Status of the Zee–Babu model for neutrino mass andpossible tests at a like-sign linear collider

Status of the Zee–Babu model for neutrino mass and
possible tests at a like-sign linear collider

Daniel Schmidt danielschmidtip@t-online.de Max-Planck-Institut für Kernphysik, Saupfercheckweg 1, 69117 Heidelberg, Germany    Thomas Schwetz schwetz@fysik.su.se Max-Planck-Institut für Kernphysik, Saupfercheckweg 1, 69117 Heidelberg, Germany Oskar Klein Centre for Cosmoparticle Physics, Department of Physics, Stockholm University, SE-10691 Stockholm, Sweden    He Zhang he.zhang@mpi-hd.mpg.de Max-Planck-Institut für Kernphysik, Saupfercheckweg 1, 69117 Heidelberg, Germany
July 15, 2019
Abstract

We provide an updated scan of the allowed parameter space of the two-loop Zee–Babu model for neutrino mass. Taking into account most recent experimental data on as well as the mixing angle we obtain lower bounds on the masses of the singly and doubly charged scalars of between 1 to 2 TeV, with some dependence on perturbativity and fine-tuning requirements. This makes the scalars difficult to observe at LHC with 14 TeV even with optimistic assumptions on the luminosity, and would require a multi-TeV linear collider to see the scalar resonances. We point out, however, that a sub-TeV linear collider in the like-sign mode may be able to observe lepton flavour violating processes such as due to contact interactions induced by the doubly charged scalar with masses up to around 10 TeV. We investigate the possibility to distinguish the Zee–Babu model from the Higgs triplet model using such processes.

I Introduction

Non-zero neutrino mass requires an extension of the Standard Model (SM). Among the plenitude of possibilities, an attractive way to explain the smallness of neutrino masses is to invoke loop processes, see for instance Farzan et al. (2013); Angel et al. (2013); Law and McDonald (2013) for recent discussions. Then the scale of the new physics responsible for generating neutrino mass can be not too far from the TeV range, which makes those type of models potentially testable at colliders and/or in experiments searching for charged lepton flavour violation (LFV). An economical way of radiative neutrino mass generation is to enlarge the scalar sector of the SM Konetschny and Kummer (1977); Cheng and Li (1980). In this work we concentrate on a particularly simple model of this kind, namely the so-called Zee–Babu model Zee (1985, 1986); Babu (1988). In this model two singlet scalars are introduced, one singly charged and one doubly charged, and neutrino masses are generated at two-loop level. Through the exchange of heavy scalars, lepton flavour violating processes such as can become observable and the new scalars could be accessible at colliders. In particular, the doubly charged scalar may induce very clean like-sign bi-lepton events. Possible connections to Dark Matter within this model have been discussed in Lindner et al. (2011); Baek et al. (2012).

In this paper we provide an update of previous phenomenological studies of the Zee–Babu model Babu and Macesanu (2003); Aristizabal Sierra and Hirsch (2006); Nebot et al. (2008); Ohlsson et al. (2009), motivated by various new experimental results relevant for this model. First, precision measurements on reactor neutrinos Abe et al. (2012); An et al. (2012); Ahn et al. (2012) have confirmed that the smallest neutrino mixing angle is non-vanishing and close to the previous upper bound, i.e.,  Gonzalez-Garcia et al. (2012). Second, in 2013 the MEG collaboration has provided a new upper limit on the LFV process , with a branching ratio less than Adam et al. (2013). We perform a parameter scan of the model taking into account up to date constraints on various LFV and other low-energy processes as well as neutrino oscillation experiments. As a consequence we find that most likely the charged scalars of the Zee–Babu model will be out of reach for the Large Hadron Collider (LHC), including the 14 TeV configuration. Below we comment on the possibilities to observe them indirectly through LFV processes at the proposed International Linear Collider (ILC) in the like-sign mode.

An alternative way to generate neutrino masses is the so-called Higgs triplet model, where an triplet scalar is introduced, which couples to the lepton doublets and gives rise to a neutrino mass term from the vacuum expectation value of the neutral component Konetschny and Kummer (1977); Cheng and Li (1980); Schechter and Valle (1980); Lazarides et al. (1981); Mohapatra and Senjanović (1981). If the triplet mass is in the TeV range the doubly charged component could be produced at colliders through the Drell–Yan process, and subsequently decay to lepton pairs, leading to similar signatures as the doubly charged scalar in the Zee–Babu model. If a doubly charged scalar should be found at a collider below the lower bounds in the Zee–Babu model discussed below, it may point towards the Higgs triplet model. In contrast, if no resonance is found the triplet can lead to similar LFV processes at a like-sign electron collider as the Zee–Babu scalar. However, due to the different mechanisms to generate neutrino masses, the specific flavour structure of those processes are distinctive in the two models. We discuss possibilities to distinguish the two models, once such LFV events were observed at a future collider.

The outline of the paper is as follows: In Sec. II, we present the framework and characteristic features of the Zee–Babu model. In Sec. III, we focus on the low-energy processes mediated by the doubly charged scalar and summarize the current constraints on the relevant Yukawa couplings. Numerical analyses on the model parameters are given in Sec. IV. In particular, we illustrate the allowed ranges of the scalar masses. We further discuss in Sec. V signatures of the doubly charged scalar at a future linear collider. The discrimination between the Zee–Babu model and the triplet model is investigated in Sec. VI. Finally, in Sec. VII, we summarize our results and conclude.

