1 Introduction

Dark matter in U(1) extensions of the MSSM with gauge kinetic mixing

Geneviève Bélanger111 E-mail: belanger@lapth.cnrs.fr, Jonathan Da Silva222 E-mail: jonathan.da.silva.physics@gmail.com and Hieu Minh Tran333 E-mail: hieu.tranminh@hust.edu.vn

LAPTH, Universite Savoie Mont Blanc, CNRS, B.P.110, F74941 Annecy Cedex, France

Laboratoire de Physique Subatomique et de Cosmologie, Université Grenoble-Alpes,

CNRS/IN2P3, 53 Avenue Des Martyrs, F-38026 Grenoble, France

Hanoi University of Science and Technology, 1 Dai Co Viet Road, Hanoi, Vietnam

Abstract

The gauge kinetic mixing in general is allowed in models with multiple Abelian gauge groups. In this paper, we investigate the gauge kinetic mixing in the framework of extensions of the MSSM. It enlarges the viable parameter space, and has an important effect on the particle mass spectrum as well as the coupling with matters. The SM-like Higgs boson mass can be enhanced with a nonzero kinetic mixing parameter and the muon tension is slightly less severe than in the case of no mixing. We present the results from both benchmark analysis and global parameter scan. Various theoretical and phenomenological constraints have been considered. The recent LHC searches for the boson are important for the case of large positive kinetic mixing where the coupling is enhanced, and severely constrain scenarios with TeV. The viable dark matter candidate predicted by the model is either the neutralino or the right-handed sneutrino. Cosmological constraints from dark matter searches play a significant role in excluding the parameter space. Portions of the parameter space with relatively low sparticle mass spectrum can be successfully explored in the LHC run-2 as well as future linear colliders and dark matter searches.

## 1 Introduction

Although the standard model (SM) has been verified to a very high accuracy, an extension is necessary both for theoretical consistency and in order to explain experimental observations. The minimal supersymmetric (SUSY) extension of the SM (MSSM) has played an important role in phenomenological studies for many years because it could address many fundamental issues such as the gauge hierarchy problem, the prediction of the Higgs boson mass, and the gauge coupling unification while also providing a dark matter (DM) candidate, the lightest neutralino. Nevertheless, the discovery of a 125 GeV Higgs boson [1, 2] has imposed some tension on the MSSM. In order to reconcile the Higgs boson mass, large loop corrections are needed, these in general require heavy squarks (especially stops) and a large mixing, thus reintroducing a certain amount of fine-tuning to the theory [3]. Moreover, the parameter regions with maximal stop mixing which allow to obtain the observed Higgs mass potentially have a metastable electroweak vacuum, and predict a global minimum which breaks charge and/or color symmetries [4, 5, 6]. In scenarios where the number of free parameters is limited due to some relations at a high energy scale, heavy squarks imply heavy sleptons [7, 8]. Hence, the SUSY contribution to the muon anomalous magnetic dipole moment can hardly explain the discrepancy between the SM prediction and the experimental result. This last issue is however easily resolved when allowing a larger hierarchy between the slepton and squark masses.

There have been attempts to resolve these tensions, see for example [9, 10, 11] and references therein. In this paper, we consider extensions of the MSSM (UMSSM) that can also improve the situation [12, 15, 13, 14, 16]. The interaction between the extra singlet superfield, whose vacuum expectation value (VEV) breaks the , and the two Higgs doublets helps to increase the mass of the SM-like Higgs at the tree level. This contribution is the same as in the next-to-MSSM (NMSSM) [17, 18, 19, 20]. In addition, the SM-like Higgs boson mass is also enhanced by the D-term contribution [21, 22]. Both these effects imply that the loop-induced contribution from the stop sector does not need to be large. In this framework, as in the NMSSM, the -term problem is solved since this term is not introduced by hand but is generated by the vacuum expectation value of the singlet after the extra gauge group is broken. The physics origin of the group depends on the specific scenario, for instance , or inherited from some grand unified theory (GUT). Here we are interested in the scenario where the is a remnant symmetry after the breaking of the GUT [23]. This scenario is also motivated by superstring models [24, 25].

