Top Quark as a Dark Portal and Neutrino Mass Generation
aa r X i v : . [ h e p - ph ] J u l Top Quark as a Dark Portal and Neutrino Mass Generation
John N. Ng a , Alejandro de la Puente a a Theory Group, TRIUMF, Vancouver BC V6T 2A3, Canada
Abstract
We present a new model for radiatively generating Majorana active neutrino masses while incorporating a viable dark mattercandidate. This is possible by extending the Standard Model with a single Majorana neutrino endowed with a dark parity, acolour electroweak singlet scalar, as well as a colour electroweak triplet scalar. Within this framework, the up -type quarks playa special role, serving as a portal for dark matter, and a messenger for neutrino mass generation. We consider three benchmarkscenarios where the abundance of dark matter can match the latest experimental results, while generating neutrino masses in themilli-electronvolt range. We show how constraints from lepton flavour violation, in particular the branching fraction of µ → e γ , canplace lower bounds on the coupling between our dark matter candidate and top quarks. Furthermore, we show that this couplingcan also be constrained using collider data from the Tevatron and the LHC.
1. Introduction
We now have compelling evidence for the existence of threeactive neutrino species [1]. Radiochemical experiments such asHomestake, Gallex / GNO and SAGE [2, 3, 4] together with theSuperK and SNO experiments [6, 7] have narrowed down themass patterns to three possibilities: A normal or an inverted hi-erarchy or almost degenerate masses. Moreover, the absolutescale of neutrino masses remains unknown. Additionally, thelast mixing angle, θ , has been measured by several reactor ex-periments [8, 9, 10] and the T2K accelerator experiment [11].This is a breakthrough for the standard picture of neutrino oscil-lations since it now paves the way towards measuring CP viola-tion in the lepton sector. Furthermore, the evidence for neutrinomasses represents one clear motivation for new physics beyondthe Standard Model (SM). Within the SM, neutrinos are mass-less; they can be accommodated in a variety of ways such asincorporating new degrees of freedom and / or new e ff ective in-teractions. Extending the SM model in this way allows us tobe sensitive to new high energy scales. Take for example theType I seesaw mechanism [12], where the SM is extended witha singlet Majorana fermion that couples to left-handed leptonsthrough the Higgs, as with the charged leptons. This class ofmodels generates viable neutrino masses with a Majorana massscale & GeV and Yukawa interaction of order one. Such ahigh seesaw scale can arise from Grand Unified models such asSO(10) [13]. However, such a high scale for new physics makesthe mechanism impossible to test. A TeV scale Majorana massis also possible in models such as left-right symmetric mod-els, (for a recent discussion see [14]). Models with flavoursymmetries are also used to explain the neutrino masses (see[15, 16] for recent reviews). Models where neutrino masses areradiatively generated have also been studied. In particular, thesimplest model where neutrino masses are induced as one-loopradiative corrections was first introduced in [17]. In this classof models a charged scalar singlet under the SM gauge group couples to left-handed lepton doublets and one is able to gen-erate active neutrino masses of the right order with a chargedscalar mass scale as low as a TeV. Neutrino masses may alsoarise as two loop radiative corrections in extensions of the SMwith an additional singlet charged scalar and a doubly chargedscalar [18, 19, 20]. The main motivation for this class of mod-els is that they employ new physics at the TeV scale and hencecan be probed at the LHC.The nature of the neutrino mass matrix can be accessedthrough data on neutrino oscillations. In the gauge basis themass matrix can be parametrized in the following way: m αβ = X i m i U α i U ∗ β i , (1)where α, β = e , µ, τ , i = , , U α, i are the neutrino