Ii The Zee–Babu Model

The particle content of the Zee–Babu model is that of the SM extended with two complex singlet scalars, a singly charged scalar and a doubly charged scalar , which couple to left-handed lepton doublets and right-handed lepton singlets , respectively. The contribution to the Lagrangian is

(1)

where the scalar potential contains additional couplings among scalar fields. The presence of the tri-linear term together with the two Yukawa-type terms in the first line of Eq. (1) implies that lepton number is violated.111In the original Zee–Babu model as displayed in Eq. (1), the trilinear term which violates lepton number by two units has to be introduced “by hand”. It is possible to have instead a lepton number conserving interaction which generates the term by spontaneous symmetry breaking, see e.g., Chang et al. (1988); Lindner et al. (2011). A Majorana mass term for neutrinos is generated via a two-loop diagram, yielding

(2)

where are charged lepton masses and is a two-loop integral McDonald and McKellar (2003), which approximates to

(3)

Here, and is a dimensionless function of order unity, see e.g., Nebot et al. (2008). Note that the charged scalars couple only to leptons and not at all to hadrons. Therefore, they might contribute to for instance the Fermi constant for leptonic processes, and hence lepton–hadron universality tests provide constraints on the couplings of the scalars, see section III.

Since is an antisymmetric matrix in flavour space, we have , and hence one of the light neutrinos is massless. The neutrino mass eigenvalues are obtained by diagonalization of (2) by means of the unitary matrix :

(4)

where correspond to the elementary rotations in the , , and planes (parametrized in what follows by three mixing angles, with and ), and and contain the Dirac and Majorana CP phases, respectively. Here only one Majorana phase is involved, since one neutrino is massless. Depending on the neutrino mass ordering, either (normal ordering, NO) or (inverted ordering, IO) is zero. The non-zero neutrino mass states are then determined by the solar and atmospheric mass-squared differences and , where .

Using the antisymmetricity of (the couplings of ), they can be expressed in terms of the neutrino mixing angles Babu and Macesanu (2003); Aristizabal Sierra and Hirsch (2006); Nebot et al. (2008). In the NO case, we have

(5)
(6)

Since is relatively small compared to the other mixing angles, we can neglect the second terms in the above expressions, and obtain the approximate relation

(7)

by assuming and . For the IO case, the two non-trivial equations are

(8)
(9)

which imply

(10)

Using Eq. (2), the Yukawa couplings of the doubly charged scalar are related to the neutrino mass matrix elements as

(11)

where (no sum) with being the charged lepton masses and is a numerical factor stemming from the loop function.

Iii Experimental Constraints

The experimental bounds on the Zee–Babu model mainly come from lepton flavour violating processes at low-energy scales mediated by the heavy scalars, and the universality of weak interactions. In this section, we summarize the relevant low-energy scale experimental limits on the Zee–Babu model.

  • Lepton flavour violating decays , which are mediated by the doubly charged scalar at tree level. The branching ratio is given by with

    (12)
  • Universality in decays: The Fermi coupling constant measured in muon and tau decays obtains corrections from the exchange of , i.e.,

    (13)

    Furthermore, by assuming the unitarity of the CKM matrix, one can test the universality of the couplings in hadronic and leptonic decays, which gives

    (14)

    In Eqs. (13) and (14) is the Fermi coupling constant as given by the SM contribution. We show only the leading terms in the couplings , which emerge from the interference of the SM diagram with the ones mediated by the Zee–Babu scalars.

  • Rare lepton decays: (for ) can be mediated at one-loop level by both and , and the branching ratios read , where

    (15)
  • Muonium to antimuonium conversion through the exchange of : The process is well bounded experimentally, leading to constraints on the effective coupling related to the following four-fermion operator

    (16)
  • Muon and electron anomalous magnetic moments: obtains addition contributions from both and , with

    (17)

    where . The bound from is very weak (only relevant for scalar masses above  TeV) and therefore we include only the constraint from .

  • conversion in nuclei: The loops which mediate the decays generate an effective vertex which induces conversion in nuclei. Using the result from Kitano et al. (2002) we obtain

    (18)

    where and represent overlap integrals of the muon and electron wave functions. The form factors are given by the same expressions as in the case of the Higgs triplet model Dinh et al. (2012)

    (19)

    where the loop function is Ma et al. (2001)

    (20)

    Note that in the Higgs triplet model both the singly and doubly charged scalars couple to left-handed leptons (since both are components of the same triplet field), whereas in the Zee–Babu model couples to left-handed and couples to right-handed leptons, see Eq. (1). Therefore, the amplitudes for singly and doubly charged scalar mediated processes do not interfere in the case of the Zee–Babu model, whereas they do in the case of the Higgs triplet model Kitano et al. (2002).

Constraint Ref. Bound ( C.L.)
Beringer et al. (2012)
universality Beringer et al. (2012)
universality Beringer et al. (2012)
universality Beringer et al. (2012)
Bennett et al. (2006); Beringer et al. (2012)
BR Bellgardt et al. (1988)
BR Hayasaka et al. (2010)
BR Hayasaka et al. (2010)
BR Hayasaka et al. (2010)
BR Hayasaka et al. (2010)
BR Hayasaka et al. (2010)
BR Hayasaka et al. (2010)
BR Adam et al. (2013)
BR Aubert et al. (2010)
BR Aubert et al. (2010)
conversion Bertl et al. (2006) see Eq. (19)
Beringer et al. (2012)
Table 1: Summary of experimental constraints and the corresponding bounds on the Yukawa couplings. Here , and and so on.