To account for the neutrino oscillations, we introduce in the model three generations of right-handed (RH) neutrinos. Assuming R-parity conservation, their superpartners which are weakly interacting massive particles can play the role of DM. This is in contrast with left-handed (LH) sneutrinos, which although also weakly interacting, have been ruled out as a DM candidate because their scattering cross section onto nuclei is too large [26]. This model offers two possible candidates for the DM, the ordinary neutralino and the right handed sneutrino, depending on which one is the lightest superparticle (LSP).

A special feature of models with two Abelian gauge groups , is that a gauge kinetic mixing term can exist in the Lagrangian without violating any underlying symmetry [27, 28, 29, 30]:

 L ⊃ −k2FμνF′νμ. (1)

Generally, even in the case that the kinetic mixing term is set to zero at some scale, it can be radiatively generated at the low energy scale due to the renormalization group (RG) evolution [31, 32]. It was found that in the case the gauge kinetic mixing effect can be significant and impact DM observables [33, 34, 36].

DM properties in U(1) extensions of the MSSM were examined in [37, 15, 38, 39] and the compatibility of the UMSSM with collider and DM observables was examined in [16] where the kinetic mixing was neglected. Here we revisit and update the constraints on the parameter space of the UMSSM inspired from GUT, while including the kinetic mixing. The radiatively generated kinetic mixing term depends on the particle content and the charge assignment of fields under the two gauge groups. For example, in the minimal SUSY model [34] the kinetic mixing parameter purely induced from the RG evolution is positive and sizable, , while in models [35] the value of k at low energies can be either positive or negative. To be completely general, we will consider that is a free parameter set at the low energy scale.

We will show that the gauge kinetic mixing can give rise to important effects on both the mass spectrum and DM properties. For example the kinetic mixing allows for a leptophobic which can more easily escape LHC constraints, gives a contribution to the mass of the Higgs boson, can shift the mass of sleptons thus providing a better agreement with the muon anomalous magnetic moment, and finally impact the DM annihilation channel. Note that in this study we include updated constraints from the LHC searches on a heavy neutral gauge boson as well as updated constraints from DM direct detection from LUX [40].

The structure of the paper is as follows. In Section 2, we briefly describe the UMSSM model with gauge kinetic mixing. The effects of the kinetic mixing term on the parameter space, the coupling with matter, and the mass spectrum are shown in Section 3. Here, benchmark analysis and results of the global parameter scan are presented with various collider constraints as well as cosmological ones taken into account. Section 4 is devoted for conclusion.

## 2 The UMSSM with gauge kinetic mixing

### 2.1 The model

The UMSSM has the gauge groups which remain after the symmetry breaking of an GUT. The particle contents of this model include the MSSM chiral supermultiplets, three generations of RH neutrino supermultiplets , the MSSM vector supermultiplets and an additional vector supermultiplet corresponding to gauge group, and a Higgs singlet superfield responsible for the breaking. Additional chiral supermultiplets are included in an anomaly free theory. For simplicity, we assume that all the fields belong to the 27 representations of that are not listed above are heavy enough to be safely neglected at low energies.

The charge of a chiral superfields is given by

 Q′ = cosθE6Q′χ+sinθE6Q′ψ, (2)

where parameterizes a linear combination of two subgroups and into . The charges and for each chiral superfield of the model are given in Table 1.

