\prime \prime }_{a12}\) 
IIIb  (2b)(4j)  \(\lambda ^{\prime \prime }_{ab3}\) 
IIIc  (2t)(4j)  \(\lambda ^{\prime \prime }_{312}\) 
IIId  (2t)(2b)(2j)  \(\lambda ^{\prime \prime }_{3b3}\) 
Best limits in RPV searches using simplified models and \(\bar{U}\bar{D}\bar{D}\) operators. Here \(a,b,c\in \{1,2\}\). The pairproduced supersymmetric particles are assumed to decay directly to the neutralino LSP if not stated otherwise. The bounds are only estimates, as they have been read off the relevant plots. The neutralino LSP always decays as \( \tilde{\chi }^0_1\rightarrow \bar{u}_i\bar{d}_j\bar{d}_k +c.c.\) for \(\bar{U}_i\bar{D}_j\bar{D}_k\). Each mass bound is for a fixed value of the RPV coupling
Particle  Lower bound (GeV)  \(\bar{U}\bar{D}\bar{D}\) coupling  Simpl. model  Comment  References 

\(\tilde{q}\)  1725 (1900) [2800]  \(\lambda ''_{112}\)  \(\tilde{q}_R\rightarrow j(\ell \ell \tilde{\chi }^0_1)_{\tilde{\chi }^0_2}\)  \(m_{\tilde{g}}\le 2400\, (1500)\, [1200]\,\)GeV  [112] 
\(\tilde{t}\)  950 (980)  \(\lambda ''_{3b3}\)  \(\tilde{t}\rightarrow t\tilde{\chi }^0_{1,2}/b\tilde{\chi }^+_1\)  \(\tilde{\chi }^0_1=\tilde{H}\) (\(\tilde{B}\)),  [119] 
\(m_{\tilde{\chi }^0_1} =300\,\)GeV  
1090 (1260)  \(\lambda ''_{3b3}\)  \(\tilde{t}\rightarrow t\tilde{\chi }^0_{1,2}/b\tilde{\chi }^+_1\)  \(\tilde{\chi }^0_1=\tilde{H}\) (\(\tilde{B}\)),  [119]  
\(m_{\tilde{\chi }^0_1} =800\,\)GeV  
\(\tilde{g}\)  1200 (1500) [2400]  \(\lambda ''_{112}\)  \(\tilde{g}\rightarrow jj(\ell \ell \tilde{\chi }^0_1)_{\tilde{\chi }^0_2}\)  \(m_{\tilde{q}}\le 2800\, (1900)\, [1725]\,\)GeV  [112] 
1850 (2100)  \(\lambda ''_{112}\)  \(\tilde{g}\rightarrow t\bar{t}\tilde{\chi }^0_1\)  \(m_{\tilde{\chi }^0_1}=100\, (800)\,\)GeV  [119]  
\(m_{\tilde{g}}>2m_t+m_{\tilde{\chi }^0_1}\)  [119]  
650 (950) [1020]  \(\lambda ''_{212}\)  \(\tilde{g}\rightarrow q\bar{q} \tilde{H}^0_1\)  \(m_{\tilde{q}}\le 100\, (500)\, [900]\,\)GeV  [138]  
\(m_{\tilde{H}^0_1}=\frac{3}{4} m_{\tilde{q}},\,m_{\tilde{c}}<m_{\tilde{g}}\)  
675 (1020) [1075]  \(\lambda ''_{212}\)  \(\tilde{g}\rightarrow q\bar{q} \tilde{H}^0_1\)  \(m_{\tilde{q}}\le 100\, (500)\, [900]\,\)GeV  [138]  
\(m_{\tilde{H}^0_1}=\frac{3}{4} m_{\tilde{q}};\;m_{\tilde{b}}<m_{\tilde{g}}\)  
650 (1020) [1100]  \(\lambda ''_{213}\)  \(\tilde{g}\rightarrow q\bar{q} \tilde{H}^0_1\)  \(m_{\tilde{q}}\le 100\, (500)\, [900]\,\)GeV  [138]  
\(m_{\tilde{H}^0_1}=\frac{3}{4} m_{\tilde{q}},\,m_{\tilde{b}}<m_{\tilde{g}}\)  
650 (990) [1075]  \(\lambda ''_{213}\)  \(\tilde{g}\rightarrow q\bar{q} \tilde{H}^0_1\)  \(m_{\tilde{q}}\le 100\, (500)\, [900]\,\)GeV  [138]  
\(m_{\tilde{H}^0_1}=\frac{3}{4} m_{\tilde{q}};\;m_{\tilde{c}}<m_{\tilde{g}}\)  
1040 (1555)  \(\lambda ''_{ijk}\)  \(\tilde{g}\rightarrow q\bar{q}\tilde{\chi }^0_1\)  \(m_{\tilde{\chi }^0_1}=100\,(900)\,\)GeV  [57]  
all \(\lambda ''_{ijk}\not =0\)  
800 (1050)  \(\lambda ''_{abc}\)  \(\tilde{g}\rightarrow 5q\)  \(m_{\tilde{\chi }^0_1}=50\,(600)\,\)GeV  [142] 
In Table 8 we have collected the best LHC lower mass bounds on the squark mass (treating the stop separately) and the gluino mass for various dominant couplings. We see that most searches have been performed for the case of gluino pair production. The mass bounds range from 650 GeV up to 2400 GeV depending on the scenario. For the case of a simplified model with only a light gluino and a neutralino mass of \(m_{\tilde{\chi }^0_1}=100\,\)GeV, ATLAS obtained a lower bound of 840 GeV. The search employed the “total jet mass of largeradius” [135]. Compared to a jetcounting analysis, it was shown that this technique allows for slightly higher sensitivity for light jets, whereas the jetcounting analysis provides the better bounds in the case of btagging requirements. The strictest bound of 2400 GeV is achieved in a CMS search for a simplified model which also contains kinematically accessible squarks [112].
In the case of the stop we have an ATLAS search, which also assumes an intermediate chargino [119], giving rise to an additional decay mode. The resulting lower mass bounds are of the order 1 TeV.
The most pertinent question for us is how well the \(\bar{U}\bar{D}\bar{D}\) models with a neutralino LSP are covered by searches at the LHC. Looking at Table 8 we see that the search from Ref. [57] in the last line seems to cover all possible \(\lambda ''_{ijk}\). However, ATLAS here explicitly assumed that all \(\lambda ''_{ijk} \not =0\) simultaneously, with every coupling taking the same value. Thus the sensitivity of the search could rely unduly on bottom quarks from \(\lambda ''_{ij3}\) and/or from top quarks from \(\lambda ''_{3jk}\) couplings. If we look at explicit searches based on the single coupling dominance assumption, then we have gluino searches for \(\lambda ''_{112,212,213}\) [112, 118, 119, 138]. All such searches make additional assumptions on the squark and/or neutralino masses. We would expect the search for \(\lambda ''_{213}\) to be equally sensitive to \(\lambda ''_{ 223}\). Similarly we would expect the \(\lambda ''_{213}\) search to apply equally to \(\lambda ''_{113}\), provided \(m_{\tilde{u},\tilde{d}}<m_{\tilde{g}}\), and to \(\lambda ''_{123}\), provided \(m_{\tilde{u},\tilde{s}}<m_{\tilde{g}}\). Thus for the case of gluino production there is no coverage only of the three couplings \(\lambda ''_{3jk}\). This should be performed by the LHC experimental groups. As we see from Table 7, the final states involve either two top quarks and four jets or two top quarks, two bottom quarks and two jets, plus of course the accompanying jets from the cascade decay.
In Sect. 6, we also take into account dominant \(\lambda ''_{3jk}\) couplings in the context of CMSSM boundary conditions and show that we can set bounds on this scenario using multijet analyses as well as searches for samesign leptons. We do however expect a boost in discovery potential when designing a dedicated search for the signature outlined above.
As we see in rows two and three of Table 8, the searches for top squarks cover at least two of these missing scenarios, namely \(\lambda ''_{3b3}\). Overall then the case \(\lambda ''_{312}\) is missing for the scenarios considered here.
Once again, none of these models have been interpreted in terms of the CMSSM, thus it is difficult to see how realistic these mass bounds are. We shall come back to this question in Sect. 6 and with explicit mass bounds in Sect. 8.
4 LHC coverage of RPVinduced stau LSP decay scenarios
As we saw in Sect. 2, even for small RPV couplings, we have substantial regions of CMSSM parameter space, where the LSP is the lightest stau, \(\tilde{\tau }_1\). The question is: how does the phenomenology change compared to a neutralino LSP, and what are the experimental constraints? The stau LSP case has been discussed from the theoretical perspective in some detail in the literature [34, 93, 95, 96, 146]. In [95] a set of appropriate LHC benchmarks was defined. Here we are interested in the case of small, but not too small \(\Lambda _{\not R_p}\). Thus the stau is the LSP but it still decays promptly. We shall consider the longlived case where the stau leads to detached vertices, or is even stable on detector scales, elsewhere.
 (a)
\(\tilde{\tau }\) LSP with the dominant operator directly coupling to a tau/stau: \(\Lambda _{\not R_p} \in \{\lambda _{aj3,a3c},\,\lambda '_{3jk}\}\).
 (b)
\(\tilde{\tau }\) LSP with the dominant operator not coupling to tau/stau \(\Lambda _{\not R_p} \in \{\lambda _{12c},\,\lambda '_{ajk},\,\lambda ''_{ijk}\}\).
In the stau LSP parameter regions, the lightest stau is mostly a \(\tilde{\tau }_R\), so that the coupling to the wino is reduced w.r.t. the bino. In conjunction with the hierarchy in the CMSSM of \(m_{\tilde{B}}<m_{\tilde{W}}\), this means that the decays via charginos, i.e. the ones resulting in \(\nu _\tau +X\) are suppressed and the final states including a \(\tau ^\pm \) dominate. However, for \(\lambda ''_{3jk}\) this is not true. In this case, all neutralinomediated channels end up in a top quark, which is kinematically suppressed w.r.t. the charginomediated \(\nu _\tau b\,d_j d_k\) mode, unless the stau is very heavy.
LHC signatures for a stau LSP from the decay of two staus. \(a,b,c=1,2\); \(i,j,k=1,2,3\). We did not distinguish between top quarks, bottom quarks and the jets arising from the first two generations. These are analogous to Tables 5 and 7. The special case in the last line denoted with \(^\mathrm{a}\) refers to scenarios where the stau LSP is light so that the decay into a top quark is kinematically disfavored, leading to the dominance of the charginomediated final state \(\nu _\tau d_3 d_j d_k\) as in Eq. (34)
Operators  LHC signatures  Couplings 

\(LL\bar{E}\)  2\(\tau \) 4\(\ell \,\not \!\!E_T\)  \(\lambda _{12c} \) 
2\(\ell \,\not \!\!E_T\)  \(\lambda _{a3b,ab3,a33}\)  
\(\tau \,\ell \) \(\not \!\!E_T\)  \(\lambda _{a33}\)  
2\(\tau \) \(\not \!\!E_T\)  \(\lambda _{a33}\)  
\(LQ\bar{D}\)  4j 2\(\tau \) 2\(\ell \)  \(\lambda '_{aij}\) 
4j 2\(\tau \) 1\(\ell \,\not \!\!E_T\)  \(\lambda '_{aij}\)  
4j 2\(\tau \) \(\not \!\!E_T\)  \(\lambda '_{aij}\)  
4j  \(\lambda '_{3ij}\)  
\(\bar{U}\bar{D}\bar{D}\)  6j 2\(\tau \)  \(\lambda ''_{ijk}\) 
6j \(\not \!\!E_T\)  \(\lambda ''_{3jk}\) \(^\mathrm{a}\) 
The complete listings of the stauLSP LHC signatures for all RPV operators, including both \(LL\bar{E}\) and \(LQ\bar{D}\) as well as the two and fourbody stau decays, are given in Ref. [96], assuming a cascade originating from a squark. See Tables 2–4 therein. We summarize the signatures of two decaying staus in Table 9. Depending on the production mechanism of the two staus at the LHC, the signature will be accompanied by extra taus, missing \(\not \!\!E_T\) and jets. We do not distinguish between top quarks, bottom quarks and the jets arising from the first two generations. These are dependent on the generation indices of the couplings and are analogous to Tables 5 and 7. We see that these scenarios always involve multiple \(\tau 's\) in the final state. For all \(LL\bar{E}\) and \(LQ\bar{D}\) couplings we can also have \(\not \!\!E_T\) arising from neutrinos. In the \(LL\bar{E}\) case it is possible to have first and secondgeneration charged leptons. Contrastingly, the \(\bar{U}\bar{D}\bar{D}\) scenario contains challenging final states with only two taus and jets.
LHC analyses explicitly looking for \(\tilde{\tau }\) LSPs are rarely performed. The only one we are aware of with prompt stau decays is Ref. [106], where the CMSSM with \(\lambda _{121}=0.032\) at \(M_X\) is considered. It is assumed there that the stau always decays via a fourbody decay into \(\tau e\mu \nu _e\) and \(\tau e e \nu _\mu \) with equal probability and the results are interpreted in the \(M_{1/2}\tan \beta \) plane, assuming \(M_0=A_0=0\). Values of \(M_{1/2}\lesssim 820~\)GeV could be excluded for most values of \(\tan \beta \). Note, however, that the RGEinduced operator \(\lambda _{233}\) and the corresponding partial twobody decay widths, \(\Gamma _2\), of \(\tilde{\tau }_1\rightarrow (\mu \nu _ \tau ,\,\tau \nu _\mu )\) have not been taken into account. While there is only a mild dependence of the ratio of the total fourbody versus twobody decay width, \(\Gamma _4/\Gamma _2\), on \(M_{1/2}\) [93], the ratio, however, scales with \(1/\tan ^2\beta \). Therefore, only taking into account \(\Gamma _4\) is a welljustified assumption in the low \( \tan \beta \) region, but the \(\tilde{\tau }\)LSP results of Ref. [106] for large \(\tan \beta \) can, unfortunately, not be trusted. Note also that Ref. [106] uses data from the 7 TeV run of the LHC. Because of the high occurrence of \(\tilde{\tau }\) LSPs in RPV models, we want to encourage the LHC collaborations to reanalyze these scenarios with more recent data.
In conclusion these models are basically not covered at all by LHC searches. Individual final states listed in Table 9 have been searched for in other contexts, however. In particular, we would like to mention a recent analysis looking for multilepton final states and including up to two hadronic tau tags, with various signal regions requiring different amounts of \(\not \!\!\!E_T\) [147]. This search is dedicated to RPC models where pairproduced electroweak gauginos decay via an intermediate slepton or sneutrino down to the lightest neutralino. Including the leptonically decaying tau modes, this search is therefore in principle sensitive to the \(LL\bar{E}\) and \(LQ\bar{D}\) signatures of Table 9. Unfortunately, so far no interpretation in terms of RPV models has been performed. Thus the corresponding mass sensitivity is also unknown. In Sect. 6, we discuss the impact of \(\tilde{\tau }\) LSPs on the LHC sensitivity, making use of nondedicated searches and taking into account RGEgenerated operators as well as fourbody decays. In Sect. 8 we present our best estimate of the LHC mass bounds in these scenarios.
5 LHC coverage of RPVinduced nonstandard LSP decay scenarios
The standard scenario in the \(\Lambda _{\not R_p}\)–CMSSM with small RPV couplings is the neutralino or stau LSP. Instead, as we saw in Sect. 2.2, for various large RPV couplings, we can also have the following LSPs: \(\tilde{e}_R,\,\tilde{\mu }_R,\,\tilde{\nu }_e,\,\tilde{\nu }_\mu ,\,(\tilde{s}_R,\tilde{d}_R),\,\tilde{b}_1,\) and \(\tilde{t}_1\), cf. Table 2. Here we discuss the phenomenology of these models and how they are covered by LHC searches. We also briefly summarize other related searches at the LHC with nonstandard LSPs, which do not occur in the \(\Lambda _{\not R_p}\)–CMSSM.
5.1 Slepton LSPs
5.1.1 Selectron or Smuon LSP
Lower mass bounds for the case of a stop LSP discussed in Sect. 5.2. Here \(b=1,2\)
Particle  Lower bound (GeV)  \(\bar{U} \bar{D} \bar{D}\) coupling  Simpl. model  Comment  References 