We summarize in Table 1 the low-energy experimental constraints used in our analysis. One can observe that lepton flavour violating processes set more stringent bounds on the Yukawa couplings, in particular the and decays. The later process could however be suppressed in the Zee–Babu model if or is vanishing, which is possible while still obtaining a valid neutrino mass matrix. The decay is mediated by both singly and doubly charged scalars, and is proportional to both Yukawa couplings and , which cannot vanish simultaneously. Therefore, the most stringent constraint on the Zee–Babu model stems from the decay.

Iv Numerical Analysis

We perform a scan of the Zee–Babu model parameters confronting the experimental data in order to obtain constraints on the scalar masses. The independent parameters can be chosen as: three leptonic mixing angles; Dirac and Majorana phases , ; the Yukawa couplings , , , ; scalar masses , ; and the parameter in the scalar potential. Neutrino masses are fixed to the values set by the best fit mass-squared differences Gonzalez-Garcia et al. (2012). Then the remaining Yukawa couplings and are fixed by Eqs. (5), (6), (8), (9), (II). For simplicity we fix the mixing angles to the values Gonzalez-Garcia et al. (2012) , , . The remaining parameters are scanned in the following ranges:

(21)

where parametrizes the requirement of perturbativity of Yukawa couplings. If not stated otherwise we take . For the tri-linear term, the parameter induces loop corrections to the scalar masses as . In the absence of fine-tuning the correction should be smaller than the tree-level masses, which leads to the constraint . Henceforth, we parameterize this requirement by a parameter , see Eq. (IV). The phases of are chosen randomly, whereas can be taken real without loss of generality. For a given set of the parameters in Eq. (IV) we check if all other values for and are less than ; if not then the point is discarded. If the perturbativity constraint is fulfilled we compare the model predictions to the experimental data with a function

(22)

where represents the data of the th experimental observable, the corresponding 1 absolute error, and the prediction of the model. The index runs over the 17 experimental observables given in Table 1. In case of upper bounds we set and use the upper bound for . In order to identify the allowed regions in parameter space we proceed as follows. For a given point in parameter space we consider the maximum of all data points:

(23)

If is fulfilled we keep the point, otherwise it is discarded. In that way we make sure that all data points are fitted within 2 standard deviations. Let us stress that we do not adopt any particular statistical interpretation of the resulting regions in parameter space in terms of confidence regions, apart from the above statement that all constraints are satisfied within .

From Eqs. (II) one can see that the contribution of the couplings (, ) is suppressed by two powers (one power) of the electron mass. Indeed, we find always viable solutions for . However, we do take into account finite values in our scan in order to allow for sub-leading effects induced by those couplings.

Figure 1: The shadowed regions correspond to allowed ranges of the scalar masses for the normal mass ordering (left panel) and the inverted mass ordering (right panel) by requiring and imposing the perturbativity criterion . The black, blue, and red curves correspond to lower limits on the scalar masses obtained from the experimental data by requiring but without the constraint. The black curve corresponds to the current experimental bounds. The blue solid and blue dashed curves show the exclusion regions from the expected conversion constraint and , respectively. The red line is given by assuming BR. Furthermore, gray curves delimit the region allowed by perturbativity without requiring that the experimental constraints are respected.

The allowed ranges of scalars masses are illustrated in Fig. 1 for both normal and inverted mass orderings. For scalar masses within the shadowed regions all the constraints are satisfied in the sense of and as explained above. We observe that the parameter space of the model is closed, however, allowing for scalar masses up to (100 TeV) are possible. If we allow for some fine-tuning in the scalar potential by setting the parameter to , the upper bound on the scalar masses will be larger than  TeV (see Fig. 2 below). Note however, that large scalar masses require Yukawa couplings close to the perturbativity limit. If Yukawa couplings assume values the scalar masses get pushed towards lower values. The lower bound on the scalar masses (black curve) is dominated by the observables from Tab. 1, most importantly from the MEG bound on . We obtain the following lower bounds by requiring that all constraints are satisfied at :

(24)

Note that the lowest possible value for the doubly charged scalar occurs for relatively large values of singly charged scalar mass and depends also on the perturbativity/fine-tuning conditions. In deriving the bounds we have assumed . If we allow values for the tri-linear coupling larger than the scalar masses (amounting to some fine-tuning in the scalar potential) the lower bounds on the Zee–Babu scalars can be relaxed. For example, we show the mass ranges in Fig. 2 by taking a relatively large constraint . One can see that the lower bounds on the scalar masses reduce to

(25)

The bounds for IO can be further weakened by fine tuning of the complex phases and , leading to a cancellation between different terms in Eq. (II). By performing a dedicated search with phases constrained to be very close to those special values we find that the IO bounds for and in Eq. (24) reduce to around 1.0 and 1.1 TeV, respectively. Let us also stress that our bounds are obtained by a random parameter scan, throwing points for a given choice of and , out of which only a fraction passes our perturbativity requirement. With such a method fine tuned solutions as the one mentioned above might be missed. In this sense our bounds for the scalar masses hold for “generic” values of the parameters.