The superpotential of the model involves the ordinary MSSM superpotential without the -term, and other terms describing interactions of the Higgs singlet and right handed neutrinos:

 W ⊃ WMSSM|μ=0+λSHuHd+NcYνLHu, (3)

where is the neutrino Yukawa coupling matrix responsible for the neutrino mass generation. After the group is broken, the -term is generated by the singlet’s VEV, as

 μ = λvS√2. (4)

The soft SUSY breaking Lagrangian of the UMSSM reads

 Lsoft ⊃ LsoftMSSM|Bμ=0−(12M′1~B′~B′+~νcRAν~LHu+h.c.)−~νcRM2~νR~νR (5) −m2S|S|2−(λAλSHuHd+h.c.),

where new soft terms are added in comparison to the MSSM: the soft mass, , the neutrino trilinear couplings, , the right handed sneutrinos soft masses, , the singlino mass, , and the Higgs trilinear coupling, . Similar to Eq. (4), the MSSM term is induced by the breaking:

 Bμ = λAλvS√2. (6)

### 2.2 Gauge kinetic mixing

The general gauge kinetic Lagrangian for Abelian gauge superfields is written as follows

 Lgaugekinetic ⊃ −∫d4θ14(WαW′α)(1kk1)(WαW′α)+h.c., (7)

where the off-diagonal element is the gauge kinetic mixing parameter. The kinetic mixing matrix can be diagonalized by a rotation among the original Abelian vector superfields, ():

 (^V^V′) = (8)

For a real rotation, the kinetic mixing parameter is limited to . The rotation (8) ensures that there is no explicit kinetic mixing in the Lagrangian written in the new basis (). However, the effect of the kinetic mixing term now transfers to the interactions between the Abelian vector superfields and chiral superfields. The gauge interaction Lagrangian is

 Lgaugeinteraction ⊃ ∫d4θΦ†eQ⋅g⋅VΦ, (9)

where

 Q⋅g⋅V = (YQ′)(gYYgYEgEYgEE)(VYVE). (10)

where is the hypercharge and the charge associated with . The gauge coupling matrix which is originally diagonal absorbs the rotation of the Abelian vector superfields, and becomes non-diagonal: 444 In our analysis in the next section, we will assume for simplicity that

 (gYYgYEgEYgEE) = (11)

To simplify the gauge coupling matrix, we perform an orthogonal rotation in the space of Abelian vector superfields such that the gauge kinetic matrix remains intact:

 (VYVE) = 1√g2EE+g2EY(gEEgEY−gEYgEE)(VV′). (12)

Eq. (10) is then rewritten as

 Q⋅g⋅V = (YQ′)(gyg′0gE)(VV′), (13)

in which

 gy= gYYgEE−gYEgEY√g2EE+g2EY =g1, (14) g′= gYYgEY+gYEgEE√g2EE+g2EY =−kg1√1−k2, (15) gE= √g2EE+g2EY =g′1√1−k2. (16)

Note that in the limit , the above Abelian gauge coupling matrix becomes diagonal. Performing matrix multiplication in Eq. 13, we obtain:

 Q⋅g⋅V = Yg1V+QpgEV′, (17)

where the new charge is defined as

 Qp = Q′−kg1g′1Y. (18)

Clearly, corresponds to the SM hypercharge and is the associated gauge superfield while the kinetic mixing induces a shift in the new charge of the chiral superfields, from , and the coupling with the new Abelian superfield, from . It is worth to note that the anomaly cancellation conditions for in the underlying theory ensure the theory to be anomaly free for the redefined charge .

### 2.3 Neutral gauge bosons

The original Abelian vector superfields are mixed to form the new ones . Their vector components in turn mix with the third component of the gauge group to form mass eigenstates when the gauge groups are broken spontaneously. The -boson mixing mass matrix is as follows

 M2Z = (M2ZZM2ZZ′M2ZZ′M2Z′Z′), (19)

where

 M2ZZ = 14g21(v2u+v2d), M2Z′Z′ = g2E[(QpHu)2v2u+(QpHd)2v2d+(QpS)2v2S], (20) M2ZZ′ = 12g1gE(QpHuv2u−QpHdv2d).

This matrix can be diagonalized by an orthogonal rotation:

 (Z1Z2) = (cosαZsinαZ−sinαZcosαZ)(Z0Z′), (21)

where is the mixing angle defined as

 sin2αZ = 2M2ZZ′M2Z2−M2Z1 (22)

The physical states and have masses:

 M2Z1,Z2 = 12[M2ZZ+M2Z′Z′∓√(M2ZZ−M2Z′Z′)2+4M4ZZ′] (23)

In our analysis, we use the measured Z-boson mass for , while and are considered as free parameters.