\(\tilde{t}\)  405  \(\lambda ^{\prime \prime }_{312}\)  \(\tilde{t}\rightarrow qq\)  445–510 GeV also excluded  [148] 
385  \(\lambda ^{\prime \prime }_{3b3}\)  \(\tilde{t}\rightarrow \bar{d}_b\bar{b} \)  [139]  
\(\tilde{g}\)  1440  \(\lambda ''_{312}\)  \(\tilde{g} \rightarrow t \tilde{t}\)  \(m_{\tilde{t}}=800\,\)GeV  [118] 
1460  \(\lambda ''_{3b3}\)  \(\tilde{g} \rightarrow t \tilde{t}\)  \(m_{\tilde{t}}=700\,\)GeV  [118] 
5.1.2 Sneutrino LSP
Note that other interesting scenarios can occur if we go beyond the CMSSM boundary conditions and/or consider more than one RPV operator at \(M_X\). We thus mention searches for an schannel production of tausneutrino LSPs, \(\tilde{\nu }_\tau \), which decay further into leptons, assuming both \(\lambda '_{311}\) and one of \(\{\lambda _{132},\,\lambda _{133},\,\lambda _{232}\}\) to be sizable [149, 150]; see also Sect. 5.3.1. However, the excluded combination of masses and couplings is not yet competitive with the more stringent bounds from the nonobservation of \(\mu \)–e conversion in nuclei [151, 152].
5.1.3 Stau LSP with large \(\lambda _{ij3}\) or \(\lambda '_{3jk}\)
This scenario directly corresponds to the case (a) in Sect. 4, just with a large coupling constant. The 2body stau decays will be prompt and the previous discussion holds.
5.2 Squark LSPs
 1.
A first or secondgeneration righthanded downlike squark with a large coupling \(\lambda ''_{212}\), decaying to first or secondgeneration quark jets. Experimentally this would lead to four jets, arising from the decay of two onshell strongly interacting particles, i.e. these are pairproduced dijet resonances.
 2.
A righthanded bottom squark decaying via \(\lambda ''_{123,2b3}\), with the same signature as in the previous case.
 3.
A righthanded top squark decaying via \(\lambda ''_{3jk}\). For \((j,k)=(1,2)\) this is again as in case 1. above. The pairproduced stops lead to two dijet resonances. For \(k=3\) there will be bottom quark jets in the final state.
Lower mass bounds on supersymmetric particles as the LSP decaying directly via an RPV operator. These are all not \(\Lambda _{\not R_P}\)–CMSSM scenarios. The bounds marked by an asterisk \(^*\) are reinterpreted leptoquark scenarios. For the leptoquark searches, we have included the SU(3), SU(2), and U(1)\(_{{\mathrm {EM}}}\) quantum numbers in the comment
Particle  Lower bound (GeV)  RPV coupling  Simpl. model  Comment  References 

\(\tilde{\nu }_\tau \)  1280 (3300)  \(\lambda '_{311}\cdot \lambda _{231,132}\)  \(d_j\bar{d}_k\rightarrow \tilde{\nu }_\tau \rightarrow e^\mu ^+\)  \(\lambda '_{311}=\lambda _{231,132}=0.01\,(0.2)\)  
2300  \(\lambda '_{311}\cdot \lambda _{231,132}\)  \(d_j\bar{d}_k\rightarrow \tilde{\nu }_\tau \rightarrow e^\mu ^+\)  \(\lambda '_{311}=\lambda _{231,132}=0.07\)  [158]  
2200  \(\lambda '_{311}\cdot \lambda _{133}\)  \(d_j\bar{d}_k\rightarrow \tilde{\nu }_\tau \rightarrow e^\tau ^+\)  \(\lambda '_{311}=\lambda _{133}=0.07\)  [158]  
1900  \(\lambda '_{311}\cdot \lambda _{233}\)  \(d_j\bar{d}_k\rightarrow \tilde{\nu }_\tau \rightarrow \mu ^\tau ^+\)  \(\lambda '_{311}=\lambda _{233}=0.07\)  [158]  
\(\tilde{u}_{Lb}\)  1050\(^*\) (1080\(^*\))  \(\lambda '_{1bc}\) (\(\lambda '_{2bc}\))  \(\tilde{u}_b\rightarrow ej\,(\mu j)\)  \({\mathrm {Br}}=1\), (3,2, \(+\frac{2}{3}\))  
\(\tilde{d}_{Lb}\)  \(625^*\)  \(\lambda ^\prime _{ib3} \)  \(\tilde{d}_{Lb}\rightarrow b\nu _i\)  \({\mathrm {Br}}=1\), (3,2,\(\frac{1}{3}\))  [159] 
\(\tilde{d}_{Rc}\)  \(900^*\) (\(850^*\))  \(\lambda ^\prime _{1bc}\ (\lambda ^\prime _{2bc})\)  \(\tilde{d}_c\rightarrow ej/\nu _e j\,(\mu j/\nu _\mu j)\)  \({\mathrm {Br}}=0.5\) each, (\(\bar{3}\),1,\(\frac{1}{3}\))  [159] 
\(480^*\)  \(\lambda ^\prime _{i3c}\)  \(\tilde{d}_c\rightarrow b \nu _i\)  \({\mathrm {Br}}=0.5\), (\(\bar{3}\),1,\(\frac{1}{3}\))  [159]  
\(\tilde{d}_R\)  650 (450)  \(\lambda ''_{313}\)  \(\tilde{d} \rightarrow \bar{b} \bar{t}\)  \(m_{\tilde{g}}=1.4\,\)TeV (\(m_{\tilde{g}}=2\,\)TeV)  [118] 
570 (420)  \(\lambda ''_{321}\)  \(\tilde{d} \rightarrow \bar{s} \bar{t}\)  \(m_{\tilde{g}}=1.4\,\)TeV (\(m_{\tilde{g}}=2\,\)TeV)  [118]  
\(\tilde{t}_L\)  1100  \(\lambda ^\prime _{133}\, (\lambda ^\prime _{233})\)  \(\tilde{t}\rightarrow e^+b\, (\mu ^+b)\)  [160]  
740  \(\lambda ^\prime _{333}\)  \(\tilde{t}\rightarrow \tau ^+ b\)  [121]  
\(1010^*\,(1080^*)\)  \(\lambda ^\prime _{132}\ (\lambda ^\prime _{232})\)  \(\tilde{t}\rightarrow e^+ j\,(\mu ^+ j)\)  \({\mathrm {Br}}=1\), (3,2,\(+\frac{2}{3}\))  [122]  
\(\tilde{b}_R\)  307  \(\lambda ^{\prime \prime }_{3b3}\)  \(\tilde{b}\rightarrow \bar{t}\bar{d}_b\)  [110]  
\(560^*\)  \(\lambda ^{\prime }_{333}\)  \(\tilde{b}\rightarrow t\tau ^ \)  Br\(=0.5\), (\(\bar{3}\),1,\(\frac{1}{3}\))  [161]  
\(\tilde{g}\)  650  \(\lambda ^{\prime \prime }_{112}\)  \(\tilde{g}\rightarrow uds\)  [140]  
835  \(\lambda ^{\prime \prime }_{113}\) (\(\lambda ^{\prime \prime }_{113}\))  \(\tilde{g}\rightarrow udb\; (csb)\)  [140]  
1360  \(\lambda ^{\prime \prime }_{323}\)  \(\tilde{g}\rightarrow tsb\)  \(m_{\tilde{q}}=5\,\)TeV  [143]  
\(917\, (929)\, [874]\)  \(\lambda ''_{abc} (\lambda ''_{ab3}) [\lambda ''_{3b3}]\)  \(\tilde{g}\rightarrow 3q\)  \(m_{\tilde{q}}=5\,\)TeV  [142] 
5.3 Nonneutralino LSP outside the CMSSM
Experimentally there are quite a few searches for sparticles directly decaying through an RPV operator. These correspond to scenarios where the sparticle at hand is the LSP. For completeness we briefly collect here the cases which are not possible within the CMSSM and which are therefore not listed in Table 2.
5.3.1 Sneutrino LSP
5.3.2 Squark LSP
On the other hand, a lefthanded downlike squark, \(\tilde{d}_{Lj}\), can only decay via a neutrino, leading to \(\not \!\!\!E_T\), cf. Eq. (47). In the table listings we interpret an ATLAS leptoquark search as the decay of a \(\tilde{d}_ {Lb},\,b=1,2\), to a bottom quark resulting in a weaker lower mass bound of 625 GeV [159]. The righthanded downlike squark has two possible decay modes, one involving a charged lepton and one involving a neutrino. Combining the two often leads to stricter bounds. For jets from the first two generations \(\lambda '_{1bc,2bc},\,b,c=1,2\), the lower experimental bound is about 900 GeV. For the case \(\lambda '_{i3c},\,c=1,2,\) the charged lepton mode involves a top quark, which would be a separate search. The neutrino mode alone leads to the much weaker bound of only 480 GeV. These are again all reinterpreted leptoquark searches.
The top squark LSP has been more widely considered in the literature, as it is naturally lighter than the other squarks, even for RPC; see for example [169, 170, 171]. The decays via \(\bar{U}\bar{D} \bar{D}\) operators are discussed in Sect. 5.2. The decays via \(LQ\bar{D}\) operators are just special cases of the decay of uplike squarks, Eq. (46), and the bounds are similar, around 1100 GeV. The case \(\lambda '_{333}\) leads to a finalstate tau and thus weaker bounds, around 750 GeV.
In the \(\Lambda _{\not R_p}\)–CMSSM, for \(\lambda ''_{3b3}\), the top squark can be the LSP, but not the bottom squark. Nevertheless the direct decay of a bottom squark via \(\lambda ''_{3b3}\) with 100% branching ratio was searched for giving a lower mass bound of 307 GeV [110]. A leptoquark search was reinterpreted as the direct decay of a righthanded bottom squark via the operator \(L_3Q_3\bar{D}_3\) to a top quark and a tau with a branching ratio of 50%. In this case the lower mass bound is 560 GeV [161].
5.3.3 Gluino LSP
However the first case, \(LL\bar{E}\), leads to identical signatures as in Sect. 3.2, the only difference is that now the intermediate neutralino is virtual. The second case, \(LQ\bar{D}\), is novel, although very similar to the electroweak production in Section 10 of Ref. [110]. There the pair production of neutralinos is investigated, followed by their threebody RPV decay. Here one should consider the pair production of gluinos, which has a significantly higher cross section, which should lead to stricter lower mass bounds.
The best bounds for the third case in Eq. (49) are listed in Table 11. The weaker bounds in the first two rows for the gluino were obtained with \(\sqrt{s}=8\,\)TeV data. Nevertheless the tbs search is the most sensitive channel. ATLAS has several searches for this scenario [57, 134, 135], however, they always allow more than one coupling to be nonzero, often even all \(\lambda ''_{ijk}\) with equal value. It is again not clear how to interpret the resulting bounds.
6 Testing the RPV–CMSSM with CheckMATE
For the remainder of this paper we are interested in the sensitivity of the LHC with respect to the \(\Lambda _{\not R_p} \)–CMSSM, i.e. to a complete supersymmetric model. In particular, we are interested in how the presence of Rparityviolating operators affects the wellknown results for the Rparityconserving CMSSM [20, 22, 172, 173]. For this purpose we shall use the program CheckMATE [63, 64, 65]. As we saw in Sects. 3 to 5, the LHC experiments mainly set bounds on simplified supersymmetric Rparityviolating models. They set little or no bounds on the complete \(\Lambda _{\not R_p}\)–CMSSM model. We here use CheckMATE to recast ATLAS and CMS searches and thus set bounds on the various \(\Lambda _{\not R_p}\)–CMSSM models.
6.1 Method
The program CheckMATE automatically determines if a given parameter point of a particular model beyond the Standard Model (BSM) is excluded or not by performing the following chain of tasks. First, the Monte Carlo generator MadGraph [174] is used to simulate proton–proton collisions. The resulting parton level events are showered and hadronized using Pythia 8 [175]. The fast detector simulation Delphes [176] applies efficiency functions to determine the experimentally accessible finalstate configuration, including the determination of the jet spectrum using FastJet [177, 178]. Afterwards, various implemented analyses from ATLAS and CMS designed to identify different potentially discriminating finalstate topologies are used.
Events which pass welldefined sets of constraints are binned in signal regions for which the corresponding prediction for the Standard Model and the number of experimentally observed events are known. By comparing the predictions of the Standard Model and the user’s BSM model of interest to the experimental result using the CL\(_{\text {S}}\) prescription [179], CheckMATE concludes if the input parameter combination is excluded or not at the 95% confidence level. For more information we refer to Refs. [63, 64, 65].
6.1.1 Model setup
For the proper description of the RPV Feynman rules in MadGraph , we take the model implementation from SARAH, which we already used in Sect. 2, and export it via the UFO format [180]. For a given set of \(\Lambda _{\not R_p}\)–CMSSM parameters, we make use of the respective SPheno libraries created from the same SARAH model used to determine the lowenergy particle spectrum, the mixing matrices and the decay tables. We calculate the SUSY and Higgs masses including RPVspecific twoloop corrections [181, 182, 183] which are particularly important for light stops [55]. As discussed in Sect. 2.4, fourbody decays of the stau can be dominant and lead to important experimentally accessible final states. In regions where this occurs, see for instance Sect. 6.2.4, the fourbody stau decays have been determined using MadGraph.
6.1.2 Monte Carlo simulation

We include only twobody supersymmetric finalstate production: \(p p \rightarrow A B+X_{\mathrm {soft}}\), and require both supersymmetric particles, A and B, to be produced onshell. Note that in RPC supersymmetry, additional hard QCD radiation, i.e. \(pp \rightarrow A B j\), is important in parameter regions with highly degenerate spectra due to the resulting kinematic boost of the decay products; see e.g. Ref. [184]. However, due to the instability of the LSP in the \(\Lambda _{\not R_p}\)–CMSSM, this additional boost is not needed and therefore this final state is not expected to contribute sizably to the final constraining event numbers.