Figure 2: The same as Fig. 1 with , i.e., allowing for a tri-linear coupling larger than the scalar masses, see Eq. (IV).

There are several future projects aiming for improving significantly the bound on conversion by about 4 to 5 orders of magnitude compared to the current limit Bertl et al. (2006), for instance the Mu2e Abrams et al. (2012); Knoepfel et al. (2013) and COMET Cui et al. (2009) experiments aim at sensitivities of order , whereas the target sensitivity of the PRISM project Witte et al. (2012) is even of order . Furthermore, an upgrade program is underway for the MEG experiment aiming at a sensitivity improvement of a further order of magnitude Adam et al. (2013). We thus also show in Figs. 1 and 2 with colored contours the future experimental constraints on the charged scalars. Those improved constraints on LFV (if no positive signal is found) will further push up the lower bounds on the scalar masses of about 1 order of magnitude. However, it is still difficult to entirely rule out the Zee–Babu model, no matter for the normal or inverted mass ordering (depending on the perturbativity requirements).

Figure 3: Allowed ranges of the scalar masses for the inverted mass ordering for by requiring . The black curve corresponds to the current experimental bounds. The shadowed area contour is allowed by the perturbativity criterion (without imposing the constraints from Tab. 1).

It should be noticed that our analysis is based on the latest measurement on the smallest mixing angle , which is indeed very crucial for the IO case. For illustration purposes, we show in Fig. 3 also a similar numerical analysis but using (which is by now excluded by oscillation data). In this case, there is no overlap between the perturbativity contours and the experimental contours, indicating that the Zee–Babu model is incompatible with the inverted mass ordering if is very small. In other words, if the neutrino mass ordering is inverted, the Zee–Babu model can then be viewed as a natural candidate for predicting sizeable , see e.g., Nebot et al. (2008). Such a feature is also manifest in Eqs. (8) and (9), showing that both, and are inversely proportional to , and therefore a too small may blow up Yukawa couplings resulting in a conflict with lepton flavour violation constraints. Conversely, there is no inverse dependence on in the case of normal mass ordering, c.f. Eqs. (5) and (6), and thus the choice of has no important impact on the Zee–Babu model parameters. We have checked this point numerically by choosing different values for , and they all give almost the same scalar mass ranges for normal ordering.


As discussed above, present data pushes the masses of the singly and doubly charged scalars of the Zee–Babu model above the TeV, see Eq. (24). This makes the direct production at colliders difficult. The production cross sections of the Zee–Babu scalars have been calculated in Babu and Macesanu (2003); Aristizabal Sierra and Hirsch (2006); Nebot et al. (2008), for recent work on di-leptons in general see e.g., del Aguila et al. (2013); Alloul et al. (2013). ATLAS Aad et al. (2012) and CMS Chatrchyan et al. (2012) have searched for doubly charged scalars decaying predominantly into muons and/or electrons based on approximately 5 fb at  TeV, obtaining lower bounds on their mass of around 400 GeV. The results of Nebot et al. (2008) show that for the Zee–Babu doubly charged scalar with masses  TeV of order 300, 10, 1 events are expected at LHC for 300 fb at  TeV, respectively, assuming 100% branching fraction of into leptons. Given the constraints on the masses derived above this implies that most likely the Zee–Babu scalars will not be observable at LHC, unless some degree of fine-tuning is accepted and the relaxed bounds of Eq. (25) apply. In that case, future experiments for LFV would also observe a positive signal, see Fig. 2.

V Tests at an collider

At a possible future linear collider the scalars can be pair-produced by photon and exchange: . Obviously this requires center of mass energies of , which in view of the bounds in Eq. (24) seems not realistic in the foreseeable future. However, a linear collider may also be operated in the like-sign mode. This offers a new window to search for LFV processes within the context of the Zee–Babu model for scalar masses up to  TeV.

The possibility to test bi-leptons at a like-sign electron collider has been considered since long time, see Refs. Rizzo (1982a, b); London et al. (1987); Cuypers and Raidal (1997); Raidal (1998); Cakir (2006); Rodejohann (2010); Rodejohann and Zhang (2011) for an incomplete list of references. Lots of work has been devoted to the search for lepton number violating reactions in the context of Higgs triplet models. In the Zee–Babu model this process is not allowed at tree level. However, lepton flavour violating (but lepton number conserving) reactions such as mediated by at tree level may be observable.

Figure 4: Upper limit on the cross section as a function of the center of mass energy for normal neutrino mass ordering for doubly charged scalar masses of 2, 2.5, and 3 TeV.

The cross section for the process for is given by

(26)

where is the width of the doubly charged scalar, and the polarization factor of the incoming electron beams. First, we note that this process is proportional to . As mentioned above, this coupling is not determined by neutrino data and in principle it could be zero. Hence, no signal can be predicted. On the other hand, also the upper bound on this coupling is rather weak and therefore a sizeable cross section would be possible in principle. The upper limit on the cross section as a function of the center of mass energy is shown in Fig. 4 for unpolarized beams (). For center of mass energies of a sharp resonance can be observed, leading to very large cross sections in excess of 100 pb, allowing for the direct discovery of the doubly charged scalar. In view of the bounds from Eq. (24), this will require a multi-TeV collider. However, for lower center of mass energies, one may still expect visible cross sections, corresponding to contact interactions mediated by the heavy scalar.