### 2.4 Sfermions and neutralinos

In the UMSSM, the D-term contributions to sfermion masses play an important role in forming the sparticle mass spectrum. They modify the diagonal components of the usual MSSM sfermion mass matrices as

 Δ~f = 12g2EQp~f(QpHuv2u+QpHdv2d+QpSv2S), (24)

where with the generation index . Sine the redefined charges and gauge coupling are functions of , the sparticle mass spectrum also depends on the kinetic mixing parameter. As we will see, this effect is particularly important.

While charginos are the same as in the MSSM, the neutralino sector of the UMSSM consists of six fermions. Their masses are eigenvalues obtained from the mass matrix that is written in the basis of neutral fermionic components of the vector supermultiplets and the Higgs supermultiplets as

 M~χ0 = ⎛⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜ ⎜⎝M10−MZZcβsWMZZsβsW000M2MZZcβcW−MZZsβcW00−MZZcβsWMZZcβcW0−μ−λvu√2QpHdgEvdMZZsβsW−MZZsβcW−μ0−λvd√2QpHugEvu00−λvu√2−λvd√20QpSgEvS00QpHdgEvdQpHugEvuQpSgEvSM′1⎞⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟ ⎟⎠,

where , , with . The value of can be derived using Eqs. (21) and (22). We have

 cos2β = 1QpHu+QpHd⎛⎜ ⎜⎝sin2αZ(M2Z1−M2Z2)v2gE√g21+g22+QpHu⎞⎟ ⎟⎠, (26)

in which and is the SU(2) coupling. The matrix diagonalizing the above mass matrix determines the components of each neutralino:

 ~χ0i = (Zn)ijψ0j,i,j={1,2,3,4,5,6}, (27)

and therefore its properties.

### 2.5 Higgs sector

The tree level mass-squared matrix of CP-even Higgs bosons is a symmetric matrix with elements computed as:

 (M0+)11 = [g21+g224+(QpHd)2g2E]v2d+λAλvSvu√2vd, (M0+)12 = −[g21+g224−λ2−QpHuQpHdg2E]vuvd−λAλvS√2, (M0+)13 = [λ2+QpHdQpSg2E]vSvd−λAλvu√2, (28) (M0+)22 = [g21+g224+(QpHu)2g2E]v2u+λAλvSvd√2vu, (M0+)23 = [λ2+QpHuQpSg2E]vSvu−λAλvd√2, (M0+)33 = (QpS)2g2Ev2S+λAλvuvd√2vS.

The lightest Higgs boson is the SM-like one. Its tree level mass can be written approximately as [41]

 m2h1|tree ≃ M2ZZcos22β+12λ2v2sin22β+g2Ev2(QpHdcos2β+QpHusin2β)2 (29) −λ4v2g2E(QpS)2[1−Aλsin22β2μ+g2Eλ2(QpHdcos2β+QpHusin2β)QpS]2.

While the second term in the above equation is the same as in the NMSSM, the last two terms only appear in the UMSSM due to the existence of and related to the extra . Therefore the Higgs boson mass depends on the kinetic mixing parameter via these terms. Similarly the masses of and can receive large corrections due to the kinetic mixing. There is one CP-odd Higgs with the mass:

 m2A0|tree = λAλ√2sin2βvS(1+v24v2Ssin22β). (30)

The mass of the charged Higgs bosons is given by

 m2H±|tree = M2W+√2λAλsin2βvS−λ22v2. (31)

These masses also depend on through the angle , see Eq. 26.

## 3 Analysis

### 3.1 Theoretical constraints

In Eq. (17), we have interpreted as a redefined gauge coupling. It is crucial to check under which condition this new coupling satisfies the perturbation limit:

 αE=gE(k)24π≲1. (32)

Replacing with the coupling definition in Eq. 16, this condition leads to an upper bound on . This constraint is weak and only excludes the regions of close to .