mixingmatrix elements. In general, if neutrinos are Majorana fermionsthen two new independent degrees of freedom, the Majoranaphases, exist and are usually assigned to the m i ’s. The experi-mental status on the neutrino oscillation parameters is summa-rized in [21] ∆ m = . + . − . × − eV sin θ = . + . − . | ∆ m | = ( . ± . × − eV Normal Hierarchy2 . + . − . × − eV Inverted Hierarchysin θ = . ± . . (2)Another strong indicator of physics beyond the SM is theample evidence pointing towards the existence of dark mat-ter [22, 23]. Velocity dispersion and rotation curves of galaxiessuggest the existence of non-luminous matter that is not com-posed by any of the known SM particles. Furthermore, the mostrecent data from Plank estimates a cold dark matter cosmolog-ical parameter Ω DM h = . ± . . Preprint submitted to Physis Letters B July 16, 2018 al evidence for dark matter is due to its gravitational proper-ties and its identity remains unknown to date. One candidateexplanation for dark matter is the existence of a weakly inter-acting massive particle (WIMP). Supersymmetric models areknown to provide a natural WIMP candidate, usually the light-est superpartner. The abundance of these particles in the uni-verse is determined by their self-annihilation rate in relation tothe expansion of the universe. When the expansion rate domi-nates over the rate of annihilation, interaction among dark mat-ter particles becomes less e ffi cient and their density becomesa constant or “freezes out”. There is, however, some possiblesignal regions for WIMP scattering with nuclei in direct detec-tion experiments, most notably the DAMA / LIBRA result [24]and CRESST [25]. Experimental upper limits on the WIMP-nucleon cross section have also been found by various experi-ments [26, 27, 28].In this work we present a new model for radiatively generat-ing Majorana active neutrino masses while incorporating a vi-able dark matter candidate. This is possible by extending theSM with a single electroweak singlet Majorana neutrino, N R , towhich we also assign an odd parity, referred to as dark parity(DP). We also add a colour electroweak singlet scalar, ψ , and acolour electroweak triplet scalar, χ . These are, respectively, oddand even under DP. Furthermore, all SM fields have even DP.Of the two particles that are odd under DP, we assume N R to bethe lightest. Such an assignment makes N R a good dark mattercandidate since it will be stable, as we engineer DP to be unbro-ken. It also forbids the usual coupling of N R to the SM leptondoublet and the Higgs doublet and hence no Dirac mass termis generated. The new colour scalars couple to quark fields, inparticular the up -type quarks. In our framework, the up -typequarks play two roles: the first one, serving as a messengerfor neutrino mass generation. This is possible given the richstructure of the Lagrangian which is used to radiatively gen-erate masses for the left-handed neutrinos via the exchange ofthe exotic colour scalars at three loops. The second role is as aportal for dark matter, where the relic abundance of dark matteris reproduced through renormalizable interactions between theMajorana neutrino and up -type quarks via ψ . We study threebenchmark scenarios where the abundance of dark matter canmatch the latest experimental results, while generating neutrinomasses in the milli-electronvolt range. Our model is consistentwith constraints from lepton flavour violation and collider datafrom the Tevatron and the LHC. The idea of using a discretesymmetry such as a Z parity to forbid a Dirac mass term forthe neutrinos and identify N R as a dark matter candidate wasfirst proposed in [29]. Radiative neutrino masses are generatedby the use of Higgs triplets or inert doublets. Here we explore anew avenue by making use of colour scalars which allow neu-trino masses to be generated at the 3-loop level. Furthermore,the phenomenology at the LHC is richer by virtue that it is verye ffi cient in producing new colour degrees of freedom.