Figure 5: Upper limits on the sum of the cross sections as a function of the doubly charged scalar mass for center of mass energies  GeV. Solid curves correspond to the Zee–Babu model, where dominates the sum. Dashed curves correspond to the Higgs triplet model, where the lightest neutrino mass is allowed to vary between 0 and 0.2 eV. The left (right) panel corresponds to normal (inverted) neutrino mass ordering.

We show in Fig. 5 the upper limits on the cross section as a function of the scalar mass for center of mass energies  GeV (solid curves). For a total luminosity of 50 , more than a few tens of events can be expected for a scalar mass . Also note that, for a smaller , the reduction of the cross section for IO is due to the LFV constraints. While such a signature will not allow for the direct discovery of the doubly charged scalar via a resonance, it would provide indirect evidence for a doubly charged particle. The flavour and chirality structure of the LFV processes would offer additional consistency checks with the Zee–Babu model, as we are going to discuss in the next section.

Vi Distinguishing the Zee–Babu and the Higgs triplet model

Once a signal induced by a doubly charged scalar is found at a collider experiment, an interesting task will be to identify the underlying model and establish the connection to the mechanism to generate neutrino mass. Here we focus on ways to distinguish the Zee–Babu model from the Higgs triplet model, which also predicts the existence of a doubly charged scalar particle. The leptonic part of the Lagrangian contains the term

(27)

where is a symmetric Yukawa coupling matrix and is a representation of the Higgs triplet containing neutral, singly charged, and doubly charged components. Neutrino masses are generated by the vacuum expectation value (VEV) of the neutral component. Hence the Yukawa couplings are directly proportional to the neutrino mass matrix. Below we outline a few possibilities to distinguish the Zee–Babu model from the Higgs triplet model.

vi.1 Signatures at a like-sign linear collider

Similar as in the Zee–Babu model, also in the Higgs triplet model the process is possible at a like-sign linear collider with a cross section in complete analogy to Eq. (26) (see e.g., Rodejohann and Zhang (2011)), with the coupling replaced by and , taking into account that for the triplet left-handed leptons couple to the doubly charged scalar, in contrast to the right-handed coupling in the Zee–Babu model. In Fig. 5 we compare the maximum obtainable value for the sum of the cross sections for lepton flavour violating processes for the Higgs triplet model (dashed) to the one for the Zee–Babu model (solid). For the Higgs triplet model, the lightest neutrino mass is not necessarily vanishing, therefore, in Fig. 5 we vary its value between zero and 0.2 eV. The VEV of the triplet is varied between 0.1 eV and 1 keV, where the largest cross sections are obtained for small values, since in this case Yukawa couplings are largest. The two Majorana phases are allowed to vary freely between 0 and , and the Dirac phase between 0 and . As for the Zee–Babu model, we use the current experimental bounds, i.e., the same constraints as for the black curves in Fig. 1. For the triplet model we include the constraints from , , muonium–antimuonium conversion, and conversion in nuclei. The corresponding expression can be found for instance in Ma et al. (2001); Kitano et al. (2002); Dinh et al. (2012). We find cross sections of order 1 fb, and those results suggest that with integrated luminosities of  fb such lepton flavour violating processes can also be expected to be observed in the Higgs triplet model. Note that here we are in the regime of much smaller than the mass of the doubly charged scalar, which implies that no resonance is seen and hence the mass cannot be determined. Furthermore, we stress again that those curves are upper bounds, and in particular, for the Zee–Babu model the cross section can be easily reduced by adjusting which is not bounded from below. Therefore the size of the cross section by itself does not allow to distinguish the two models.

Figure 6: Upper limits on the sum of the cross sections in the Higgs triplet model as a function of the lightest neutrino mass for center of mass energies  GeV. The doubly charged scalar mass is set to 2 TeV. The left (right) panel corresponds to normal (inverted) neutrino mass ordering.

Before we discuss the possibility to use the flavour structure to distinguish the models, let us mention the importance of the lightest neutrino mass in the case of the triplet model. Since the cross section for is proportional to the value of the lightest neutrino mass is important, especially for normal mass ordering, where the possible size of depends strongly on . In Fig. 6 we show the maximum obtainable value for the sum of the cross sections for lepton flavour violating processes as a function of the lightest neutrino mass. We observe that for strongly hierarchical spectrum with normal ordering the cross section becomes very small. Hence, establishing normal mass ordering by oscillation experiments plus setting an upper bound on the lightest neutrino mass below 0.05 eV (for instance by cosmology or neutrinoless double beta decay) would make the signal in the triplet model very small. In that case a sizeable signal at the like-sign collider would favour the Zee–Babu model. Note also that even for large neutrino masses no relevant lower bound on the cross section can be derived in the triplet model, since Yukawa couplings can be made very small by increasing the VEV of the triplet above the keV range (up to the GeV scale).

vi.2 Flavour structure of the Yukawa couplings

Let us now assume that either a leptonically decaying doubly charged resonance is found at LHC or a linear collider, or that lepton flavour violating processes are seen at a like-sign linear collider. In such a case the flavour structure of the decays or the LFV processes will be rather different in the two cases. Note that ratios of decay rates and LFV cross sections will be the same and proportional to the corresponding Yukawa couplings:

(28)

where in the Zee–Babu model we have

(29)

and in the triplet model an analogous relation holds but replacing the Yukawa coupling by . The important observation is that the flavour structure of those couplings will be rather different in the two models, with proportional to the neutrino mass matrix , while for the relation to the neutrino mass matrix is more complicated.