In this model, is not chosen as an independent parameter as in the MSSM. It depends on the values of four other free parameters , , , and the kinetic mixing as expressed in Eq. (26). The reality condition on the angle ,

 0≤cos2β≤1, (33)

defines the regions in the parameter space of where further calculations can be carried out. In Figs 1, 2 and 3, we show the parameter regions allowed by the constraint (33). For a specific choice of , the kinetic mixing parameter is limited to a specific range that is usually smaller than the open range . Thus by allowing a nonzero kinetic mixing term, the acceptable ranges for other parameters change significantly.

First note that . Thus at , there is a unique value of that satisfies Eq. 26 for each choice of and . This can be seen in Fig. 1 where the allowed parameter regions in the plane for the case of GeV are depicted. Moreover this value of is large and positive for , Fig. 0(b). Given a choice of , for larger values of the range of allowed values for increases. The sign of is generally anticorrelated with that of for large values of the mixing to allow for a cancellation between the two terms in Eq. 26, except when . Moreover approaches 1 as the mixing increases. Note that for all cases where the first term in Eq. 26 dominates, the allowed regions are symmetric with respect to a sign flip of .

In Fig. 2, we plot the allowed regions in the plane for various values of and two choices of . In the limit of no mixing, , the range of values of become independent of and are only set by the conditions for , and for . Thus the non-zero kinetic mixing implies that regions of parameter space with small values of and small mixing are accessible while they were not with  [16]. However phenomenological constraints that will be discussed in the next section further restrict this region. For non-zero mixing angles, , larger values of are required to increase and compensate an increase in in the first term in Eq. 26. Note that the allowed range for is quite narrow at large values of and that the allowed regions in the plane become much larger for than for .

We also show in Fig. 3 the allowed regions in the plane for various values of the boson mass . Figs. 2(a) and 2(b) correspond to and 1.5 respectively. For , only a narrow range of mixing angles are allowed for , while for any value is allowed. Indeed in this case the first term in Eq. 26 becomes strongly suppressed. As mentioned above, for , a larger area of parameter space is theoretically consistent.

In summary, the presence of the kinetic mixing enlarges significantly the theoretically allowed regions of parameter space, in particular regions with small values of , large mixing and low boson mass. Moreover the large positive kinetic mixing () is slightly favoured as compared to large negative () as shown in Figs. 1, 2 and 3.

Besides the above theoretical constraints, we also impose perturbative Yukawa couplings, for this we require the Yukawa couplings to be smaller than at the SUSY scale. This constraint excludes the possibilities of very small or large values of . We also require that the width to mass ratios of Higgs particles should satisfy .

### 3.2 Phenomenological constraints

In our analysis, various phenomenological constraints are taken into account. For the Higgs boson mass, the combined result of the ATLAS and CMS measurements is employed [42] with a theoretical uncertainty of about 2 GeV. The deflection of the electroweak -parameter with respect to 1 is computed and compared to current upper bound [43]. We also consider the constraint on the muon anomalous magnetic moment [44, 45, 46]. A variety of constraints from flavor physics are taken into account. Observables in the -meson sector that are of interest include: the oscillation parameters , [47], the branching ratios of the following processes: [48], [49], [43], at low and high dilepton invariant mass [50], [51, 52], [47], [53], [47], [47], and the ratios , [54]. Observables in the Kaon sector include: the branching ratios of the processes [55], [56], the mass difference between and [43], and the indirect CP-violation in the system [43]. When calculating these observables, we take into account theoretical uncertainties as well as those from CKM matrix, rare decays, and hadronic parameters. The experimental limits of these constraints are as follows,