2. Model
The model we consider in this study is an extension to the SMthat incorporates a dark matter candidate and generates Majo- rana masses for the active left-handed neutrinos, radiatively andat the three loop level. Within this framework, N R couples toright handed up -type quarks through a colour electroweak sin-glet scalar, ψ . Furthermore, we incorporate a coupling betweenthe electroweak lepton doublets and the up -type quark doubletsthrough a colour electroweak triplet, χ . The new physics can beparametrized in the following way: L BS M = X i y i ψ u i P L N c ψ + X ℓ, i n λ i ℓ h u i P R (cid:16) χ ν c ℓ + χ ℓ c (cid:17) + d i P R (cid:16) χ ℓ c − χ ν c ℓ (cid:17)io + hc , (3)where l = e , µ, τ and i = , , y i ψ denotes the strength of the interaction between N R and u iR via ψ , while λ il the strength between the quark doublets( u i , d i ) L and ( ν, l ) L via χ . Throughout this work, we make useof P R / L = ± γ . Furthermore, unless otherwise stated, we workin the charged fermion mass basis.Under the SM gauge group S U (3) c × S U (2) W × U (1) Y , ψ transforms as a ( , , / ) and we write the field χ as χ = χ / √ χ χ − χ / √ ! (4)which transforms as a ( , , − / ). These assignments yieldelectric charges of Q = / , − / , − / χ , χ and χ re-spectively.The gauge covariant derivatives for the scalars are given by L kin = ( D µ ψ ) † ( D µ ψ ) + Tr( D µ χ ) † ( D µ χ ) , (5)where D µ ψ = ∂ µ − ig s G a µ λ a − ig ′ ( 23 ) B µ ! ψ D µ χ = ∂ µ − ig s G a µ λ a χ − ig W i µ σ i , χ ] − ig ′ ( −
13 ) B µ χ. (6)The implicit sums are over the generators λ a of SU(3), a = , ...
8, and the generators σ i of SU(2), i = , , N R Majorana mass term, M N R ¯ N cR N R , can be added. This term is even under DP, andwe treat M N R as a free parameter. We further assume that M N R < m ψ which makes N R a suitable dark matter candidate .The gauge and Z invariant potential is given by V ( H , ψ, χ ) = − µ H † H + λ
4! ( H † H ) + m χ Tr( χ † χ ) + λ χ (Tr( χ † χ )) + m ψ ψ † ψ + λ ψ ( ψ † ψ ) + κ H † H Tr( χ † χ ) + κ H † χ † χ H + κ H † H ψ † ψ + ρ Tr( χ † χ ) ψ † ψ (7)where H is the SM Higgs field. In order not to have a colourbreaking vacuum we take m χ , m ψ to be positive. Since all the We have assumed that m ψ > M N R , such that M N R is the lightest stableparticle under the DP. We may also have M N R > m ψ . In the latter case, ourmodel will be one with a strongly interacting dark matter candidate, an analysisthat is beyond the scope of this paper.
3. Dark Matter
As mentioned in the previous section, the unbroken Z sym-metry stabilizes N R . Due to the interaction introduced in Equa-tion (1), the mechanism that leads to a reduction in the relicabundance of N R is via t -channel annihilation into right-handedtop and charm quarks through the exchange of the colour elec-troweak singlet scalar, ψ . In this work we consider three bench-mark points which depict three important regions of parameterspace: M N R = , ,
450 GeV.The evolution of the comoving particle density is given bythe Boltzmann equation˙ nn eq = Γ · n n eq − − H nn eq (8)where n is the particle density at time t and n eq is the density atequilibrium, H is the Hubble expansion rate and Γ parametrizesthe interaction rate, Γ = h σ v i n eq , where h σ v i denotes the ther-mally average annihilation cross section. By solving numeri-cally the above equation one can find the temperature at whichparticles depart from equilibrium and freeze out. This tempera-ture is given by x FO ≡ mT FO ≈ log . g mM Pl h σ v i g / ∗ x / FO , (9)where g denotes the number of degrees of freedom of the par-ticle under consideration and g ∗ the number of relativistic de-grees of freedom at the freeze out temperature. The present dayrelic abundance is then given by Ω DM h ≈ . × GeV − Jg / ∗ M Pl , (10)where J ≡ Z ∞ x FO h σ v i x dx . (11)The thermalized cross section at temperature T can be calcu-lated from the annihilation cross section of our dark matter can-didate, N R . The thermalized cross section is given by (cid:10) σ N R N R v (cid:11) = Z ∞ M NR ds ( s − M N R ) s / K (cid:16) s / / T (cid:17) M N R T K ( M N R / T ) σ ( s ) , (12)where σ ( s ) is the annihilation cross section as a function ofthe center of mass energy squared of the interaction, and K ( z ), K ( z ) are Modified Bessel function of the first and second kindrespectively. We calculated the relic abundance using the latest version of MicOMEGAs [30] and the model files were gener-ated with the latest version of FeynRules [31]. We carried outa scan over three parameters, y t , c ψ , and m ψ , for the three bench-mark points. The results are shown in Figures 1 and 2.