Figure 7: Cross section ratios for processes (or equivalently ratios of branching fractions of the doubly charged scalar decays). The left (right) panel corresponds to normal (inverted) neutrino mass ordering. Black points denote the ratios in the Zee–Babu model, while red points correspond to the ratios in the Higgs triplet model. All scalar masses are taken to be . For the triplet model we vary the Majorana phases, the lightest neutrino mass between zero and , and the triplet VEV between 0.1 eV and 1 keV.

We illustrate in Fig. 7 the cross section ratios of the LFV processes (or equivalently decay branching fractions) in the two models. For the sake of definiteness, all the scalar masses are taken to be . Those results depend only weakly on the scalar masses. The different flavour structure of the two models can be clearly seen from the plots. In particular, in the Zee–Babu model, the dominating decay mode is always the channel, no matter of the neutrino mass ordering. An observation of a significant fraction of events different from the di-muon channel would exclude the model. The largest contribution of a different flavour combination may occur for IO with a fraction of events with .

In order to see this point more clearly, we insert the neutrino mixing parameters into Eq. (II) and obtain for the NO case

(30)

where (defined after Eq. (II)) and terms proportional to the small parameters , , and have been neglected. Therefore, one has approximately for a nearly maximal , and the ratios between Yukawa couplings are

(31)

Since the branching ratios are proportional to , one finds and , in good agreement with the left plot of Fig. 7. Using the first line in Eq. (VI.2) with  eV, one finds that is bounded from below by neutrino masses Nebot et al. (2008). Numerically we find

(32)

Taking all the scalar masses to be 3 TeV as an example, the above condition indicates , where we have used Eq. (7) and the constraint from on from Tab. 1. This lower bound on is in good agreement with our numerical results.

In the case of IO the ratios in Eq. (31) hold only approximately. Because of the relations in Eq. (10), cancellations between the various terms in Eq. (II) become possible, leading to the correlation visible in the right panel of Fig. 7. We have checked numerically that for IO the lower bound corresponding to Eq. (32) is one order of magnitude weaker. Numerically we obtain a lower bound of for scalar masses of 3 TeV, which can be understood by using Eq. (10) and the bound on from CKM unitarity shown in Tab. 1.

Finally, the contribution to the neutrino masses is strongly suppressed by the electron mass, and thus is allowed to be relatively sizeable compared to other Yukawa couplings. In the case that lies close to its lower bound, the dominating decay channel could be instead of . The most interesting channels are then the and channels. Note however, that this channel is not observable at the like-sign collider due to the Standard Model Møller scattering background. Hence this signature can only be explored if decays of an on-shell doubly charged scalar are observed.

vi.3 Further different signatures

Apart from exploring the flavour structure of Yukawas there are several different signatures to distinguish the two models. In particular, for the triplet model there is large literature on additional observables. Below we give a brief review of a few possibilities.

Starting with collider signatures, we note that observing a doubly charged resonance below the lower bounds in the Zee–Babu model (Eqs. (24) or (25)) would favour the Higgs triplet. In such a case exploring the flavour structure of the decays (as illustrated in Fig. 7) may be used to establish the relation to neutrino mass, see for instance Garayoa and Schwetz (2008). If a resonance is observed consistent with the Zee–Babu bounds, one may also look for the singly charged scalar, which is predicted in both models. In the triplet model the mass difference between the singly and doubly charged scalars is given by the VEV of the Higgs doublet, , times a dimensionless coupling in the scalar potential. Therefore, generically one expects a mass difference . In the Zee–Babu model the two scalar masses are unrelated. Moreover, the triplet couples to , which allows processes like for instance . Signatures of the singly charged triplet component have been investigated in Fileviez Perez et al. (2008); Akeroyd and Chiang (2009); Aoki et al. (2012); Akeroyd and Sugiyama (2011).

Another difference of the models is that the doubly charged scalar in the Zee–Babu model (triplet model) couples to right-handed (left-handed) leptons. In the case of LFV processes at a like-sign linear collider one can use the polarization of the beams to find out the chirality structure of the effective operator induced by the heavy doubly charged scalar Cuypers and Raidal (1997). The possibility to determine the chirality at a hadron collider by using tau decays has been investigated in Sugiyama et al. (2012).

Apart from collider experiments, measurement sensitive to the absolute neutrino mass scale will be important, see Weinheimer and Zuber (2013) for a recent review. Since the Zee–Babu model predicts the lightest neutrino mass to be zero, it can be ruled out by establishing a non-zero lightest neutrino mass for instance in neutrinoless double beta decay experiments, kinematical neutrino mass measurement, and/or in cosmology, eventually combined with a determination of the neutrino mass ordering from oscillation experiments Blennow et al. (2013).

Vii Conclusions

We have studied the current experimental constraints on the Zee–Babu model, taking into account recent data on lepton mixing angles and the MEG limit on . By performing a numerical parameter scan of the model we find that most likely the charged scalars of the Zee–Babu model will be out of reach for the Large Hadron Collider (LHC), including the 14 TeV configuration. If a signal should indeed be seen at LHC this would push the model into a fine tuned parameter region close to the limit of perturbativity. In such a case a signal in upcoming experiments searching for charged lepton flavour violation, such as conversion on nuclei or is guaranteed.