We do not consider finalstate combinations which are strongly suppressed by the relevant parton density distributions and/or which only exist in RPV supersymmetry. Most importantly, this excludes flavoroffdiagonal combinations of “sea”squarks (we clarify the meaning below) or squark–slepton combinations.

We include in our simulations production processes, which can only proceed via the electroweak interactions, i.e. the production of sleptons, electroweak gauginos and the mixed production of electroweak gauginos and squarks or gluinos. However, in the case of electroweak gauginos we only include the production of the two lightest neutralinos and the lightest chargino, i.e. the dominantly bino and wino states in a CMSSM setup. We expect no sizable contributions from the ignored Higgsinos, since these are typically significantly heavier and therefore have negligible production rates in comparison with the lighter winos. Similarly, we do not include flavoroffdiagonal slepton combinations and mixed electroweak gaugino–squark production with “sea”squarks.
 With the above considerations, the resulting set of production channels that we consider are listed below: Strong processes:Electroweak processes:

\(\tilde{g} \tilde{g}\)

\(\tilde{g} \tilde{q}_{V}^{(*)}\)

\(\tilde{q}_{V}^{(*)} q_{V}^{(*)}\) (all combinations)

\(\tilde{q}_{S}\tilde{q}_{S}^{*}\) (only flavordiagonal)
Mixed processes:
\(\tilde{\chi } \tilde{\chi }\) [all (nonHiggsino) combinations]

\((\tilde{\ell }_{L}, \tilde{\ell }_{R}, \tilde{\nu }_{\ell , L}) (\tilde{\ell }_{L}^* , \tilde{\ell }_{R}^* , \tilde{\nu }_{\ell , L}^*)\) (only flavordiagonal)
Here, \(\tilde{q}_V\) refers to the superpartners of the light quarks: \(\tilde{u}_{L,R}, \tilde{d}_{L,R}\) and \(\tilde{s}_{L,R}\), while \(\tilde{q}_S\) refers to the remaining squarks \(\tilde{c}_{L,R}, \tilde{b}_{1,2}\) and \(\tilde{t}_{1,2}\). Furthermore, \(\tilde{\chi }\) subsumes the two lightest neutralinos \(\tilde{\chi }_{1, 2} ^0\) and the lightest chargino \(\tilde{\chi }_1^\pm \).
\(\tilde{g} \tilde{\chi }\)

\(\tilde{q}_{V}^{(*)} \tilde{\chi }\) (all combinations)


Decays of supersymmetric particles are performed within Pythia 8 , using the information from the decay table determined in SPheno. This ignores any potential spindependent information, which could be relevant when performing the proper matrixelement calculation. However, as we do not assume spineffects to be important here, we take the computationally faster approach of using decay tables.