200 300 400 500 600 700 800 900 1000 m ψ [GeV] y t ψ Figure 1: The normalized relic abundance in the y t ψ − m ψ plane. The greyregion corresponds to the region of parameter space consistent with a Majorananeutrino with mass M N R =
150 GeV contributing 75 − M N R =
450 GeV.
200 300 400 500 600 700 800 900 1000 m ψ [GeV] y c ψ Figure 2: The normalized relic abundance in the y c ψ − m ψ plane. The blackregion corresponds to the region of parameter space consistent with a Majorananeutrino with mass M N R =
80 GeV contributing 75 − M N R = , The dependence of the relic abundance on y t ψ and m ψ isshown in Figure 1. The grey region denotes the parameter spaceconsistent with a relic with mass M N R =
150 GeV contributing75 − M N R =
450 GeV. In Figure 2 we3how the dependence of the relic abundance as a function of y c ψ and m ψ . The grey and maroon regions correspond to a relic withmass of 150 and 450 GeV respectively. The scattered behaviourof the grey and maroon regions in Figures 1 and 2 is due to thefact that a combination of annihilation channels are open: c ¯ c and t ¯ c / c ¯ t for a 150 GeV Majorana neutrino and c ¯ c , t ¯ t and t ¯ c / c ¯ t for a 450 GeV Majorana neutrino. This is not the case for arelic with M N R =
80 GeV, where the c ¯ c annihilation channelis the only one open. Here one finds that the relic abundancedepends only on y c ψ and m ψ . For an 80 GeV Majorana neutrino,the region consistent with 75 −
4. Radiative Neutrino Mass generation
The conserved DP allows us to identify N R as a candidate fordark matter and it also forbids Dirac neutrino mass terms for theactive neutrinos, ν i . Therefore, the usual seesaw mechanism isnot operative in this model. However, the Lagrangian of Equa-tion (3) has enough structure to radiatively generate masses for ν i via the exchange of the exotic colour scalars. In particular, ithas the novel feature of using the t R and c R quarks as a portalto communicate with the dark sector and as messengers for theneutrinos. Within this framework, the lowest order diagram forneutrino mass generation is at 3-loops. The diagram is due toexchanges of both ψ and χ fields. This is depicted in Figure 3which gives the ℓ, ℓ ′ element of the active neutrino mass matrix M ν . ν ℓ L ν ℓ ′ L t N t χ χψ ψ Figure 3: 3-loop generation of a Majorana mass for active neutrinos from thet-quark. The crosses on the fermion lines indicate mass insertions. Similardiagrams from the c-quark will also play a role although it gives smaller con-tribution.