 122.1 GeV≤mh≤128.1 GeV, (34) Δaμ=aexpμ−aSMμ=(24.9±8.7)×10−10, (35) Δρ<8.8×10−4, (36) 17.715 ps−1≤ΔMs≤17.799 ps−1, [2σ] (37) 0.504 ps−1≤ΔMd≤0.516 ps−1, [2σ] (38) 0.70×10−4≤BR(B±→τ±ντ)≤1.58×10−4, [2σ] (39) 2.99×10−4≤BR(¯B0→Xsγ)≤3.87×10−4, [2σ] (40) 1.7×10−9≤BR(B0s→μ+μ−)≤4.5×10−9, [2σ] (41) 0.84×10−6≤BR(¯B0→Xsℓ+ℓ−)low≤2.32×10−6, [2σ] (42) 2.8×10−7≤BR(¯B0→Xsℓ+ℓ−)high≤6.8×10−7, [2σ] (43) 2.7×10−6≤BR(b→dγ)≤25.5×10−6, [2σ] (44) BR(B0d→μ+μ−)≤8.7×10−10, [3σ] (45) BR(B→Xsν¯ν)<6.4×10−4, [90%CL] (46) BR(B+→K+ν¯ν)<1.6×10−5, [90%CL] (47) BR(B0→K∗0ν¯ν)<5.5×10−5, [90%CL] (48) 0.299≤RD≤0.495, [2σ] (49) 0.259≤RD∗≤0.373, [3σ] (50) BR(K+→π+ν¯ν)<4.03×10−10, [2σ] (51) BR(K0L→π0ν¯ν)<2.6×10−8, [90%CL] (52) 5.275×10−3 ps−1≤ΔMK≤5.311×10−3 ps−1, [2σ] (53) 2.206×10−3≤ϵK≤2.250×10−3. [2σ] (54)

Various constraints from direct searches for new particles at colliders are relevant for the scenarios we consider. While scenarios with light sfermions are severely restricted by LEP, the LSP can be light enough to contribute to the invisible decay width, we impose the constraint  MeV [57].

Searches for a heavy neutral gauge boson in the dilepton and dijet channels have been performed at the LHC both at 8 TeV and 13 TeV. We use the most recent data on the dilepton final state corresponding to an integrated luminosity of 3.2 fb at TeV [58]. Here, the mass limit is interpolated for each specific value of . Limits from the dijet resonance searches at the LHC are obtained with the method described in Ref.  [59] using a combination of ATLAS [60, 61] and CMS [62, 63] dijet data at 8 TeV and 13 TeV. These constraints are included in micrOMEGAs4.3 [64]. In the UMSSM, there are cases where the lightest chargino is long-lived, typically when the chargino is nearly pure wino or nearly pure higgsino. For such points, we take into account the results from long-lived chargino searches at the Tevatron and the LHC. To derive this constraint, the observed limits for the cross sections of long-lived chargino pair production at D0 [65] experiment are employed in combination with the observed limit for chargino pair production and neutralino-chargino production cross section at the ATLAS [66] experiment. We follow the procedure described in [16].

In addition to the constraints from collider physics, we take into account those from cosmological observations. The most recent measurement of the DM relic density by Planck experiment [67] reads

 ΩCDMh2 = 0.1188±0.0010. (55)

In the global parameter scan, we impose only an upper bound on the DM relic density of the LSP. Thus we implicitly assume that there could be an additional DM candidate.

For DM direct detection the LUX experiment sets the most severe constraint on the spin-independent (SI) cross section between a DM particle and nucleons [68, 69, 40], while PICO-60 [70] sets the best direct limit on the spin-dependent (SD) cross section on protons. The SD cross section on protons is also constrained by IceCube [71] by observing the neutrino flux from DM captured in the Sun, this limit however depends on specific annihilation channels.

The UMSSM model with kinetic mixing was implemented in LanHEP version 3.2.0 [72, 73, 74] which produces the model files suitable for CalcHEP [75]. The spectrum and all the DM observables are calculated using micrOMEGAs version 4.3.1 [15, 16, 76, 64] with the help of UMSSMTools [77] adapted from NMSSMTools v5.0.2 routines [78, 79]. The latter includes in particular all flavour physics observables. For collider observables, we use a routine of micrOMEGAs to compute the limits from LHC as well as the invisible width. An interface to HiggsBounds [80] allows to test the Higgs sector of the model with respect to CL exclusion limits from the LEP, Tevatron and LHC experiments. Finally the points satisfying all the above collider constraints are analysed with SModelS 1.0.4 which decomposes the signal of any BSM model into simplified topologies in order to test it against LHC bounds [81, 82].