To take into account the sizable contributions from higherorder QCD effects in the production cross section, we multiply the leadingorder production cross sections taken from MadGraph with the nexttoleadinglogarithm Kfactors determined by NLLFast [185, 186, 187, 188, 189, 190, 191]. This tool interpolates gluino and squarkmassdependent higherorder cross sections for all “strong Processes” listed above. NLLFast assumes degenerate first and secondgeneration squark sector where we use the median of the squark masses for the calculation. Note that this degeneracy is present in models with small \(\Lambda _{\not R_p}\) but not necessarily if \(\Lambda _{\not R_p}\) is large.^{8} Stops and sbottoms are always treated separately and obtain individual Kfactors. The consideration of higherorder effects for the remaining processes is computationally far more involved as tools like PROSPINO [192] need to perform the full NLO calculation. These effects are, however, expected to be significantly smaller compared to the strong production processes and thus we neglect them here.
6.1.3 Incorporated analyses
CheckMATE provides a large set of implemented ATLAS and CMS results from both the \(\sqrt{s} = 8\) and the 13 TeV runs of the LHC. These analyses target a large variety of possible final states which typically appear in theories beyond the Standard Model. The vast majority, however, are designed to target RPC supersymmetry. This implies cuts which require significant amounts of missing transverse momentum and/or highly energetic finalstate objects, namely leptons or jets. As many of the most prominent decay chains in the \(\Lambda _{\not R_p}\)–CMSSM indeed correspond to these signatures, it is therefore interesting to determine the relative exclusion power of these tailored analyses in comparison to the RPC–CMSSM.
In this study we consider all \(\sqrt{s} = 8\) TeV LHC analyses implemented in CheckMATE 2.0.1. For a full, detailed list we refer to the documentation in Refs. [63, 64, 65] and the tool’s website.^{9} We discuss the target final states of the relevant analyses in more detail below, alongside our results. Some final states have been reanalyzed and the corresponding bounds have been updated with new LHC results taken at \(\sqrt{s} = 13\) TeV centerofmass energy. For our purpose of comparing the relative exclusion power when going from an RPC to an RPV scenario, using the \(\sqrt{s} = 8\) TeV analysis set has the advantage of covering a much larger variety of final states.
To set the limit, CheckMATE tests all signal regions in all selected analyses, determines the one signal region with the largest expected sensitivity and checks if the corresponding observed result of that signal region is excluded at 95% C.L. or not. In the following, the analysis which contains this limitsetting signal region is referred to as the “most sensitive analysis” for a given \(\Lambda _{\not R_p}\)–CMSSM parameter point. Due to the lack of information about correlations in systematic uncertainties between different signal regions, CheckMATE is currently incapable of combining information from different signal regions.^{10}
6.1.4 Scanned parameter regions
6.2 Results
 1.
To what extent do the existing ATLAS and CMS analyses, which largely focus on RPC supersymmetry, exclude the parameter space of the \(\Lambda _{\not R_p}\)–CMSSM?
 2.Does breaking Rparity weaken the bounds of the RPC–CMSSM due to a gap in the coverage of possible final states?
 (a)
If the answer is yes, how could these gaps be closed?
 (b)
Alternatively, if the answer is no, in the cases where the bounds become stronger, which of the effects mentioned in previous sections lead to this result?
 (a)
6.2.1 Rparity conserving CMSSM
We start with a short discussion of the RPC–CMSSM in Fig. 11. On the left we denote in the \(M_0\)–\(M_{1/2}\) plane by the thick, solid black line the 95%CL exclusion range we obtained using CheckMATE for this model. Thus below the curve, in red is the excluded parameter area. Above the curve, in green is the allowed area. The remaining CMSSM parameters have been set to \(\tan \beta =30\), \(\mu > 0\) and \(A_0 =2 M_0\). In light gray we present supersymmetric mass isocurves for the LSP (dotdashed), the first two generation squarks (dotted), the gluino (solid), and the lightest stop (dashed). The white region on the far left at low \(M_0\) results in a \(\tilde{\tau }\)LSP, which is not viable phenomenologically if Rparity is conserved: as there is no possible decay channel for the stau, these regions result in stable charged particles which e.g. spoil big bang nucleosynthesis [195]. This is why here and in the following, we do not show any RPC results in the stauLSP region. We do include them in the RPV cases.
In the right plot, we show in the \(M_0\)–\(M_{1/2}\) plane, which analysis implemented in CheckMATE is most sensitive at a given parameter point or region. We use the color and hash code of Fig. 10. Thus for example the point (\(M_0=1000\,\)GeV, \(M_{1/2}=500\,\)GeV) is excluded by the analysis denoted atlas_1405_7875 [129] in CheckMATE.
The most sensitive analyses target either a 0lepton multijet (atlas_1405_7875), or \(a \ge 3b\)jet final state (atlas_conf_2013_061, Ref. [196]), both requiring a significant amount of missing transverse momentum. The former final state is especially sensitive when light gluinos decay into jets via onshell squarks and therefore – as can be seen in our results – covers the low \(M_{0}\) region where the squarks are relatively light. On the contrary, the latter targets stop and gluino pair production, the dominant modes for large \(M_0\), which can produce four bjets due to the resulting top quarks in the final decay chain.
6.2.2 \(LL\bar{E}\), \(\Lambda _{\not R_p}\)CMSSM
We now consider the case of a small nonzero \(LL\bar{E}\) operator as discussed in Sects. 3.2 and 4. We determine excluded parameter regions of the corresponding \(\Lambda _{\not R_p}\)–CMSSM and compare with the LHC exclusion line obtained in the RPC case. Since we consider a small RPV coupling, the particle spectrum remains virtually unchanged with respect to the RPC–CMSSM. However, as emphasized before, such an operator leads to the decay of the LSP. Therefore, a neutralino LSP will decay into two leptons and one neutrino for a generic \(\lambda _{ijk}\) coupling, cf. Eq. (18) and Table 3. In parameter regions where the lightest stau is the LSP its possible decay modes are: (i) directly into \(\ell _i \nu _j\), if either i, j or k equals 3, (ii) via the RGEgenerated \(\lambda _{i33}\) coupling if the nonzero RPV coupling at \(M_X\) is of form \(\lambda _{ijj}\) and \(\{i,j\}\ne 3\) (see Sect. 2.4) or (iii) via a fourbody decay, cf. see Sect. 4, and Table 9 for the corresponding LHC signatures. However, the fourbody decay does not happen here due to the large \(\tan \beta \) value employed and the consequential dominance of the twobody decay through the RGEgenerated operators.
The generic LHC searches for RPC supersymmetry look for missing energy in combination with jets and/or leptons. They should thus also perform well for the \(LL\bar{E}\) models, due to the many extra leptons from the RPV decay. Although the amount of missing energy is in general not as pronounced as for an RPC model, the energy carried away by the neutrino in the final decay can still be sizable enough to produce a striking signature; see Refs. [197, 198].
In the top row of Fig. 12 we show the CheckMATE exclusion in the \(M_0\)–\(M_{1/2}\) plane for the \(\Lambda _{\not R_p}\)–CMSSM with \(\lambda _{122}\ne 0\). The remaining CMSSM parameters and the light gray isomass curves are as in Fig. 11. The excluded region is shown below the thick solid black curve in red. The allowed region is shown above this curve in green. The dark red/green colored regions correspond to a \(\tilde{\tau }\)LSP, which must be considered in the RPV case. The light red/green colored regions correspond to a \(\tilde{\chi }^0_1\)LSP, as in the RPC case. The RPC–CMSSM exclusion line of Fig. 11 (thick solid black curve there) is shown here as a thick dashed black curve for comparison. It does not extend into the \(\tilde{\tau }\)LSP region, as that is not viable in the RPC–CMSSM.
The plot on the upper right in Fig. 12 shows in thick solid black the same CheckMATE exclusion from the upperleft plot, as well as the RPC exclusion from Fig. 11 as a solid dashed line. It furthermore shows which LHC analysis implemented in CheckMATE is most sensitive at a given parameter region using the same color code as in Fig. 11. When comparing with Fig. 10, we see that most of the parameter range is most sensitively covered by atlas_conf_2013_036, Ref. [56].
As a result of \(\lambda _{122}\ne 0\), the neutralino decays will lead to four more charged first or secondgeneration leptons compared to the RPC case, in regions where the neutralino is the LSP, cf. Table 3. Consequently, analyses looking for four or more leptons, Ref. [56], are very sensitive to this scenario and yield a stronger limit than in the RPC case. Thus the solid black curve in the upperleft plot is more restrictive than the dashed black curve. The search in Ref. [56] contains separate signal regions designed for both RPC and RPV signatures, respectively. It is interesting that, although their signal regions designed for the RPV signatures are the ones with the best exclusion power, the RPC signal region performs almost equally well.
When looking more closely at the CheckMATE output we see that it is a specific search region in Ref. [56], which is most sensitive to the \(LL\bar{E}\) case we are considering here, namely the electroweak pair production of neutralinos and charginos. This production channel is not very promising for RPC models in the CMSSM, as the largest electroweak cross section is usually obtained by the production of a charged and a neutral wino. Within the CMSSM boundary conditions, both would decay to the bino by emitting a W and a Higgs/Zboson, respectively. This comparably small electroweak signal rate is usually not enough for these final states to be detected over the background. Thus the RPC–CMSSM is most stringently constrained by gluino pair production. In contrast, in the \(\Lambda _{\not R_p}\)–CMSSM the neutralino decays via \(\lambda _{122}\) lead to a clean signal with many charged leptons.
Specifically, for \(M_{0}\gtrsim 500\,\hbox {GeV}\), corresponding to regions with a neutralino LSP, we can exclude values of \(M_{1/2} \lesssim 950\,\hbox {GeV}\), which feature a bino of \(m_{\tilde{\chi }^0_1}\simeq 400\,\hbox {GeV}\) and winos of \(m_{\tilde{\chi }^0_2}\simeq m_{\tilde{\chi }^\pm _1}\simeq 800\,\hbox {GeV}\), as well as gluinos over 2.1 TeV. Thus within the \(\Lambda _{\not R_p}\)CMSSM, the indirect constraint on the gluino mass (via the universal gaugino mass) is much stricter than the RPC gluino search can reach.
These \(\Lambda _{\not R_p}\)CMSSM bounds obtained using CheckMATE can, to some degree, be compared to the results from Ref. [110], where bounds on the pair production of winos decaying via \(LL\bar{E}\) are set. This analysis excludes wino masses up to 900 GeV also for \(\lambda _{122}\ne 0\), when assuming that the neutral wino is the LSP. In the case at hand we have a lighter bino to which the wino will decay. This change in kinematics (for instance, the finalstate leptons will be less energetic) with respect to the simplified model analysis in Ref. [110] explains the small differences observed in the bounds on the wino mass.
In the upperleft plot of Fig. 12, the low \(M_{0}\) region features a stau LSP and we observe that the exclusion power close to the LSPboundary drops significantly. This occurs due to the produced neutralinos (charginos) decaying into \(\tau \tilde{\tau }_1\) \((\nu \tilde{\tau }_1)\) and \(\tilde{\tau }_1\) decaying via the RGEinduced \(\lambda _{133}\) coupling. As a result the stau has equal branching ratios into both \(e\nu \) and \(\tau \nu \) final states. Compared to the expected signatures of neutralinoLSP regions explained above, several finalstate first and secondgeneration leptons are now replaced by \(\tau \) leptons. The expected event rates therefore drop by powers of the leptonic tau branching ratio and hence significantly affect the resulting bound from the same analysis.
As \(M_0\) approaches 0, the mass of all sleptons and sneutrinos further decreases. This slowly opens further decays of the neutralino (chargino) into other \(\ell _i \tilde{\ell }_i\) and \(\nu _i \tilde{\nu }_i\) (\( \nu _i \tilde{\ell }_i\) and \(\ell _i \tilde{\nu }_i\)) combinations, which for \(i \ne 3\) lead to the same decay signatures via \(\lambda _{122}\) as discussed before for the neutralino LSP region. Hence, the exclusion line approaches the earlier, stricter bound for \(M_0 \rightarrow 0\).
Next, we consider the bottom row of Fig. 12 with a nonzero \(\lambda _{123}\) coupling. The labeling and the included curves are to be understood as for the upper two plots. In the lowerright plot we see that again atlas_conf_2013_036, Ref. [56], is most sensitive over most of the parameter region, however only if \(M_0\) is large and the squarks are decoupled. For small \(M_0\), analyses which look for jets and likesign charged leptons, for example Ref. [133] (denoted atlas_1404_2500 in Fig. 10), become more sensitive. In fact, these two analyses are similarly sensitive in this parameter region but the second gets contributions from light squark decays and thus starts dominating the exclusion line for small \(M_0\).^{11} To the right in the lowerleft plot, the neutralinoLSP decay now always involves tau leptons, cf. Eq. (18). The resulting overall bound, the thick solid black curve, is weaker than in the \(\lambda _ {122}\) case in the neutralinoLSP region, as the multilepton signal is diluted by these taus. However, it is still stricter than the RPC case, the thick black dashed curve. We furthermore observe in the lowerleft plot that the considered analyses in the stau LSP region, i.e. for small \(M_0\), are more sensitive, than in the \(\lambda _{122}\) case. The search is more sensitive, as the stau now decays via \(\lambda _{123}\) into a neutrino and a first or secondgeneration charged lepton, rather than a tau final state as in the \(\lambda _{122}\) case, cf. the second line in Eq. (31).
For completeness we in turn show the results for nonzero values of \(\lambda _{131}\) and \(\lambda _{133}\) in Fig. 13, respectively. The notation is as in Fig. 12. Again, the differences with respect to Fig. 12 can be explained by, respectively, considering the number of charged first and secondgeneration charged leptons versus the number of tau leptons in the relevant final states. For the \(\lambda _{131}\) coupling, the neutralino will decay with almost equal branching ratios into both \(e e\nu \) and \(e\tau \nu \), which is why the excluded region is larger than for \(\lambda _{123}\) but smaller than for \(\lambda _{122}\). For the case of \(\lambda _{133}\), the neutralino decays either into \(\tau \tau \nu \) or \(e \tau \nu \), which is why the LHC sensitivity is lower compared to all previous cases. Remarkably, however, it is still more sensitive than the RPC case for most values of \(M_0\). Lastly, we note there are only minor differences between the exclusion lines if we were to exchange RPV couplings to (s)electrons by couplings to (s)muons which is due to the comparable identification efficiency between electrons and muons at both ATLAS and CMS. All other \(LL\bar{E}\) couplings are obtained by exchanging flavor indices \(1\leftrightarrow 2\) and the respective bounds can therefore be inferred from the scenarios shown above.
For all of the \(LL\bar{E}\) cases, we see that the \(4\ell \)+\({\mathrm {MET}}\) search of atlas_conf_2013_036, Ref. [56] and the jets\(+\)SS\(\ell \) search of atlas_1404_2500, Ref. [133], are the most sensitive in CheckMATE over most of the parameter range. The most important message from looking at the different \(LL\bar{E}\) operators and comparing to the RPC case is, however, that within the CMSSM, as a complete supersymmetric model, the LHC is actually more sensitive to scenarios in which Rparity is violated via an \(LL\bar{E}\) operator than if Rparity is conserved. This statement holds even if the signal regions designed for RPV in Ref. [56] are disregarded.
6.2.3 \(LQ\bar{D}\), \(\Lambda _{\not R_p}\)CMSSM
We now turn to the discussion of the \(LQ\bar{D}\) operator. In general, when compared to the previous \(LL\bar{E}\) case, it is clear that the LHC sensitivity is reduced, as we have to replace either a neutrino and a charged lepton or even two charged leptons by two quarks in the final RPV decay of a neutralino LSP; see Eq. (24) and Sect. 3.3. Stau LSPs, in turn, will mostly decay either into a pair of quarks or via a fourbody decay into a tau, a charged lepton or neutrino, and two quarks; see Eq. (31) and Sect. 4.
Let us discuss the results for the individual couplings. The first row in Fig. 14 shows the case of a nonzero \(\lambda '_{222}\) operator. As just mentioned, the overall exclusion sensitivity is significantly lower than in the \(LL\bar{E}\) case, and is comparable to the RPC–CMSSM case, shown here as the thick black dashed line. In a small region around \(M_0=750\,\)GeV, the RPC is even stricter. In the neutralinoLSP region with high \(M_0\), we find that analyses which look for jets and likesign charged leptons, for example Ref. [133] (denoted atlas_1404_2500 in Fig. 10), are most sensitive. See the righthand plot. In this region the first and secondgeneration squarks are relatively heavy. Therefore gluino and stop pair production are the most dominant production modes. These produce final states with many bjets, lower quark generation jets, and leptons and hence populate the 3b signal region of a “2 samesign \(\ell \) or \(3 \ell \,\)” analysis (atlas_1404_2500), for which the Standard Model background is nearly zero. Here, the high finalstate multiplicity induced by the \(LQ\bar{D}\) decay results in a slightly increased sensitivity when compared to the RPC case.
In regions with lower \(M_0\) where gluino–squark associated production and squark pair production become relevant, generic squark–gluino searches like atlas_conf_2013_062, Ref. [199], which look for jets, leptons and missing transverse momentum dominate. For these, the increased finalstate multiplicity via the additional \(LQ\bar{D}\)induced decays results in a worse bound than for the RPC case. This is due to the signal regions setting strong cuts on the required momentum of the finalstate objects and the missing transverse momentum of an event. These are necessary to sufficiently reduce the Standard Model background contribution, especially from multiboson production, which also produces final states with high jet and lepton multiplicity and some missing transverse momentum. Since the expected missing transverse momentum of the event is significantly larger in RPC models for which the LSP does not decay, breaking Rparity weakens the bounds in these regions, cf. Refs. [197, 198].
We continue with the discussion of \(\lambda '_{113}\), with the only phenomenologically relevant difference that \(\bar{D}_2\) is replaced by \(\bar{D}_3\). Hence, in the neutralino LSP case, the only phenomenological difference is that two bjets replace two normal jets. (We found that the sensitivity in the \(\lambda '_{222}\) and \(\lambda '_{112}\) cases are almost identical, since the experimental efficiencies for muons and electrons are similar.) Due to the good bjet tagging efficiency, this clearly improves the distinguishability with respect to the Standard Model background and results in an increase in sensitivity. This effect is most prominent for large values of \(M_0\). The same analysis as in the previous \(\lambda '_{222}\) case, see atlas_1404_2500, Ref. [133], provides the most stringent bounds as it contains special signal regions which tag additional bjets. For smaller values of \(M_0\) barely any change in sensitivity is visible in comparison to before.
In the \(\tilde{\tau }\)LSP region of the \(\lambda '_{113}\) case, the stau will almost always undergo a fourbody decay, thereby decaying into both \(\tau e b j\) and \(\tau \nu b j\) at approximately equal rates. The increase in sensitivity with respect to the neutralinoLSP region comes from the additional tau leptons in the final state.
We continue with the cases \(\lambda '_{131}\) and \(\lambda '_{133}\) in the lower two rows in Fig. 14 and focus on the neutralino LSP region first. Here, the top quark in the decay products does not improve the sensitivity when compared to the \(\lambda '_{222}\) case. When comparing to the \(\lambda '_{113}\) case, we see the sensitivity also goes down. On the one side we no longer have the bottom quark jet in every decay and on the other hand the operator \(\lambda '_{13i}\) in principle allows for neutralino decays into both \(t+\ell +j_i\) and \(b + \nu +j_i\). However, the mass of the LSP is so low in the relevant parameter range that the decay into the top quark is kinematically suppressed. Hence, most of the neutralinos will decay via the neutrino mode and as such do not produce the finalstate leptons which are required for the aforementioned “2 samesign \(\ell \) or \(3 \ell \,\)” analysis to be sensitive. Instead, the most relevant analysis in the large\(M_0\) region turns out to be a search looking for events with more than seven jets plus missing energy; see atlas_1308_1841, Ref. [128]. The high jet multiplicity for \(\lambda '_{131,133}\) arises from hadronically decaying tops, produced from the standard \(\tilde{g} \rightarrow t \tilde{t}, \tilde{t} \rightarrow t\tilde{\chi }^0_1\) decay chains in this parameter region, as well as jets from the final neutralino decay \(\tilde{\chi }^0_1\rightarrow jj\not \!\!E_T\).
Comparing \(\lambda '_{133}\) to \(\lambda '_{131}\), the two additional bjets in the final state result in a slightly improved exclusion power via the multib analysis in atlas_conf_2013_061, Ref. [196].
In the \(\tilde{\tau }\)LSP region, the case \(\lambda '_{133}\) is analogous to the \(\lambda '_{222}\) case in that the RGEgenerated \(\lambda _ {133}\) operator determines the \(\tilde{\tau }\) decay, leading to similar bounds. In the \(\lambda '_{131}\) case, the situation is similar to \(\lambda '_{113}\ne 0\) in that the fourbody decay dominates. However, as the final state including the top quark, \(\tilde{\tau } \rightarrow \tau e t j\), is kinematically suppressed, the stau almost exclusively decays into \(\tau \nu b j\). This scenario therefore exhibits the worst LHC measurement prospects of all the \(\lambda '_{aij} \ne 0\), \(a=1,2\), \(\tilde{\tau }\)LSP scenarios.
Turning to the \(\lambda '_{311}\) scenario shown in the top row of Fig. 15, all differences with respect to the former \(\lambda ^\prime _{222}\) case can be explained by the exchange of muons by taus in the final state, which reduces the overall final state identification efficiency. As a result, the searches looking for leptons lose sensitivity and, similar to the above \(\lambda ^\prime _{i3i}\) cases, the best constraints are instead provided by the high jet multiplicity analysis described in atlas_1308_1841, Ref. [128]. Whilst in the \(\lambda ^\prime _{222}\) scenario the lower \(M_0\) region was most constrained by the squark–gluino searches in atlas_conf_2013_062, Ref. [199], here this region is again covered by the high multiplicity jet analyses. A closer look at the event rates, however, reveals that these two analyses are almost equally sensitive and hence the resulting bounds are nearly the same.
In the \(\lambda '_{311}\) scenario, the LHC sensitivity does not change significantly when traversing from the neutralino into the stauLSP region since the stau itself decays directly into light quark jets. Hence, only the kinematics change when the LSP crossover occurs, while the finalstate signatures stay the same. This is why we see exactly the same behavior for the other \(i=3\) cases \(\lambda '_{313}\), \(\lambda '_{331}\) and \(\lambda '_{333}\ne 0\), the former of which we show in the second row of Fig. 15. Consequently, the additional btagging in these scenarios does not noticeably improve the exclusion power of atlas_1308_1841, Ref. [128].
Here we have considered all distinct types of nonzero \(LQ\bar{D}\) operators, which in principle have differing LHC phenomenology. In the region where the neutralino is the LSP and \(M_0\lesssim 1.2~\)TeV, the corresponding LHC bounds that we obtain using CheckMATE are slightly weaker compared to the Rparityconserving CMSSM. This corresponds to the region where squark pair production dominates and the additional decay of the neutralino LSP reduces the \(\not \!\!E_T\). In the parameter region where gluino and stop pair production dominates, i.e. for large \(M_0\), we instead find most \(LQ \bar{D}\) scenarios are as constrained as the RPC–CMSSM because of the equally good performance of the multijet searches preferred by RPV and the multib searches sensitive to RPC. The special cases \(L_i Q_j \bar{D}_3\) with \(i, j \in \{1, 2\}\) are significantly more constrained in this region of the Rparityviolating CMSSM, due to the additional extra leptons and bjets in the final state. In all cases, regions with stau LSP are well covered by either multilepton or combined lepton+jet searches and yield comparable bounds as in parameter regions with a neutralino LSP.
6.2.4 \(\bar{U} \bar{D} \bar{D}\)
Here we discuss the \(\bar{U} \bar{D} \bar{D}\) operator for which one typically expects the weakest LHC bounds as there is no striking missing energy signal nor any additional leptons; see for example Refs. [169, 200, 201].
In Fig. 16, we show the results in analogy with the previous subsections. We first consider the case of \(\lambda ''_{121}\), the top row, where the neutralino LSP decays into three light jets. Therefore multijet searches should yield the most stringent limits for such scenarios. Indeed, as can be seen in the top right plot of Fig. 16, the analysis in atlas_1308_1841, Ref. [128] provides the best exclusion power for the entire neutralinoLSP region. Interestingly, the bounds on the parameter space which we obtain are almost as strong as the bounds on the RPC scenario, the thick black dashed line, cf. Fig. 11. This can be regarded as an impressive success for the experimental groups, since multijet analyses belong to the most challenging signatures at a hadron collider.
In the large \(M_0\) region where the exclusion lines from RPC and RPV are very similar, gluino pair production has the highest cross section. The gluinos then decay down to a top quark and a stop which itself decays to a top and a neutralino. The dominant \(\tilde{g} \rightarrow t \tilde{t}\) decay occurs because of the large stop mixing in this region which significantly reduces the \(\tilde{t}_1\) mass with respect to the other squark masses. The neutralino then eventually undergoes a threebody decay into three light jets. As a result, the signal region looking for \(\ge 10\) jets and missing energy (denoted “10j50” in Ref. [128]) provides the best constraints. This is somewhat surprising as the analysis vetoes against isolated leptons while requiring missing energy. Naively, one would have expected searches for bjets, missing energy and leptons to dominate. However, we find that only the nextbest analysis looks for that, Ref. [202], with the best applicable signal region “SR1\(\ell \)6jC” looking for one lepton, more than six jets and missing energy. Furthermore, note that also in Ref. [128], a RPV interpretation has been performed, assuming gluino pair production, which decay into \(\tilde{t} \bar{t}\), with \(\tilde{t}\rightarrow bs\), obtaining bounds of \(m_{\tilde{g}}\gtrsim 1\,\)TeV. Translating the bounds we obtain for the CMSSMlike scenario to gluino mass bounds, we obtain even stricter mass limits, which is due to the higher jet multiplicity in the final state from including the intermediate neutralino in the decay chain.
In the lower \(M_0\) regions where the exclusion in RPC parameter space is stricter, squark–gluino associated production is dominant. Therefore, while the RPC–CMSSM provides a large missing energy signal and is therefore probed by analyses like atlas_1405_7875, Ref. [129], the RPV counterpart is still best covered by the 10jet signal region of [128].
We once more want to emphasize that the bounds we obtain on the parameter space rely on the boundary conditions which we impose at the high scale and the corresponding (s)particle spectrum. In particular, as seen in the large \(M_0\) region, the presence of top quarks in the decay is of major importance. In Ref. [203], a reinterpretation of LHC results in a natural SUSY context and \(\lambda _{212}'' \ne 0\) has been performed. In their scenario the stop and gluino masses are varied independently and a Higgsino LSP decaying into three light jets is assumed. The results show that in this case, gluino and stop masses are generically less constrained when compared to the RPC analog.
In the stau LSP region, at low \(M_0\), the decay into one tau lepton and three light jets dominates for \(\lambda ''_{121}\) and into one tau, one bjet and two light jets for \(\lambda ''_{113}\). Interestingly, most of this area is best covered by the 4lepton analysis atlas_conf_2013_036 [56], which requires, in the case of a \(\tau \)tag, at least three additional light leptons. This means that the other three \(\tau \)leptons can only be identified through their leptonic decay, reducing the overall acceptance by a factor \([{\mathrm{BR}}(\tau \rightarrow \ell \nu \nu )]^3 \simeq 0.044\). In addition, we find that wino pair production is important in this part of parameter space and that the charged wino state decays into \(e/\mu +\tilde{\nu }\) in up to 30% of the cases, further contributing a charged lepton in the final state.
Turning to the cases where thirdgeneration quarks are among the LSP decays, namely the bottom three rows in Fig. 16, we see that the bounds in the low \(M_0\) region are similar to the \(\lambda ''_{121}\) case, and that again the multijet search is most sensitive. In the high \(M_0\) region, where gluino and stop pair production becomes relevant, searches for samesign leptons, atlas_1404_2500 [133], become effective for \(\lambda ''_{3ij}\). This is due to leptonically decaying top quarks in the final state and has already been analyzed in detail in Ref. [204]. Note that in comparison to the \(LQ\bar{D}\) operator \(\lambda '_{i3j}\), where the top in the final state was phasespace suppressed and as such the alternative neutrino decay mode was favored, there is no other comparable decay mode for the neutralino in the case of \(\bar{U} \bar{D} \bar{D}\). As such, it will always decay into a top quark whenever kinematically accessible, which is the case for \(M_{1/2}\gtrsim 400~\)GeV. Therefore, both gluino and stop pair production can lead to samesign leptons from the leptonic top decay modes, rendering this scenario slightly more constrained than the \(\lambda ''_{aij}\), \(a=1,2\), cases. Interestingly, we find that in addition electroweak gaugino production is even more important for the limit setting in the large\(M_0\) area than stop pair production.
For the \(\lambda ''_{323}\) case, the additional possibility of tagging more bjets further improves sensitivity, such that the large\(M_0\) region is considerably more constrained than the RPC analog.
For \(\lambda ''_{312}\), in regions with the stau being the LSP, the kinematical suppression of final states with top quarks results in the most abundant decay chains going via offshell charginos into a neutrino, a bottom quark and two lightflavor jets, cf. Eq. (34). Therefore, searches for missing energy and several jets, e.g. atlas_1308_1841 and atlas_1405_7875, Refs. [128, 129], provide a good coverage. At very low \(M_0\), where the mass difference \(m_{\tilde{\chi }^0_1}  m_{\tilde{\tau }_1}\) is largest, even the search for samesign leptons, atlas_1404_2500 [133], which is sensitive to the leptonically decaying taus from \(\tilde{\chi }^0 \rightarrow \tilde{\tau } \tau \), becomes effective enough to exclude the area below \(M_{1/2}\lesssim 730~ \)GeV. However, the sensitivity of this analysis to the scenario at hand quickly drops off with decreasing \(m_{\tilde{\chi }^0_1}m_ {\tilde{\tau }_1}\), as can be seen in the \(\lambda ''_{312}\) case of Fig. 16. For \(\lambda ''_{323} \ne 0\), which features at least four bjets in the final state, the search for large missing transverse momentum and at least three bjets, atlas_conf_2013_061 [196], is furthermore able to exclude the rest of the \(\tilde{\tau }\)LSP parameter space below around \(M_{1/2}\simeq 760~\)GeV.
Summarizing, \(\bar{U} \bar{D} \bar{D}\) couplings within the CMSSM are almost as well covered by LHC analyses as the RPC counterpart. Similarly to the \(LQ\bar{D}\) case, regions with low \(M_0\) are harder to detect at the LHC than the RPC scenario. For large \(M_0\) the searches for many jets and missing energy are very sensitive, leading to bounds as strong as in the RPC–CMSSM, while in the case of a \(\lambda ''_{3i3}\) coupling the bounds are even stricter. We stress again that these results are, in particular in the large \(M_0\) region, specific to the CMSSM boundary conditions. For instance, if the stops were heavier than the gluinos, the bounds which one could set on the corresponding scenario would be considerably weaker [203]. In the stau LSP region, searches for several leptons provide the best constraints whereas for \(\lambda ''_{323}\), multibjet analyses are even more sensitive.
7 Large \(\Lambda _{\not R_p}\): other LSP scenarios
Here, we briefly comment on scenarios with a large \(\Lambda _{\not R_p}\) coupling, as discussed in Sect. 5. In this case, the mass spectrum changes with respect to the RPC–CMSSM so that a direct comparison is no longer meaningful. However, qualitative changes only occur if not only the spectrum but also the relative hierarchy of particle masses is altered. This can lead to (i) changes in the finalstate signature and/or (ii) changes in the kinematic distributions.
A drastic example of case (i) is the squark LSP scenario, which we envisage for large \(\bar{U} \bar{D} \bar{D}\) couplings, cf. Table 2. Then, as discussed in Sect. 5, squark LSPs will be pairproduced and decay directly into pairs of dijet resonances via the (large) \(\lambda ''\) coupling. Even though we operate within CMSSM boundary conditions here, the possible squark LSP scenarios correspond quite closely to the simplified models employed in the experimental analyses searching for this exact scenario, as the pair production of the comparably light squarks will dominate over all other production modes. We therefore refer to the analyses summarized in Sect. 5 for the respective bounds on the squark masses.
We cannot perform a recast of the bounds on the squark LSP scenarios with the current version of CheckMATE since the respective analyses containing 4jet final states of which two combine to a dijet resonance are not yet handled by this tool. We leave the inclusion of these results for future work.
In the case of a slepton or sneutrino LSP, the change in the finalstate signature is milder as the pair production of the LSP is suppressed with respect to squark/gluino and also wino production. There will therefore at most be changes in the intermediate cascade decays as well as the kinematics of the finalstate particles; we do not expect drastic changes.
It is nevertheless instructive to check this statement with our tools at hand. In Fig. 17 we show an example of parameter space with either neutralino or sneutrino LSP, corresponding to the scenario shown earlier in Fig. 3. In the top plot on the left we have turned off the Rparityviolating coupling, \(\lambda '_{233}=0\), and show only the \(\tilde{\chi }^0_1\)LSP region, as is appropriate in the RPC–CMSSM. This agrees with Fig. 11. On the right we see that the most sensitive signature in the exclusion region, the lowerrighthand corner of the plot, is the \(0\ell \), 2–6j+\(\not \!\!E_T\) search of atlas_1405_7875 [129].
Coupling  \(\tilde{\chi }^0_1\) LSP region  \(\tilde{\tau }_1\) LSP region  