Even if the doubly charged scalar of the model is too heavy to be produced at a collider we point out that a sub-TeV linear collider operated in the like-sign mode may reveal lepton flavour violating processes due to contact interactions induced by the heavy doubly charged scalar. Assuming luminosities of several 10 fb such processes might be observable for scalar masses up to 10 TeV. We stress however, that no signal can be guaranteed, since it is proportional to the Yukawa coupling which is essentially unconstrained by neutrino data.

Furthermore we have considered the same signature for an alternative model for neutrino mass, the Higgs triplet model, which has a similar particle spectrum as the Zee–Babu model, although the mechanism of neutrino mass generation is very different. We have shown that those two models lead to a very different flavour structure of LFV signatures at a like-sign collider (or equivalently to ratios of branching fractions of doubly charged scalar decays, in case they are kinematically accessible). We have outlined various characteristic signatures of the two models. If neutrino mass should indeed be generated by one of those two extensions of the scalar sector it seems likely that the correct model can be identified by using an interplay of various collider signatures as well as absolute neutrino mass measurement.


Note: During the final stage of this work we became aware of Ref. Herrero-Garcia et al. (2014) where also an updated parameter scan in the Zee–Babu model is performed. Taking into account the slightly different method to derive the allowed parameter region and different perturbativity and fine-tuning requirements our results are consistent.

Acknowledgment

We thank the authors of Herrero-Garcia et al. (2014) for sharing their preliminary results with us. One of the authors (H.Z.) is indebted to Werner Rodejohann for useful discussions. This work was supported by the International Max Planck Research School for Precision Tests of Fundamental Symmetries (D.S.) and the Max Planck Society in the project MANITOP (H.Z.). T.S. acknowledges partial support from the European Union FP7 ITN INVISIBLES (Marie Curie Actions, PITN-GA-2011-289442).