This mechanism yields finite contributions to all the elementsof M ν and it is best seen using the mass insertion technique. The ℓℓ ′ element of the active neutrino mass matrix is given by( M ν ) ℓℓ ′ = X i , j K i j λ i ℓ λ j ℓ ′ (13)where i , j = u , c , t and K i j which controls the scale of neutrino masses is given by K ij = y i ψ y j ψ ρ (16 π ) m i m j M N R ( m χ − m i )( m χ − m j ) I ( m χ , m ψ ) , I = Z ∞ du uu + M N R · Z dx ln m χ (1 − x ) + m ψ x + ux (1 − x ) m i (1 − x ) + m ψ x + ux (1 − x ) . (14) From the above equation we see that the u -quark yields a neg-ligible contribution to the neutrino masses and we can concen-trate on the t and c quarks. Furthermore, if only one type ofquark is involved in the neutrino mass generation, then Equa-tion (13) gives rise to two massless active neutrinos, excludedby experimental data. Therefore, at least two quark familiesmust come into play. To simplify the model we assume thatthe top quark gives the main contribution and also demand that λ ce ,µ << λ c τ , such that the c -quark contribution only modifies the3 , M ν . These requirements are su ffi cient to lift thedegeneracy of two massless neutrinos. y t ψ K t [ G e V ] Figure 4: K t , t factor as a function of the N R − t R coupling, y t ψ . The region inblack corresponds to a Majorana neutrino with M N R =
80 GeV while the greyand maroon regions correspond to Majorana neutrino masses of 150 and 450GeV respectively.
Using this framework for neutrino mass generation we ana-lyzed the parameter space consistent with 75-100% of the darkmatter relic abundance, and calculated the K i j factors. In Fig-ure 4 we show the K t , t factor as a function of y t ψ . The blackregion corresponds to a Majorana neutrino with M N R =
80 GeVand the grey and maroon regions correspond to Majorana neu-trino masses of 150 and 450 GeV respectively. We use a colourelectroweak triplet with mass m χ = χ and ψ of ρ = .
1. The bulk of theneutrino mass is due to K t , t since K t , t ≫ K c , c . The K t , t pa-rameter ranges from ∼ one meV to 100 eV for parameter pointsresponsible for 75-100% of the dark matter relic abundance.This range of K t , t values can naturally provide this model with4 milli-electronvolt active neutrino mass. It is easy to see whythe neutrino masses are naturally small. Let us consider the t-quark contribution. The 3-loop suppression yields a factor of10 − . Since the LHC has not seen any new colour states we canassume that m χ > m t m χ ) ∼ − . For M N R =
100 GeV, the factor ( M NR m χ ) gives an-other 10 − suppression. Therefore, sub-eV active neutrinos arenatural in this model and no fine tuning of y t , c ψ or ρ is required. µ → e γ From Equation 3 one can see that the colour electroweaktriplet scalar states will give rise to lepton flavour violating de-cays. In particular, the decay µ → e γ can be used to place alower bound on the y t ψ coupling. In our framework, the branch-ing fraction of µ → e γ is given by Br ( µ → e γ ) = . TeV m χ ! × − | λ t µ λ te + λ c µ λ ce | . (15)Given that K t , t ≫ K c , c , we see that we have no sensitivity to λ c µ λ ce in the definition of M ν . In this work we have maximizedthe contribution to the branching fraction in the limit where λ c µ λ ce ∼ λ t µ λ te . We then extract the value of λ t µ λ te using the re-sults from Figure 4 together with the latest values of m e µ [32]and the current experimental upper bound on Br ( µ → e γ ) ≤ . × − [33]. In the analysis, we have used the best fit rangefor m e µ assuming a normal hierarchy of active neutrino masses, | m e µ | = . − . m χ = Br ( µ → e γ ) = . × − (cid:18) m e µ K t , t (cid:19) (16)Our results are shown in Figure 5, where we plot the normalizedbranching fraction, ξ ( µ → e γ ) = Br ( µ → e γ ) / Br ( µ → e γ ) exp ,as a function of y t ψ using the lower limit on m e µ ; and in Figure 6using the upper limit on m e µ . The black region corresponds to M N R =
80 GeV while the grey and maroon regions to M N R = ,
450 GeV respectively. We see that the lower bound on y t ψ increases with decreasing M N R . This behaviour is due to thefact that the branching fraction is proportional to M N R whileit is inversely proportional to ( y t ψ ) . In particular, we find anupper bound of y t ψ . . − . M N R =
80 GeV and y t ψ . . − . , . − . M N R = ,
450 GeV.An important fact to note is that the constraints placed on y t ψ using the current experimental bound on Br ( µ → e γ ) are not atall sensitive to the mass of the colour electroweak singlet scalar.This scalar plays an important role in mediating the annihilationof the Majorana neutrinos. As we will see below, bounds on themass of this scalar as well as upper bounds on the y t ψ can beobtained using collider data. y t ψ ξ ( µ −> e γ ) Figure 5: Lower limit on the branching fraction normalized to the experi-mental upper bound as a function of y t ψ using the best fit values for m e µ us-ing an electroweak triplet scalar mass, m ψ = M N R =