\(m_{\tilde{g}}\)  \(m_{\tilde{t}_1}\)  \(m_{\tilde{q}_{\mathrm{1st/2nd}}}\)  \(m_{\tilde{\chi }^0_1}\)  \(m_{\tilde{\chi }^\pm _1}\)  \(m_{\tilde{g}}\)  \(m_{\tilde{t}_1}\)  \(m_{\tilde{q}_{\mathrm{1st/2nd}}}\)  \(m_{\tilde{\chi }^0_1}\)  \(m_{\tilde{\chi }^\pm _1}\)  \(m_{\tilde{\tau }_1}\)  
RPC  1280  710  1560  220  430  –  –  –  –  –  – 
\(\lambda _{122}\)  2070  1320  1960  400  750  1690  1140  1520  320  600  230 
\(\lambda _{123}\)  1700  980  1630  310  600  1790  1220  1620  340  640  260 
\(\lambda _{131}\)  1850  1120  1700  350  670  1740  1180  1580  330  620  260 
\(\lambda _{133}\)  1590  920  1540  290  560  1690  1140  1520  320  600  230 
\(\lambda '_{111}\)  1220  700  1520  210  410  1690  1140  1520  320  600  230 
\(\lambda '_{113}\)  1480  850  1530  260  510  1690  1140  1520  320  600  230 
\(\lambda '_{131}\)  1310  750  1450  230  440  1690  1150  1520  320  600  220 
\(\lambda '_{133}\)  1310  750  1470  220  440  1690  1140  1520  320  600  230 
\(\lambda '_{311}\)  1250  750  1400  210  420  1530  1040  1360  280  530  190 
\(\lambda '_{313}\)  1290  730  1410  220  440  1530  1040  1360  280  530  190 
\(\lambda '_{323}\)  1280  720  1400  220  430  1530  1040  1370  280  540  200 
\(\lambda '_{331}\)  1330  750  1440  230  450  1580  1080  1420  290  560  210 
\(\lambda '_{333}\)  1350  770  1420  240  470  1620  1060  1460  310  600  240 
\(\lambda ''_{113}\)  1250  720  1350  210  420  1420  970  1270  260  490  180 
\(\lambda ''_{121}\)  1260  730  1350  210  420  1480  1010  1330  270  520  200 
\(\lambda ''_{312}\)  1250  730  1350  210  420  1430  960  1290  260  500  180 
\(\lambda ''_{323}\)  1400  780  1350  250  480  1530  1040  1360  280  530  190 
In the RPV case, we now want to compare the bounds on \(M_{1/2}\) in the sneutrinoLSP and the neutralinoLSP region, respectively. In both cases, the production of stop squarks dominates the LHC supersymmetric production cross section. Despite the large \(\lambda '_{233}\) coupling invoked, the stop mainly decays into a top quark and a bino. In the region on the righthand side of the figure, the binoLSP has a dominant threebody decay. While in the lefthand region, the bino is not the LSP and it first decays into \(\nu _\mu \tilde{\nu }_\mu \) or \(\mu \tilde{\mu }\), with the nearly degenerate \(\tilde{\nu }\) or \(\tilde{\mu }\) onshell. The latter then decay further via the \(\lambda '\) coupling. The dominant final state is therefore the same in the \(\tilde{\chi }^0_1\)LSP and the \(\tilde{\nu }\)LSP scenarios, and only the finalstate particles’ kinematics differ. As expected, we thus observe that, except for a small dip in the crossover region (\(m_{\tilde{\chi }^0_1}m_{\tilde{\nu }_1}\) is small) the bounds in the sneutrinoLSP and the neutralinoLSP region are comparable.
Interestingly, in some of the sneutrinoLSP region, the cross section of the pair production of sneutrinos and smuons is even comparable to the stop pair production. However, due to the additional top quark in the final state, the latter provides a better discrimination against Standard Model background and sneutrino/smuon pair production does not provide any mentionable constraints on the parameter space by itself.
From this, we conclude that even though the mass hierarchies of the lightest supersymmetric particles may be affected for larger RPV couplings, the resulting bounds are hardly dependent on the details of this hierarchy, as long as both LSP and NLSP are only electroweakly interacting. Small, fine tuned parameter regions with degenerate LSPNLSP masses form a mild exception as here the soft decay kinematics of the NLSPtoLSP decay weaken the resulting bounds.
8 Absolute lower mass bounds on RPVCMSSM scenarios
In this last section, we present a set of lower supersymmetric mass bounds within the CMSSM. These thus assume a complete supersymmetric model, with possibly involved cascadedecay chains. This is unlike the experimental bounds in Tables 4, 6 and 8, which are based on simplified models.
The bounds here in Table 12 are the result of the analyses of Sect. 6, which lead to the Figs. 11, 12, 13, 14, 15 and 16. For each case the allowed parameter range (green) is scanned and the lightest respective sparticle mass is determined. We list separately the bounds for the case of a neutralino LSP (left) and for a stau LSP (right). In both cases we give the lower mass bounds for: the gluino, the lightest stop, the first/secondgeneration squarks, the lightest neutralino and the lightest chargino. These are the particles which are also directly produced. Bounds on other particles also exist, but these are always indirect, and obtained only through the CMSSM boundary conditions. They inform us about the \(\Lambda _{\not R_p}\)–CMSSM, not necessarily the sensitivity of the LHC. For the stau LSP scenario we include the lower bound on the lightest stau, as this is an essential parameter of these models. We emphasize that all the bounds in the stauLSP case are new, as such bounds do not yet exist in the literature.
Looking at the bounds more closely, for example in the upperleft plot of Fig. 12 for \(\lambda _{122}\not =0\), we would expect the lightest allowed gluino mass in the \(\tilde{\chi }^0_1\)LSP case to correspond to the dip in the exclusion curve near \((M_0,\,M_{1/2})\simeq (800\,{\mathrm {GeV}},\,920\,{\mathrm {GeV}})\). Looking at the light gray dotdashed gluino mass isocurve, we therefore expect a lower mass exclusion bound of just over 2000 GeV. In Table 12 in the row for \(\lambda _{122}\), we see the lower gluino mass bound is indeed 2070 GeV. Correspondingly in the lowerleft plot on the left in Fig. 12, for \(\lambda _{123}\not =0\), the exclusion curve is significantly lower, about three quarters of the way between the 1 and 2 TeV gluino isomass curves. It also has no marked dip. The bound in Table 12 is 1700 GeV. In these cases, the strong bounds on the gluino mass are caused by direct bounds on the electroweak gaugino sector through multilepton searches which translate into bounds on the gluino mass via the CMSSM boundary conditions. Since the cross section for gluino production at such high masses is of order \(\mathcal {O}(\)ab), direct measurements of gluinoinduced topologies cannot provide competetive bounds.
Turning to the case \(\lambda '_{113}\not =0\) shown in Fig. 14, we see the exclusion curve sloping downwards for large \(M_0\), thus the bound is obtained at the limit of our scan region, \(M_0=3000\,\)GeV. This is similar to the extended simplified models considered in Ref. [112], where both the squarks and the gluinos were kinematically accessible. The lower mass bounds on the gluino/squarks shown in the second to last row of the gluino section and the last row of the squark section in Table 6 strongly depend upon the chosen squark/gluino masses, i.e. the gluino mass bound depends upon the assumed or allowed squark masses and vice versa.
The remaining gluino mass bounds in the \(\tilde{\chi }^0_1\)LSP scenario for \(\lambda '\) and \(\lambda ''\) are all very similar, mainly around 1300 GeV. They are determined by the limit of the scanning region and rest on the production cross sections at \(M_0=3000\,\) GeV. Contrarily to the above \(LL\bar{E}\) discussion, the most sensitive signatures require the production of gluinos and hence set a comparably weaker bound. In all these cases a lighter gluino should be possible for completely decoupled squarks.
The lower mass bounds on the lightest top squark are typically significantly weaker, in the range of 700–800 GeV over all couplings in the \(\tilde{\chi }^0_1\)LSP scenario. This is similar to the RPC CMSSM bound of 710 GeV. The exception are the cases \(\lambda _{122}\, (\lambda _{131})\) with \(m_{\tilde{t}_1}> 1320\) GeV (1120 GeV). The reason is that, because of the employed relation \(A_0=2\,M_0\), the stop mass splitting increases with increasing \(M_0\). Thus the lowest stop mass bounds come from the bounds at large \(M_0\). As is seen in Figs. 12 and 13, the cases \(\lambda _{122,131}\) are the only ones where the lower bound on \(M_{1/2}\) is as severe or even stricter for large \(M_0\) as for lower \(M_0\). For the first/secondgeneration squarks the lower mass bounds come from the low\(M_0\) region and are consistently around 1400–1600 GeV, as in the RPC case. They are only markedly stricter in the \(LL\bar{E}\) scenarios.
For the stau LSP scenarios the lower gluino mass bound is typically obtained in the dip region along the neutralinoLSP–stauLSP crossover or close by. For \(\lambda _{122}\) this is clearly significantly lower, for \(\lambda _{123,131,133}\) it is comparable to the neutralinoLSP case and for \(\lambda '_{111,113,131,133,311,313},\) \(\lambda ''_{121,113,312,323}\) the lower gluino mass bound in the stauLSP case should be considerably stricter than in the neutralino LSP case. This is confirmed in Table 12. For the \(LL\bar{E}\) cases these gluino mass lower bounds are due to electroweak gaugino production. For \(LQ\bar{D}\) and \(\bar{U}\bar{D}\bar{D}\) the production process leading to the most sensitive limits is gluino and/or squark production, possibly involving top squarks. The other mass bounds are then determined indirectly via the CMSSM boundary conditions. Thus the gluino and lightest neutralino mass bounds can be roughly understood as the mass ratio \(M_1/M_3\simeq 1/6\). The chargino mass bounds are also due to direct electroweak gaugino production for the \(LL\bar{E}\) case but are otherwise also derived quantities in the RPVCMSSM. The stop and lightest generation squark bounds are a mixture, sometimes derived, but sometimes also obtained via direct production. The stau mass bounds in the stauLSP scenarios are all derived quantities. Overall the stauLSP parameter range is very narrow and thus the lower mass bounds are very similar across all couplings.
To emphasize these bounds are the result of using CheckMATE and therefore contain all the same deficiencies as discussed in Sect. 6.1.2. Also, experimental bounds are typically interpreted within simplified models whereas CMSSMbased scenarios like ours have various potentially interesting decay signatures which appear simultaneously. Lacking a statistical combination of the numerous search channels then leads to a significant dilution of the bounds that can be derived compared to a single simplified model.
Despite being conservative our RPVCMSSM gluino mass bounds are stricter in the \(LL\bar{E}\) \(\tilde{\chi }^0_1\)LSP case than in Table 4. This is because they are in fact due to electroweak gaugino production, which can be reinterpreted within the \(\Lambda _{\not R_p}\)CMSSM. In the \(LQ\bar{D}\) case the bounds are largely similar, except for those from Refs. [118, 119] which are based on \(\sqrt{s}=13\,\)TeV data. In the \(\bar{U}\bar{D}\bar{D}\) case the bounds are similar.
9 Summary
 1.
In Sect. 2, starting from the small set of \(\Lambda _{\not R_p}\)CMSSM parameters at \(M_X\) in Eq. (5), we have dynamically determined the possible LSPs at the weak scale, taking in particular the Higgs mass constraint into account. This is an update of Ref. [62]. The results are presented in Table 2. We find an extensive parameter range with either a neutralino or a stau LSP. For special large RPV couplings we can also have one of \(\{\tilde{e}_R,\,\tilde{\mu }_R,\,\tilde{\nu }_{e,\mu },\,\tilde{s}_R,\,\tilde{d}_R,\,\tilde{b}_1,\,\tilde{t}_1\}\) as the LSP.
 2a.In Sect. 3, we focused first on the \(\tilde{\chi }^0_1\)LSP scenarios. For the various possible dominant operators, we have compiled tables detailing all possible LHC signatures: Tables 3 (\(LL\bar{E}\)), 5 (\(LQ\bar{D}\)), and 7 (\(\bar{U}\bar{D}\bar{D}\)). We have then compiled all relevant LHC analyses by ATLAS and CMS, and we have presented the resulting bounds on the simplified supersymmetric mass spectra in Tables 4 (\(LL\bar{E}\)), 6 (\(LQ\bar{D}\)), and 8 (\(\bar{U}\bar{D}\bar{D}\)), again depending on the nature of the dominant RPV operator. These bounds are independent of the assumption of CMSSMlike boundary constraints and can thus be applied to all RPV models, provided the appropriate branching ratios are implemented. Comparing the two sets of tables we can thus determine the coverage of these models at the LHC. We have observed the following:

\(\tilde{\chi }^0_1\)LSP, \(L L \bar{E}\): These scenarios are very well covered via LHC analyses looking for four leptons (including a number of taus) plus missing transverse energy, cf. Ref. [107]. We also note that electroweak gaugino production can play an important role due to the large number of additional leptons in the finals states, significantly boosting the efficiencies.

\(\tilde{\chi }^0_1\)LSP, \(L Q \bar{D}\): The typical signatures containing charged leptons with a number of jets, btagged or otherwise, are well covered. However, final states lacking charged leptons and instead containing hadronically decaying taus and or missing transverse energy are not completely covered, cf. cases IIgk in Table 5. Most existing analyses focus on high jet multiplicity plus missing transverse energy which provide some sensitivity to the above scenarios. Note the recent analysis in Ref. [119] tags an isolated electron or muon, further requiring high jet multiplicity with no veto on missing transverse energy. This search is sensitive to many of the RPV scenarios beyond just the \(L Q \bar{D}\) operators.

\(\tilde{\chi }^0_1\)LSP, \(U\bar{D} \bar{D}\): Many scenarios are well covered, especially in the case that top quarks are produced in the cascadedecay chain, yielding leptons in the final states. Searches for only jets with and without bjets are in principle sensitive to all possible final states. However in Ref. [57] all \(\bar{U}\bar{D} \bar{D}\) couplings where simultaneously switched on. This makes it very difficult to reinterpret the particular analysis in comparison to single operator dominance. Finally based on the simplified analyses with single operator dominance, \(\lambda ''_{312}\) coverage is lacking.