References

  • Farzan et al. (2013) Y. Farzan, S. Pascoli, and M. A. Schmidt, JHEP JHEP03, 107 (2013), eprint 1208.2732.
  • Angel et al. (2013) P. W. Angel, N. L. Rodd, and R. R. Volkas, Phys. Rev. D87, 073007 (2013), eprint 1212.6111.
  • Law and McDonald (2013) S. S. C. Law and K. L. McDonald (2013), eprint 1303.6384.
  • Konetschny and Kummer (1977) W. Konetschny and W. Kummer, Phys. Lett. B70, 433 (1977).
  • Cheng and Li (1980) T. P. Cheng and L.-F. Li, Phys. Rev. D22, 2860 (1980).
  • Zee (1985) A. Zee, Phys. Lett. B161, 141 (1985).
  • Zee (1986) A. Zee, Nucl. Phys. B264, 99 (1986).
  • Babu (1988) K. S. Babu, Phys. Lett. B203, 132 (1988).
  • Lindner et al. (2011) M. Lindner, D. Schmidt, and T. Schwetz, Phys. Lett. B705, 324 (2011), eprint 1105.4626.
  • Baek et al. (2012) S. Baek, P. Ko, and E. Senaha (2012), eprint 1209.1685.
  • Babu and Macesanu (2003) K. S. Babu and C. Macesanu, Phys. Rev. D67, 073010 (2003), eprint hep-ph/0212058.
  • Aristizabal Sierra and Hirsch (2006) D. Aristizabal Sierra and M. Hirsch, JHEP 12, 052 (2006), eprint hep-ph/0609307.
  • Nebot et al. (2008) M. Nebot, J. F. Oliver, D. Palao, and A. Santamaria, Phys. Rev. D77, 093013 (2008), eprint arXiv:0711.0483.
  • Ohlsson et al. (2009) T. Ohlsson, T. Schwetz, and H. Zhang, Phys. Lett. B681, 269 (2009), eprint 0909.0455.
  • Abe et al. (2012) Y. Abe et al. (DOUBLE-CHOOZ Collaboration), Phys. Rev. Lett. 108, 131801 (2012), eprint 1112.6353.
  • An et al. (2012) F. An et al. (DAYA-BAY Collaboration), Phys. Rev. Lett. 108, 171803 (2012), eprint 1203.1669.
  • Ahn et al. (2012) J. Ahn et al. (RENO collaboration), Phys. Rev. Lett. 108, 191802 (2012), eprint 1204.0626.
  • Gonzalez-Garcia et al. (2012) M. Gonzalez-Garcia, M. Maltoni, J. Salvado, and T. Schwetz, JHEP 1212, 123 (2012), www.nu-fit.org, eprint 1209.3023.
  • Adam et al. (2013) J. Adam et al. (MEG Collaboration) (2013), eprint 1303.0754.
  • Schechter and Valle (1980) J. Schechter and J. W. F. Valle, Phys. Rev. D22, 2227 (1980).
  • Lazarides et al. (1981) G. Lazarides, Q. Shafi, and C. Wetterich, Nucl. Phys. B181, 287 (1981).
  • Mohapatra and Senjanović (1981) R. N. Mohapatra and G. Senjanović, Phys. Rev. D23, 165 (1981).
  • Chang et al. (1988) D. Chang, W.-Y. Keung, and P. Pal, Phys. Rev. Lett. 61, 2420 (1988).
  • McDonald and McKellar (2003) K. L. McDonald and B. H. J. McKellar (2003), eprint hep-ph/0309270.
  • Kitano et al. (2002) R. Kitano, M. Koike, and Y. Okada, Phys. Rev. D66, 096002 (2002), eprint hep-ph/0203110.
  • Dinh et al. (2012) D. Dinh, A. Ibarra, E. Molinaro, and S. Petcov, JHEP 1208, 125 (2012), eprint 1205.4671.
  • Ma et al. (2001) E. Ma, M. Raidal, and U. Sarkar, Nucl. Phys. B615, 313 (2001), eprint hep-ph/0012101.
  • Beringer et al. (2012) J. Beringer et al. (Particle Data Group), Phys. Rev. D86, 010001 (2012).
  • Bennett et al. (2006) G. Bennett et al. (Muon G-2 Collaboration), Phys. Rev. D73, 072003 (2006), eprint hep-ex/0602035.
  • Bellgardt et al. (1988) U. Bellgardt et al. (SINDRUM Collaboration), Nucl. Phys. B299, 1 (1988).
  • Hayasaka et al. (2010) K. Hayasaka, K. Inami, Y. Miyazaki, K. Arinstein, V. Aulchenko, et al. (Belle Collaboration), Phys. Lett. B687, 139 (2010), eprint 1001.3221.
  • Aubert et al. (2010) B. Aubert et al. (BaBar Collaboration), Phys. Rev. Lett. 104, 021802 (2010), eprint 0908.2381.
  • Bertl et al. (2006) W. H. Bertl et al. (SINDRUM II Collaboration), Eur. Phys. J. C47, 337 (2006).
  • Abrams et al. (2012) R. Abrams et al. (Mu2e Collaboration) (2012), eprint 1211.7019.
  • Knoepfel et al. (2013) K. Knoepfel et al. (R. C. Group) (2013), eprint 1307.1168.
  • Cui et al. (2009) Y. Cui et al. (COMET Collaboration) (2009), eprint KEK-2009-10.
  • Witte et al. (2012) H. Witte, B. Muratori, K. Hock, R. Appleby, H. Owen, et al., Conf. Proc. C1205201, 79 (2012).
  • del Aguila et al. (2013) F. del Aguila, M. Chala, A. Santamaria, and J. Wudka, Phys. Lett. B725, 310 (2013), eprint 1305.3904.
  • Alloul et al. (2013) A. Alloul, M. Frank, B. Fuks, and M. R. de Traubenberg, Phys. Rev. D88, 075004 (2013), eprint 1307.1711.
  • Aad et al. (2012) G. Aad et al. (ATLAS Collaboration), Eur. Phys. J. C72, 2244 (2012), eprint 1210.5070.
  • Chatrchyan et al. (2012) S. Chatrchyan et al. (CMS Collaboration), Eur. Phys. J. C72, 2189 (2012), eprint 1207.2666.
  • Rizzo (1982a) T. G. Rizzo, Phys. Rev. D25, 1355 (1982a).
  • Rizzo (1982b) T. G. Rizzo, Phys. Lett. B116, 23 (1982b).
  • London et al. (1987) D. London, G. Belanger, and J. Ng, Phys. Lett. B188, 155 (1987).
  • Cuypers and Raidal (1997) F. Cuypers and M. Raidal, Nucl. Phys. B501, 3 (1997), eprint hep-ph/9704224.
  • Raidal (1998) M. Raidal, Phys. Rev. D57, 2013 (1998), eprint hep-ph/9706279.
  • Cakir (2006) O. Cakir, New J. Phys. 8, 145 (2006), eprint hep-ph/0604183.
  • Rodejohann (2010) W. Rodejohann, Phys. Rev. D81, 114001 (2010), eprint 1005.2854.
  • Rodejohann and Zhang (2011) W. Rodejohann and H. Zhang, Phys. Rev. D83, 073005 (2011), eprint 1011.3606.
  • Garayoa and Schwetz (2008) J. Garayoa and T. Schwetz, JHEP 0803, 009 (2008), eprint 0712.1453.
  • Fileviez Perez et al. (2008) P. Fileviez Perez, T. Han, G.-y. Huang, T. Li, and K. Wang, Phys. Rev. D78, 015018 (2008), eprint 0805.3536.
  • Akeroyd and Chiang (2009) A. Akeroyd and C.-W. Chiang, Phys. Rev. D80, 113010 (2009), eprint 0909.4419.
  • Aoki et al. (2012) M. Aoki, S. Kanemura, and K. Yagyu, Phys. Rev. D85, 055007 (2012), eprint 1110.4625.
  • Akeroyd and Sugiyama (2011) A. Akeroyd and H. Sugiyama, Phys. Rev. D84, 035010 (2011), eprint 1105.2209.
  • Sugiyama et al. (2012) H. Sugiyama, K. Tsumura, and H. Yokoya, Phys. Lett. B717, 229 (2012), eprint 1207.0179.
  • Weinheimer and Zuber (2013) C. Weinheimer and K. Zuber, Annalen der Physik, 525, 565 (2013), eprint 1307.3518.
  • Blennow et al. (2013) M. Blennow, P. Coloma, P. Huber, and T. Schwetz (2013), eprint 1311.1822.
  • Herrero-Garcia et al. (2014) J. Herrero-Garcia, M. Nebot, N. Rius, and A. Santamaria (2014), to appear.
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