### 3.3 Benchmark analysis

We examine the effect of the kinetic mixing on the sparticle spectrum for a benchmark set of the UMSSM inputs. The simplified UMSSM input parameters are taken to be: the common gaugino masses TeV, the common slepton and squark soft masses TeV and TeV respectively, the boson mass TeV, the common trilinear coupling TeV, the -parameter GeV, the mixing angle between two -bosons , and the angle . Note that these values of , and are chosen randomly such that all the phenomenological constraints, especially the 125 GeV Higgs boson mass and the DM relic abundance, can be satisfied for a suitable value of . Letting the kinetic mixing parameter to be a free input, we find that the range with is theoretically acceptable. The values outside this range are excluded by the reality condition (33) and the tachyonic slepton condition.

In Fig. 4, we show the dependence of slepton masses of the first generation on the kinetic mixing parameter. For this particular choice of inputs, the behaviors of the second and third slepton generations are very similar. The two sfermions belonging to the LH slepton doublet, and , have masses too degenerate to be distinguished in the plot. When increasing the kinetic mixing, the LH slepton masses increase while the RH selectron mass decrease, becoming tachyonic for . The RH sneutrino mass is nearly independent on .

Fig. 5 shows the first generation squark masses as functions the kinetic mixing parameter. Here, only the RH up-squark becomes heavier for larger . The other squark masses (LH up-squark, LH down-squark, and RH down-squark) decrease with the kinetic mixing parameter . As for the slepton case, the other two generations of squarks have a similar behavior as the first generation and are therefore not shown in the figure.

The -dependence of sfermion masses can be explained using Eqs. (24), (16), and (18). Within the allowed range of , corrections to the sfermion masses are dominantly controlled by . For the benchmark value , the quantity in the brackets of the right side of (24) is positive. Therefore, the D-term correction to a sfermion mass is approximately proportional to and its hypercharge . The dependence on is therefore stronger for the sparticle with a large hypercharge . The mass increases (decreases) with for negative (positive) . The RH sneutrino has a hypercharge , hence its mass remains almost constant. The kinetic mixing enters the neutralino masses only through the mixing between higgsinos, singlino and bino’, hence the neutralino masses are almost independent of the kinetic mixing.

The SM-like Higgs boson mass is plotted as a function of the kinetic mixing in Figs (6). We see that the Higgs boson mass decreases with and that in the absence of kinetic mixing the mass would be much below the observed value. Thus, enabling a negative nonzero kinetic mixing, in this case , allows to bring the Higgs boson mass in agreement with the observed value, GeV.

For illustration, we show in Tables 2, the sparticle mass spectrum as well as the constrained observables for the benchmark just discussed. Assuming theoretical uncertainties in the calculations, we find that only the muon g-2 and satisfy the corresponding constraints at level, while all other observables comply with the experimental limits at level. The LEP limits, the invisible width, and the dilepton and dijet constraints for the boson from the LHC, the constraint from long-lived chargino searches at D0 and the ATLAS experiments are all satisfied. This benchmark is also compatible with limits on the Higgs sector obtained by Lilith and HiggsBounds as well as with limits on sparticles obtained with SModelS. We note that the kinetic mixing induces large shifts in the heavy Higgs doublet, from  TeV when to  TeV when while the singlet mass, in Table 2, remains constant. Such heavy masses are in any case out of reach of the LHC. The DM candidate for this benchmark is a higgsino-like neutralino. Its relic density is achieved by annihilation into gauge bosons and coannihilation with the second lightest neutralino and the chargino NLSP whose masses are almost degenerate. The SI and SD cross sections of the DM scattering on nuclei meet the requirement from the LUX and IceCube experiments. Since the sparticles are quite heavy, it is challenging to test this benchmark at the LHC. However, the future XENON1T will be able to test the model via the SI interaction of the neutralino DM.