 2b.
In Sect. 4, we next considered the stauLSP scenarios in detail. The list of LHC signatures has been presented in Ref. [96]. However, these models, irrespective of the dominant RPV coupling, are not explicitly searched for at the LHC. The only exception is a CMSSM model with nonzero \(\lambda _{121}\) [106]. This analysis only uses \(\sqrt{s}=7\,\)TeV data and makes a number of assumptions about the fourbody decay channels, completely ignoring twobody decay channels, which become relevant for large \(\tan \beta \). We thus do not present a list of lower mass bounds on the supersymmetric particles in these scenarios. We are, however, encouraged by the recent search for multiple leptons with up to two hadronic tau candidates, motivated by electroweak gaugino production in Rparityconserving models [147]. These signatures are also highly relevant for many stauLSP scenarios. We further encourage experimentalists to perform dedicated analyses looking for final states with high tau, charged lepton and jet multiplicities, cf. Table 9.
 2c.
In Sect. 5 we summarize the experimental LHC bounds on the nonstandard LSP scenarios, i.e. those listed in Table 2, which are obtained for large RPV couplings. The results are summarized in Tables 10 and 11. There are typically no direct searches for these scenarios at the LHC, except in the \(\tilde{t}_1\)LSP scenario. However in most cases the \(\tilde{\chi }^0_1\) is the NLSP and for example the gluino cascade decay proceeds through the same chain as in the corresponding RPV \(\tilde{\chi }^0_1\)LSP case. The only difference is that the neutralino decay is now twobody instead of threebody. Thus the finalstate kinematic distributions should be slightly different. We expect this to only moderately affect the search sensitivities. In special cases entire decay modes can be kinematically blocked due to the heavy top quark for example, which in turn can affect bounds more significantly. We have also collected a set of related searches which involve nonneutralino/stau LSPs, which do not arise in the \(\Lambda _{\not R_p}\)–CMSSM, but for which there are experimental searches. Here we have a sneutrino LSP, a squark LSP, and a gluino LSP. These are compiled in Table 11. Here we have also included reinterpreted leptoquark searches.
 3.
In Sect. 6 we performed collider studies using the program CheckMATE to assess the coverage at the LHC of RPV models in comparison to the CMSSM with Rparity conserved. We consistently only use analyses implemented in CheckMATE for the \(\sqrt{s}=8\,\)TeV data. We also only considered all supersymmetric production cross sections at leading order for both the Rparityviolating and the Rparityconserving case, multiplied with Kfactors determined by NLLFast for strongly produced final states. We found that the LHC constraints on RPV models are, for most regions of parameter space, at least comparable to the RPC case, while for the \(LL\bar{E}\) operator the constraints are significantly stronger (cf. Fig. 12). The main caveats are \(\bar{U} \bar{D} \bar{D}\) and \(L Q \bar{D}\) operators with \(M_0\) in the range 300–1000 GeV, cf. Figs. 14, 15 and 16. In the RPC case, the most sensitive analyses in these regions are searches looking for high\(p_T\) jets plus missing transverse momentum. Including these RPV operators decreases sensitivity through both the reduction of the missing transverse energy and the distribution of the jet \(p_T\) over many jets. Therefore searches for many jets and a moderate amount of missing energy are usually most sensitive here, but not competitive to the RPC sensitivity. One should note that the analysis in Ref. [119] is not currently available in CheckMATE. We expect this search to be far more sensitive in these parameter regions, especially as it does not trigger on missing transverse energy. The increased sensitivity outside of this \(M_0\) range occurs as many RPV operators can lead to the production of particles in the final state which are not only easier to detect experimentally but can also lead to greatly reduced SM background contamination. For example \(\bar{U}\bar{D} \bar{D}\) operators which involve couplings to the top squarks, \(\lambda ''_{3ij}\), can lead to final states with jets (including bjets) and two likesign leptons or even three or more leptons, which is what analyses like Ref. [133] have been designed for. Because of these rather special lepton signatures, these final states can be better discriminated against the SM background compared to the typical signatures arising from the RPCCMSSM in the large\(M_0\) region, which are one lepton, bjets and missing transverse momentum. We stress that we have used all available 8 TeV analyses implemented in CheckMATE, however, this does not yet include a large number of ATLAS and CMS analyses optimized for RPV models. We find that many of these missing analyses are essential when one considers RPV couplings which are large enough to directly affect the mass spectrum of the model. Since CheckMATE is currently restricted to cutandcountbased analyses, no resonance searches could be considered which might also be relevant for certain RPV scenarios, as discussed in the main text.
 4.
In Sect. 8 we have collected the resulting mass bounds of the RPV–CMSSM CheckMATE analysis of Sect. 6. We present explicit lower mass bounds for the gluino, the first/secondgeneration squarks, the lightest stop, the lightest neutralino and the lightest chargino. We present these separately for the case of a \(\tilde{\chi }^0_1\)LSP and for a \(\tilde{\tau }\)LSP. We compare the bounds with the corresponding RPCbounds obtained also with CheckMATE, as well as with the experimental bounds collected here in Sect. 3. The lower mass bounds in the stauLSP case are all new, as both ATLAS and CMS have not yet determined any lower mass bounds in this case.
Footnotes
 1.
See also Ref. [7] on the early history of supersymmetry, 1967–1976, and references therein.
 2.
This large value of the RPV coupling is consistent with the lowenergy bounds, cf. Table 1 and Appendix A.
 3.
 4.
 5.
As discussed above, for example an operator \(L_1L_2\bar{E}_1\) can generate at 1loop \(L_2Q_j \bar{D}_k\) operators. The corresponding neutralino decay branching ratios are however suppressed, as there is no kinematic or other suppression of the leading operators.
 6.
Here again: \(a,b,c \in \{1,2\}\) and \(i,j,k\in \{1,2,3\}\).
 7.
In the case of \(LL\bar{E}\) operators, however, the most stringent bounds may arise from electroweakino pair production, as we shall see in Sect. 6.
 8.
For the only benchmark case with large \(\Lambda _{\not R_p}\) we consider, the squark sector is not degenerate. However, as in this case stop production dominates, we can still use the NLLFast Kfactors for rescaling the cross sections.
 9.
 10.
The statistical combination of signal regions for CheckMATE is work in progress [193].
 11.
Atlas_1404_2500 only becomes the most sensitive analysis if higherorder cross sections are used for the strong production modes – at leading order the electroweak processes are dominant setting negligibly weaker bounds via analysis atlas_conf_2013_036.
Notes
Acknowledgements
We thank Santiago Folgueras for interesting discussions about multilepton searches at CMS, as well as Florian Staub for technical support with SARAH. One of us, H.K.D. thanks the organizers of the conference ‘Is SUSY Alive and Well?’ held at the IFT UA Madrid in Sept. 2016, which partially stimulated this work. We also thank Howie Haber, Steve Martin and Tim Stefaniak for discussions. D.D. acknowledges the support of the Collaborative Research Center SFB 676 “Particles, Strings and the Early Universe” of the DFG. M.E.K. thanks the DFG for financial support through the Research Unit 2239 “New Physics at the LHC” and the IFIC Valencia for hospitality while part of this work was completed. T.O. is supported by the SFB–Transregio TR33 “The Dark Universe”. A.R. thanks the Cusanuswerk for funding, and Tel Aviv University and the Weizmann Institute for hospitality while part of this work was completed. H.K.D. thanks Nikhef and MITP Mainz for hospitality while part of this work was completed.
References
 1.Yu A Golfand, E.P. Likhtman, JETP Lett. 13, 323 (1971). [Pisma Zh. Eksp. Teor. Fiz. 13, 452 (1971)]Google Scholar
 2.D.V. Volkov, V.P. Akulov, Phys. Lett. B 46, 109 (1973)ADSCrossRefGoogle Scholar
 3.J. Wess, B. Zumino, Phys. Lett. B 49, 52 (1974)ADSCrossRefGoogle Scholar
 4.J. Wess, B. Zumino, Nucl. Phys. B 70, 39 (1974)ADSCrossRefGoogle Scholar
 5.S.R. Coleman, J. Mandula, Phys. Rev. 159, 1251 (1967)ADSCrossRefGoogle Scholar
 6.R. Haag, J.T. Lopuszanski, M. Sohnius, Nucl. Phys. B 88, 257 (1975)ADSCrossRefGoogle Scholar
 7.P. Ramond (2016). https://workshops.ift.uamcsic.es/files/205/Ramond.pdf
 8.E. Gildener, Phys. Rev. D 14, 1667 (1976)ADSCrossRefGoogle Scholar
 9.M.J.G. Veltman, Acta Phys. Polon. B 12, 437 (1981)Google Scholar
 10.
 11.M. Drees, J.S. Kim, Phys. Rev. D 93, 095005 (2016). arXiv:1511.04461 ADSCrossRefGoogle Scholar
 12.S.S. AbdusSalam et al., Eur. Phys. J. C 71, 1835 (2011). arXiv:1109.3859 ADSCrossRefGoogle Scholar
 13.S. Dimopoulos, H. Georgi, Nucl. Phys. B 193, 150 (1981)ADSCrossRefGoogle Scholar
 14.H.P. Nilles, Phys. Lett. B 115, 193 (1982)ADSCrossRefGoogle Scholar
 15.R. Barbieri, S. Ferrara, C.A. Savoy, Phys. Lett. B 119, 343 (1982)ADSCrossRefGoogle Scholar
 16.A.H. Chamseddine, R.L. Arnowitt, P. Nath, Phys. Rev. Lett. 49, 970 (1982)ADSCrossRefGoogle Scholar
 17.G.L. Kane, C.F. Kolda, L. Roszkowski, J.D. Wells, Phys. Rev. D 49, 6173 (1994). arXiv:hepph/9312272 ADSCrossRefGoogle Scholar
 18.S. Cassel, D.M. Ghilencea, S. Kraml, A. Lessa, G.G. Ross, JHEP 05, 120 (2011). arXiv:1101.4664 ADSCrossRefGoogle Scholar
 19.P. Bechtle et al., JHEP 06, 098 (2012). arXiv:1204.4199 ADSCrossRefGoogle Scholar
 20.O. Buchmueller et al., Eur. Phys. J. C 74, 2922 (2014). arXiv:1312.5250 ADSCrossRefGoogle Scholar
 21.P. Bechtle et al., PoS EPS–HEP2013, 313 (2013). arXiv:1310.3045
 22.P. Bechtle et al., Eur. Phys. J. C 76, 96 (2016). arXiv:1508.05951 ADSCrossRefGoogle Scholar
 23.A. Djouadi et al. (MSSM Working Group), in GDR (Groupement De Recherche)—Supersymetrie Montpellier, France, April 15–17, 1998 (1998). arXiv:hepph/9901246. http://inspirehep.net/record/481987/files/arXiv:hepph_9901246.pdf
 24.C.F. Berger, J.S. Gainer, J.L. Hewett, T.G. Rizzo, JHEP 02, 023 (2009). arXiv:0812.0980 ADSCrossRefGoogle Scholar
 25.J.A. Conley, J.S. Gainer, J.L. Hewett, M.P. Le, T.G. Rizzo, Eur. Phys. J. C 71, 1697 (2011). arXiv:1009.2539 ADSCrossRefGoogle Scholar
 26.K.J. de Vries et al., Eur. Phys. J. C 75, 422 (2015). arXiv:1504.03260 ADSCrossRefGoogle Scholar
 27.G. Bertone, F. Calore, S. Caron, R. Ruiz, J.S. Kim, R. Trotta, C. Weniger, JCAP 1604, 037 (2016). arXiv:1507.07008 ADSCrossRefGoogle Scholar
 28.H.K. Dreiner (1997). [Adv. Ser. Direct. High Energy Phys. 21, 565 (2010)]. arXiv:hepph/9707435
 29.R. Barbier et al., Phys. Rep. 420, 1 (2005). arXiv:hepph/0406039 ADSCrossRefGoogle Scholar
 30.N. Escudero, D.E. LopezFogliani, C. Munoz, R. Ruiz de Austri, JHEP 12, 099 (2008). arXiv:0810.1507 ADSCrossRefGoogle Scholar
 31.S. Weinberg, Phys. Rev. D 26, 287 (1982)ADSCrossRefGoogle Scholar
 32.L.J. Hall, M. Suzuki, Nucl. Phys. B 231, 419 (1984)ADSCrossRefGoogle Scholar
 33.H.K. Dreiner, M. Thormeier, Phys. Rev. D 69, 053002 (2004). arXiv:hepph/0305270 ADSCrossRefGoogle Scholar
 34.B.C. Allanach, A. Dedes, H.K. Dreiner, Phys. Rev. D 69, 115002 (2004). [Erratum: Phys. Rev. D 72, 079902 (2005)]. arXiv:hepph/0309196
 35.R.M. Godbole, P. Roy, X. Tata, Nucl. Phys. B 401, 67 (1993). arXiv:hepph/9209251 ADSCrossRefGoogle Scholar
 36.J.L. Goity, M. Sher, Phys. Lett. B 346, 69 (1995). [Erratum: Phys. Lett. B 385, 500 (1996)]. arXiv:hepph/9412208
 37.L.E. Ibanez, G.G. Ross, Phys. Lett. B 260, 291 (1991)ADSCrossRefGoogle Scholar
 38.T. Banks, M. Dine, Phys. Rev. D 45, 1424 (1992). arXiv:hepth/9109045 ADSMathSciNetCrossRefGoogle Scholar
 39.H.K. Dreiner, C. Luhn, M. Thormeier, Phys. Rev. D 73, 075007 (2006). arXiv:hepph/0512163 ADSCrossRefGoogle Scholar
 40.H.K. Dreiner, C. Luhn, H. Murayama, M. Thormeier, Nucl. Phys. B 774, 127 (2007). arXiv:hepph/0610026 ADSCrossRefGoogle Scholar
 41.H.K. Dreiner, M. Hanussek, J.S. Kim, C.H. Kom, Phys. Rev. D 84, 113005 (2011). arXiv:1106.4338 ADSCrossRefGoogle Scholar
 42.H.K. Dreiner, M. Hanussek, C. Luhn, Phys. Rev. D 86, 055012 (2012). arXiv:1206.6305 ADSCrossRefGoogle Scholar
 43.P. Fileviez Perez, Int. J. Mod. Phys. A 28, 1330024 (2013). arXiv:1305.6935
 44.H.K. Dreiner, H. Murayama, M. Thormeier, Nucl. Phys. B 729, 278 (2005). arXiv:hepph/0312012 ADSCrossRefGoogle Scholar
 45.C. Csaki, Y. Grossman, B. Heidenreich, Phys. Rev. D 85, 095009 (2012). arXiv:1111.1239 ADSCrossRefGoogle Scholar
 46.S. Davidson, M. Losada, JHEP 05, 021 (2000). arXiv:hepph/0005080 ADSCrossRefGoogle Scholar
 47.D.E. LopezFogliani, C. Munoz, Phys. Rev. Lett. 97, 041801 (2006). arXiv:hepph/0508297 ADSCrossRefGoogle Scholar
 48.H.B. Kim, J.E. Kim, Phys. Lett. B 527, 18 (2002). arXiv:hepph/0108101 ADSCrossRefGoogle Scholar
 49.E.J. Chun, H.B. Kim, Phys. Rev. D 60, 095006 (1999). arXiv:hepph/9906392 ADSCrossRefGoogle Scholar
 50.E.J. Chun, H.B. Kim, JHEP 10, 082 (2006). arXiv:hepph/0607076 ADSCrossRefGoogle Scholar
 51.H.K. Dreiner, F. Staub, L. Ubaldi, Phys. Rev. D 90, 055016 (2014). arXiv:1402.5977 ADSCrossRefGoogle Scholar
 52.W. Buchmuller, L. Covi, K. Hamaguchi, A. Ibarra, T. Yanagida, JHEP 03, 037 (2007). arXiv:hepph/0702184 ADSCrossRefGoogle Scholar
 53.G. Arcadi, L. Covi, M. Nardecchia, Phys. Rev. D 92, 115006 (2015). arXiv:1507.05584 ADSCrossRefGoogle Scholar
 54.P.A.R. Ade et al. (Planck), Astron. Astrophys. 571, A16 (2014). arXiv:1303.5076
 55.H.K. Dreiner, K. Nickel, F. Staub, Phys. Lett. B 742, 261 (2015). arXiv:1411.3731 ADSCrossRefGoogle Scholar
 56.ATLAS, ATLASCONF2013036 (2013)Google Scholar
 57.Tech. Rep. ATLASCONF2016057, CERN, Geneva (2016). https://cds.cern.ch/record/2206149
 58.R. Caminal Armadans, Nucl. Part. Phys. Proc 273–275, 618 (2016)Google Scholar
 59.B.C. Allanach, A. Dedes, H.K. Dreiner, Phys. Rev. D 60, 056002 (1999). [Erratum: Phys. Rev. D 86, 039906 (2012)]. arXiv:hepph/9902251
 60.H.K. Dreiner, G.G. Ross, Nucl. Phys. B 365, 597 (1991)ADSCrossRefGoogle Scholar
 61.H.K. Dreiner, F. Staub, A. Vicente, W. Porod, Phys. Rev. D 86, 035021 (2012). arXiv:1205.0557 ADSCrossRefGoogle Scholar
 62.H.K. Dreiner, S. Grab, Phys. Lett. B 679, 45 (2009). arXiv:0811.0200 ADSCrossRefGoogle Scholar
 63.M. Drees, H. Dreiner, D. Schmeier, J. Tattersall, J.S. Kim, Comput. Phys. Commun. 187, 227 (2015). arXiv:1312.2591 ADSCrossRefGoogle Scholar
 64.J.S. Kim, D. Schmeier, J. Tattersall, K. Rolbiecki, Comput. Phys. Commun 196, 535 (2015). arXiv:1503.01123 ADSCrossRefGoogle Scholar