80 GeV while the grey and maroon regions correspond to M N R = ,
450 GeV respectively.
6. Collider constraints
This model is also highly constrained by data from high en-ergy colliders such as the Tevatron and the LHC. In particular,our model yields two very distinct signatures for which verystringent bounds exist. We used Madgraph 5 [34] to calculatethe parton-level signal prediction and implemented the initialand final state radiation using Pythia [35]. Our signal accep-tances were are calculated with the PGS detector simulationimplementing the cuts in the corresponding LHC and Tevatronanalyses.One constraint is due to dijet plus missing energy (MET)searches at the Tevatron. The latest bounds on this processwere carried out by the CDF collaboration using p ¯ p collisionsat a center of mass energy of √ s = .
96 TeV and 2 . fb − ofintegrated luminosity [36]. Within our framework, two chan-nels can lead to this final state. The first one is t ¯ t productionfollowed by a three body decay of the top quark into two Ma-jorana neutrinos and a charm quark, t → N R N R c . This channelis open as long as N R has a mass below ∼
86 GeV. The secondchannel is through pair production of two colour electroweaksinglets, followed by the decay ψ → N R c . These two channelsare sensitive to y t , c ψ and m ψ . In order to generate exclusions onall three parameters of our model we implemented the experi-mental sample with tight kinematic thresholds of MET > H T >
225 GeV, where H T denotes the scalar sum ofthe two jet transverse energies: H T = E T (jet ) + E T (jet ) (17)The second constraint is due to top squark pair productionin pp collisions with a center of mass energy of √ s = . − of integrated luminosity. We used the results ob-tained with the Compact Muon Solenoid (CMS) detector at the5 y t ψ ξ ( µ −> e γ ) Figure 6: Upper limit on the branching fraction normalized to the experi-mental upper bound as a function of y t ψ using the best fit values for m e µ us-ing an electroweak triplet scalar mass, m ψ = M N R =
80 GeV while the grey and maroon regions correspond to M N R = ,
450 GeV respectively.
LHC. This search looks for decays of a stop squark into a topquark and a neutralino [37]. Top squarks are the scalar partnersof the top quark in supersymmetric extensions of the SM suchas the Minimal Supersymmetric Standard Model (MSSM), andthe neutralino is a linear combination of the fermionic partnersof the neutral gauge bosons and the two neutral Higgs bosons.Within our framework, the colour electroweak singlet, ψ , hasthe same gauge quantum numbers as the top squark but addi-tional decay modes, in particular ψ → N R c . We apply the CMSconstraint using their cut based analysis for three di ff erent METcuts: > , ,
300 GeV.In Figure 7, we show the parameter regions excluded for an80 GeV Majorana neutrino from the four experimental observ-ables mentioned at the beginning of this section. On the top, weplot the excluded region in the y t ψ − m ψ plane for y c ψ = .
25. Theregion labeled 1 corresponds to regions of parameter space ex-cluded by the CMS observable with MET >
200 GeV, while theregions labeled 2 and 3 correspond to MET >
150 and > y c ψ = .
5. For this value of y c ψ the excluded region is smallersince the branching fraction of ψ → N R t is reduced, and thus,the CMS analysis is less sensitive to our model. From the plotsin Figure 7 we also see that no region is excluded by the CDFexperiment for m ψ >
300 GeV. This is not true for masses be-low 300 GeV, where the CDF experiment rules out the entiremodel for M N R =
80 GeV.In Figure 8, we show the regions of parameter space excludedfor a 150 GeV Majorana neutrino. The plot on the top corre-sponds to y c ψ = .