 65.D. Dercks, N. Desai, J.S. Kim, K. Rolbiecki, J. Tattersall, T. Weber (2016). arXiv:1611.09856
 66.B. Brahmachari, P. Roy, Phys. Rev. D 50, R39 (1994). [Phys. Rev. D 50(1), R39 (1994)], arXiv:hepph/9403350
 67.R. Hempfling, Nucl. Phys. B 478, 3 (1996). arXiv:hepph/9511288 ADSCrossRefGoogle Scholar
 68.V.D. Barger, M.S. Berger, R.J.N. Phillips, T. Wohrmann, Phys. Rev. D 53, 6407 (1996). arXiv:hepph/9511473 ADSCrossRefGoogle Scholar
 69.B. de Carlos, P.L. White, Phys. Rev. D 54, 3427 (1996). arXiv:hepph/9602381 ADSCrossRefGoogle Scholar
 70.E. Nardi, Phys. Rev. D 55, 5772 (1997). arXiv:hepph/9610540 ADSCrossRefGoogle Scholar
 71.F. Staub (2008). arXiv:0806.0538
 72.F. Staub, Comput. Phys. Commun. 181, 1077 (2010). arXiv:0909.2863 ADSCrossRefGoogle Scholar
 73.F. Staub, Comput. Phys. Commun. 182, 808 (2011). arXiv:1002.0840 ADSCrossRefGoogle Scholar
 74.F. Staub, Comput. Phys. Commun. 184, 1792 (2013). arXiv:1207.0906 ADSCrossRefGoogle Scholar
 75.F. Staub, Comput. Phys. Commun. 185, 1773 (2014). arXiv:1309.7223 ADSCrossRefGoogle Scholar
 76.F. Staub, Adv. High Energy Phys 2015, 840780 (2015). arXiv:1503.04200 MathSciNetGoogle Scholar
 77.W. Porod, Comput. Phys. Commun. 153, 275 (2003). arXiv:hepph/0301101 ADSCrossRefGoogle Scholar
 78.W. Porod, F. Staub, Comput. Phys. Commun. 183, 2458 (2012). arXiv:1104.1573 ADSCrossRefGoogle Scholar
 79.G. Aad et al. (ATLAS, CMS), Phys. Rev. Lett 114, 191803 (2015). arXiv:1503.07589
 80.P. Athron, J.H. Park, T. Steudtner, D. Stöckinger, A. Voigt, JHEP 01, 079 (2017). arXiv:1609.00371
 81.J.M. Frere, D.R.T. Jones, S. Raby, Nucl. Phys. B 222, 11 (1983)ADSCrossRefGoogle Scholar
 82.J.A. Casas, A. Lleyda, C. Munoz, Nucl. Phys. B 471, 3 (1996). arXiv:hepph/9507294 ADSCrossRefGoogle Scholar
 83.J.E. CamargoMolina, B. O’Leary, W. Porod, F. Staub, Eur. Phys. J. C 73, 2588 (2013). arXiv:1307.1477 ADSCrossRefGoogle Scholar
 84.J.E. CamargoMolina, B. O’Leary, W. Porod, F. Staub, JHEP 12, 103 (2013). arXiv:1309.7212 ADSCrossRefGoogle Scholar
 85.M. Davier, A. Hoecker, B. Malaescu, Z. Zhang, Eur. Phys. J. C 71, 1515 (2011). [Erratum: Eur. Phys. J. C 72, 1874 (2012)]. arXiv:1010.4180
 86.D. Stockinger, in SUSY 2007 Proceedings, 15th International Conference on Supersymmetry and Unification of Fundamental Interactions, July 26–August 1, 2007, Karlsruhe, Germany (2007). pp. 720–723. arXiv:0710.2429, http://www.susy07.unikarlsruhe.de/Proceedings/proceedings/susy07.pdf
 87.F. Jegerlehner (2017). arXiv:1705.00263
 88.A. Gould, B.T. Draine, R.W. Romani, S. Nussinov, Phys. Lett. B 238, 337 (1990)ADSCrossRefGoogle Scholar
 89.L.E. Ibanez, C. Lopez, Nucl. Phys. B 233, 511 (1984)ADSCrossRefGoogle Scholar
 90.L.E. Ibanez, C. Lopez, C. Munoz, Nucl. Phys. B 256, 218 (1985)ADSCrossRefGoogle Scholar
 91.M. Drees, S.P. Martin (1995). arXiv:hepph/9504324
 92.S.P. Martin (1997). [Adv. Ser. Direct. High Energy Phys. 18, 1 (1998)]. arXiv:hepph/9709356
 93.H.K. Dreiner, S. Grab, M.K. Trenkel, Phys. Rev. D 79, 016002 (2009). [Erratum: Phys. Rev. 79, 019902 (2009)]. arXiv:0808.3079
 94.B. de Carlos, P.L. White, Phys. Rev. D 55, 4222 (1997). arXiv:hepph/9609443 ADSCrossRefGoogle Scholar
 95.B.C. Allanach, M.A. Bernhardt, H.K. Dreiner, C.H. Kom, P. Richardson, Phys. Rev. D 75, 035002 (2007). arXiv:hepph/0609263 ADSCrossRefGoogle Scholar
 96.K. Desch, S. Fleischmann, P. Wienemann, H.K. Dreiner, S. Grab, Phys. Rev. D 83, 015013 (2011). arXiv:1008.1580 ADSCrossRefGoogle Scholar
 97.D. Choudhury, H.K. Dreiner, P. Richardson, S. Sarkar, Phys. Rev. D 61, 095009 (2000). arXiv:hepph/9911365 ADSCrossRefGoogle Scholar
 98.H.K. Dreiner, C. Hanhart, U. Langenfeld, D.R. Phillips, Phys. Rev. D 68, 055004 (2003). arXiv:hepph/0304289 ADSCrossRefGoogle Scholar
 99.H.K. Dreiner, S. Heinemeyer, O. Kittel, U. Langenfeld, A.M. Weber, G. Weiglein, Eur. Phys. J. C 62, 547 (2009). arXiv:0901.3485 ADSCrossRefGoogle Scholar
 100.H.K. Dreiner, M. Hanussek, J.S. Kim, S. Sarkar, Phys. Rev. D 85, 065027 (2012). arXiv:1111.5715 ADSCrossRefGoogle Scholar
 101.C. Patrignani et al. (Particle Data Group), Chin. Phys. C 40, 100001 (2016)Google Scholar
 102.S. Dawson, Nucl. Phys. B 261, 297 (1985)ADSMathSciNetCrossRefGoogle Scholar
 103.H.K. Dreiner, P. Morawitz, Nucl. Phys. B 428, 31 (1994). [Erratum: Nucl. Phys. B 574, 874 (2000)]. arXiv:hepph/9405253
 104.K. Agashe, M. Graesser, Phys. Rev. D 54, 4445 (1996). arXiv:hepph/9510439 ADSCrossRefGoogle Scholar
 105.B.C. Allanach, A. Dedes, H.K. Dreiner, Phys. Rev. D 60, 075014 (1999). arXiv:hepph/9906209 ADSCrossRefGoogle Scholar
 106.G. Aad et al. (ATLAS), JHEP 12, 124 (2012). arXiv:1210.4457
 107.G. Aad et al. (ATLAS), Phys. Rev. D 90, 052001 (2014). arXiv:1405.5086
 108.G. Aad et al. (ATLAS), JHEP 08, 138 (2015). arXiv:1411.2921
 109.CMS, CMSPASSUS13010 (2013)Google Scholar
 110.V. Khachatryan et al. (CMS), Phys. Rev. D 94, 112009 (2016). arXiv:1606.08076
 111.CMS, CMSPASSUS16022 (2016). https://cds.cern.ch/record/2205165
 112.CMS, CMSPASSUS12027 (2012). http://cds.cern.ch/record/1494689
 113.ATLAS, ATLASCONF2016075 (2016).Google Scholar
 114.S. Chatrchyan et al. (CMS), Phys. Rev. Lett. 111, 221801 (2013). arXiv:1306.6643
 115.CMS, CMSPASSUS13010 (2013). https://cds.cern.ch/record/1550552
 116.Tech. Rep. ATLASCONF2015018, CERN, Geneva (2015). https://cds.cern.ch/record/2017303
 117.Tech. Rep. CMSPASSUS13005, CERN, Geneva (2013). https://cds.cern.ch/record/1547780
 118.M. Aaboud et al. (ATLAS), (2017). arXiv:1706.03731
 119.M. Aaboud et al. (ATLAS), (2017). arXiv:1704.08493
 120.V. Khachatryan et al. (CMS), Phys. Lett. B 760, 178 (2016). arXiv:1602.04334
 121.V. Khachatryan et al. (CMS), Phys. Lett. B 739, 229 (2014). arXiv:1408.0806
 122.V. Khachatryan et al. (CMS), Phys. Rev. D 93, 032004 (2016). arXiv:1509.03744
 123.H.K. Dreiner, P. Richardson, M.H. Seymour, Phys. Rev. D 63, 055008 (2001). arXiv:hepph/0007228 ADSCrossRefGoogle Scholar
 124.H.K. Dreiner, S. Grab, M. Kramer, M.K. Trenkel, Phys. Rev. D 75, 035003 (2007). arXiv:hepph/0611195 ADSCrossRefGoogle Scholar
 125.G. Aad et al. (ATLAS), JHEP 04, 116 (2015). arXiv:1501.03555
 126.Tech. Rep. CMSPASSUS16042, CERN, Geneva (2017). http://cds.cern.ch/record/2257294
 127.Tech. Rep. CMSPASSUS16037, CERN, Geneva (2017). http://cds.cern.ch/record/2256652
 128.G. Aad et al. (ATLAS), JHEP 10, 130 (2013). [Erratum: JHEP 01, 109 (2014)]. arXiv:1308.1841
 129.G. Aad et al. (ATLAS), JHEP 09, 176 (2014). arXiv:1405.7875
 130.M. Aaboud et al. (ATLAS), Eur. Phys. J. C 76, 392 (2016). arXiv:1605.03814
 131.A.M. Sirunyan et al. (CMS), (2017). arXiv:1704.07781
 132.Tech. Rep. CMSPASSUS16036, CERN, Geneva (2017). http://cds.cern.ch/record/2256872
 133.G. Aad et al. (ATLAS), JHEP 06, 035 (2014). arXiv:1404.2500
 134.G. Aad et al. (ATLAS), JHEP 12, 086 (2012). arXiv:1210.4813
 135.G. Aad et al. (ATLAS), Phys. Rev. D 91, 112016 (2015). [Erratum: Phys. Rev. D 93(3), 039901 (2016)]. arXiv:1502.05686
 136.G. Aad et al. (ATLAS), JHEP 06, 067 (2016). arXiv:1601.07453
 137.ATLAS, ATLASCONF2016022 (2016)Google Scholar
 138.V. Khachatryan et al. (CMS), Phys. Lett. (2016). arXiv:1608.01224
 139.V. Khachatryan et al. (CMS), Phys. Lett. B 747, 98 (2015). arXiv:1412.7706
 140.S. Chatrchyan et al. (CMS), Phys. Lett. B 730, 193 (2014). arXiv:1311.1799
 141.S. Chatrchyan et al. (CMS), JHEP 01, 163 (2014). [Erratum: JHEP 01, 014 (2015)]. arXiv:1311.6736
 142.ATLAS, ATLASCONF2013091 (2013)Google Scholar
 143.CMS (CMS Collaboration), Tech. Rep. CMSPASSUS16013, CERN, Geneva (2016). https://cds.cern.ch/record/2205147
 144.ATLAS, ATLASCONF2016094 (2016)Google Scholar
 145.ATLAS, ATLASCONF2017013 (2017)Google Scholar
 146.F. de Campos, O.J.P. Eboli, M.B. Magro, W. Porod, D. Restrepo, M. Hirsch, J.W.F. Valle, JHEP 05, 048 (2008). arXiv:0712.2156 CrossRefGoogle Scholar
 147.CMS, CMSPASSUS16039 (2017)Google Scholar
 148.Tech. Rep. ATLASCONF2015026, CERN, Geneva (2015). http://cds.cern.ch/record/2037653
 149.G. Aad et al. (ATLAS), Phys. Rev. Lett 115, 031801 (2015). arXiv:1503.04430
 150.V. Khachatryan et al. (CMS), Eur. Phys. J. C 76, 317 (2016). arXiv:1604.05239
 151.A. Faessler, T.S. Kosmas, S. Kovalenko, J.D. Vergados (1999). arXiv:hepph/9904335
 152.W.H. Bertl et al. (SINDRUM II), Eur. Phys. J. C 47, 337 (2006)Google Scholar
 153.Tech. Rep. CMSPASEXO16029, CERN, Geneva (2016). https://cds.cern.ch/record/2231062
 154.Tech. Rep. ATLASCONF2016084, CERN, Geneva (2016). https://cds.cern.ch/record/2206277
 155.J.A. Evans, Y. Kats, JHEP 04, 028 (2013). arXiv:1209.0764 ADSCrossRefGoogle Scholar
 156.Y. Bai, A. Katz, B. Tweedie, JHEP 01, 040 (2014). arXiv:1309.6631 ADSCrossRefGoogle Scholar
 157.Tech. Rep. CMSPASEXO16001, CERN, Geneva (2016). https://cds.cern.ch/record/2149046
 158.M. Aaboud et al. (ATLAS), Eur. Phys. J. C 76, 541 (2016). arXiv:1607.08079
 159.G. Aad et al. (ATLAS), Eur. Phys. J. C 76, 5 (2016). arXiv:1508.04735
 160.Tech. Rep. ATLASCONF2015015, CERN, Geneva (2015). https://cds.cern.ch/record/2002885
 161.V. Khachatryan et al. (CMS), JHEP 07, 042 (2015). [Erratum: JHEP 11, 056 (2016)]. arXiv:1503.09049
 162.S. Dimopoulos, R. Esmailzadeh, L.J. Hall, G.D. Starkman, Phys. Rev. D 41, 2099 (1990)ADSCrossRefGoogle Scholar
 163.J.L. Feng, J.F. Gunion, T. Han, Phys. Rev. D 58, 071701 (1998). arXiv:hepph/9711414 ADSCrossRefGoogle Scholar
 164.J.L. Hewett, T.G. Rizzo, in Highenergy physics. Proceedings, ICHEP’98, Vancouver, Canada, July 23–29, 1998, vol. 1, 2, pp. 1698–1702 (1998). arXiv:hepph/9809525. http://wwwpublic.slac.stanford.edu/sciDoc/docMeta.aspx?slacPubNumber=SLACPUB7926
 165.H.K. Dreiner, T. Stefaniak, Phys. Rev. D 86, 055010 (2012). arXiv:1201.5014 ADSCrossRefGoogle Scholar
 166.G. Moreau, E. Perez, G. Polesello, Nucl. Phys. B 604, 3 (2001). arXiv:hepph/0003012 ADSCrossRefGoogle Scholar
 167.M.A. Bernhardt, S.P. Das, H.K. Dreiner, S. Grab, Phys. Rev. D 79, 035003 (2009). arXiv:0810.3423 ADSCrossRefGoogle Scholar
 168.S. Chatrchyan et al. (CMS), JHEP 12, 055 (2012). arXiv:1210.5627
 169.J.A. Evans, Y. Kats, PoS EPS–HEP2013, 287 (2013). arXiv:1311.0890
 170.E.J. Chun, S. Jung, H.M. Lee, S.C. Park, Phys. Rev. D 90, 115023 (2014). arXiv:1408.4508 ADSCrossRefGoogle Scholar
 171.Z. Marshall, B.A. Ovrut, A. Purves, S. Spinner, Phys. Lett. B 732, 325 (2014). arXiv:1401.7989 ADSCrossRefGoogle Scholar
 172.G. Aad et al. (ATLAS), JHEP 10, 054 (2015). arXiv:1507.05525
 173.S. Chatrchyan et al. (CMS), Phys. Rev. D 90, 112001 (2014). arXiv:1405.3961
 174.J. Alwall, R. Frederix, S. Frixione, V. Hirschi, F. Maltoni, O. Mattelaer, H.S. Shao, T. Stelzer, P. Torrielli, M. Zaro, JHEP 07, 079 (2014). arXiv:1405.0301 ADSCrossRefGoogle Scholar
 175.T. Sjöstrand, S. Ask, J.R. Christiansen, R. Corke, N. Desai, P. Ilten, S. Mrenna, S. Prestel, C.O. Rasmussen, P.Z. Skands, Comput. Phys. Commun. 191, 159 (2015). arXiv:1410.3012 ADSCrossRefGoogle Scholar
 176.J. de Favereau, C. Delaere, P. Demin, A. Giammanco, V. Lemaître, A. Mertens, M. Selvaggi (DELPHES 3), JHEP 02, 057 (2014). arXiv:1307.6346
 177.M. Cacciari, G.P. Salam, G. Soyez, Eur. Phys. J. C 72, 1896 (2012). arXiv:1111.6097 ADSCrossRefGoogle Scholar
 178.M. Cacciari, G.P. Salam, Phys. Lett. B 641, 57 (2006). arXiv:hepph/0512210 ADSCrossRefGoogle Scholar
 179.A.L. Read, J. Phys. G 28, 2693 (2002)ADSCrossRefGoogle Scholar
 180.C. Degrande, C. Duhr, B. Fuks, D. Grellscheid, O. Mattelaer, T. Reiter, Comput. Phys. Commun. 183, 1201 (2012). arXiv:1108.2040 ADSCrossRefGoogle Scholar
 181.M.D. Goodsell, K. Nickel, F. Staub, Eur. Phys. J. C 75, 32 (2015). arXiv:1411.0675 ADSCrossRefGoogle Scholar
 182.M. Goodsell, K. Nickel, F. Staub, Eur. Phys. J. C 75, 290 (2015). arXiv:1503.03098 ADSCrossRefGoogle Scholar
 183.M.D. Goodsell, F. Staub, Eur. Phys. J. C 77, 46 (2017). arXiv:1604.05335 ADSCrossRefGoogle Scholar
 184.H.K. Dreiner, M. Kramer, J. Tattersall, Europhys. Lett. 99, 61001 (2012). arXiv:1207.1613 ADSCrossRefGoogle Scholar
 185.W. Beenakker, R. Hopker, M. Spira, P. Zerwas, Nucl. Phys. B 492, 51 (1997). arXiv:hepph/9610490 ADSCrossRefGoogle Scholar
 186.W. Beenakker, M. Kramer, T. Plehn, M. Spira, P.M. Zerwas, Nucl. Phys. B 515, 3 (1998). arXiv:hepph/9710451 ADSCrossRefGoogle Scholar
 187.A. Kulesza, L. Motyka, Phys. Rev. Lett. 102, 111802 (2009). arXiv:0807.2405 ADSCrossRefGoogle Scholar
 188.A. Kulesza, L. Motyka, Phys. Rev. D 80, 095004 (2009). arXiv:0905.4749 ADSCrossRefGoogle Scholar
 189.W. Beenakker, S. Brensing, M. Kramer, A. Kulesza, E. Laenen et al., JHEP 0912, 041 (2009). arXiv:0909.4418 ADSCrossRefGoogle Scholar
 190.W. Beenakker, S. Brensing, M. Kramer, A. Kulesza, E. Laenen, I. Niessen, JHEP 08, 098 (2010). arXiv:1006.4771 ADSCrossRefGoogle Scholar
 191.W. Beenakker, S. Brensing, M. Kramer, A. Kulesza, E. Laenen et al., Int. J. Mod. Phys. A 26, 2637 (2011). arXiv:1105.1110 ADSCrossRefGoogle Scholar
 192.W. Beenakker, R. Hopker, M. Spira (1996). arXiv:hepph/9611232
 193.S. Belkner, Master’s Thesis (2017)Google Scholar
 194.S.P. Martin, J.D. Wells, Phys. Rev. D 64, 035003 (2001). arXiv:hepph/0103067 ADSCrossRefGoogle Scholar
 195.K. Kohri, T. Takahashi, Phys. Lett. B 682, 337 (2010). arXiv:0909.4610 ADSCrossRefGoogle Scholar
 196.Tech. Rep. ATLASCONF2013061, CERN, Geneva (2013). http://cds.cern.ch/record/1557778
 197.H.K. Dreiner, S. Grab, T. Stefaniak, Phys. Rev. D 84, 035023 (2011). arXiv:1102.3189 ADSCrossRefGoogle Scholar
 198.M. Hanussek, J.S. Kim, Phys. Rev. D 85, 115021 (2012). arXiv:1205.0019 ADSCrossRefGoogle Scholar
 199.Tech. Rep. ATLASCONF2013062, CERN, Geneva (2013). http://cds.cern.ch/record/1557779
 200.B.C. Allanach, A.J. Barr, L. Drage, C.G. Lester, D. Morgan, M.A. Parker, B.R. Webber, P. Richardson, JHEP 03, 048 (2001). arXiv:hepph/0102173 ADSCrossRefGoogle Scholar
 201.B.C. Allanach, A.J. Barr, M.A. Parker, P. Richardson, B.R. Webber, JHEP 09, 021 (2001). arXiv:hepph/0106304 ADSCrossRefGoogle Scholar
 202.G. Aad et al. (ATLAS), JHEP 10, 024 (2014). arXiv:1407.0600
 203.M.R. Buckley, D. Feld, S. Macaluso, A. Monteux, D. Shih (2016). arXiv:1610.08059
 204.D. Bardhan, A. Chakraborty, D. Choudhury, D.K. Ghosh, M. Maity (2016). arXiv:1611.03846
 205.C. Anders, C. Bernaciak, G. Kasieczka, T. Plehn, T. Schell, Phys. Rev. D 89, 074047 (2014). arXiv:1312.1504 ADSCrossRefGoogle Scholar
 206.A. Abdesselam et al., Eur. Phys. J. C 71, 1661 (2011). arXiv:1012.5412 ADSCrossRefGoogle Scholar
Copyright information
Open AccessThis article is distributed under the terms of the Creative Commons Attribution 4.0 International License (http://creativecommons.org/licenses/by/4.0/), which permits unrestricted use, distribution, and reproduction in any medium, provided you give appropriate credit to the original author(s) and the source, provide a link to the Creative Commons license, and indicate if changes were made.
Funded by SCOAP^{3}