25 while the plot on the bottom to y c ψ = . m ψ lies above 200 GeV but it is not able to ex-clude any of that region of parameter space. Therefore, the onlyrelevant observable is the CMS analysis, which is able to ex-clude a region of parameter space where 320 . m ψ .
550 for y t ψ & . y c ψ = .
25. Again, the excluded region is sig- nificantly smaller for larger values of y c ψ , since the branchingfraction of ψ → N R t is suppressed.The above collider constraints were also applied to a Majo-rana neutrino with M N R =
450 GeV. We found that these con-straints were not strong enough to rule out any of the parameterspace consistent with 75 − + MET was enoughto exclude it as a viable dark matter candidate.
7. Discussion
In this study, we have investigated the possibility of extend-ing the Standard Model with an electroweak singlet Majorananeutrino, stabilized by a new Z symmetry, to explain the abun-dance of the dark matter in the universe. In this model, wecoupled the dark matter candidate to up -type quarks via a newcolour electroweak singlet scalar. Throughout the study weconsidered three benchmark scenarios: M N R = , , − y t , c ψ , and scalar masses, m ψ . This however was not thecase for M N R =
80 GeV, where the only available annihilationchannel was into charm quarks. In this case we found a veryclear dependence of the coupling y c ψ on m ψ .We have also investigated the possibility of radiatively gener-ating Majorana masses for the active neutrinos of the StandardModel by incorporating a colour electroweak triplet scalar inaddition to the colour electroweak singlet scalar. This setup al-lowed us to generate active neutrino masses at three loops. Wefound that the neutrino mass was mostly sensitive to the y t ψ cou-pling, and that for points consistent with 75 − µ → e γ . Wefound that the current experimental bound on the branchingfraction placed lower bounds on the coupling y t ψ independent onthe colour electroweak singlet mass, m ψ . This lower bound wasalso higher for lighter Majorana neutrinos. The second con-straint was due to two di ff erent collider searches. We foundthat these constraints place upper bounds on the coupling y t ψ .These constraints were also dependent on m ψ and y c ψ ; the lat-ter responsible for the size of the excluded region, since thiscoupling modifies the branching fraction of ψ → N R t .Our framework o ff ers an attractive avenue that naturally gen-erates small active neutrino masses while providing a largerange of masses for a viable dark matter candidate. The modelwe presented here is a minimal one as only couplings to t and c quarks are employed. The model also predicts new colour de-grees of freedom which lie below the TeV scale, and are nowbeing probed at the LHC. Further signatures at the LHC, suchas rare top quarks decays, monotop production and e ff ects onthe LHC Higgs signals, will be reported elsewhere.6 cknowledgements ADP would like to thank Jorge de Blas Mateo and TravisMartin for useful discussions and essential feedback regardingthe progress of this work. This work is supported in parts bythe National Science and Engineering Council of Canada.
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25. Theregion labeled 1 corresponds to regions of parameter space excluded by theCMS observable with MET >
200 GeV, while the regions labeled 2 and 3correspond to MET >
150 and >
300 GeV respectively. The plot on the bottomcorresponds to y c ψ = . y t Ψ m Ψ @ G e V D y t Ψ m Ψ @ G e V D Figure 8: Collider constraints for a 150 GeV Majorana neutrino from the fourexperimental observables mentioned at the beginning of this section. On thetop we show the excluded region in the y t ψ − m ψ plane for y c ψ = .
25. Theregion labeled 1 corresponds to regions of parameter space excluded by theCMS observable with MET >
200 GeV, while the regions labeled 2 and 3correspond to MET >
150 and >
300 GeV respectively. The plot on the bottomcorresponds to y c ψ = .5.