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Thus, for example, curves such as those shown in Fig. 2 show that the Néel model state (φ = 0) gives the minimum gs energy for all values of κ < κ c 1 where κ c 1 is also dependent on the level of LSUBn approximation, as we also see below in Fig. 3. By contrast, for κ > κ c 1 the minimum in the energy is found to occur at a value φ = 0. If we consider the pitch E/N 2φ/π J 2 '=0.0 J 2 '=0.25 J 2 '=0.5 J 2 '=0.75 J 2 '=0.8 J 2 '=1.0 J 2 '=1.2 J 2 '=1.4 FIG. 2: (Color online) Ground-state energy per spin of the spin-1/2 J 1 -J ′ 2 Hamiltonian of Eq. | cond-mat.str-el |
Naturally, the proper approach is to take into account both of these effects (under which the work is in progress) but in what follows we merely concentrate on the problem of two coupled chains connected by this interladder coupling, see Fig. 1. This will enable us to study the generic role of the interladder coupling. While the detailed derivation of the interladder coupling is somewhat lengthy and will be described elsewhere, a simple cartoon-picture shown in the left panel of Fig. 2 explains its basic idea. Here we see that the two Zhang- Rice singlets [9] which are situated next to each other in the two neighboring chains (see dotted rings in Fig. | cond-mat.str-el |
For U = 0 the ground state of ( 4) is of course the Fermi sea, |ψ 0 = k<kF c † k↑ c † k↓ |0 ,(5) with particle density n = 1 + k F /π. For U = ∞, on the other hand, the ground state is [22,34] |ψ ∞ = i (1 -n i↑ n i↓ ) |ψ 0 ,(6) i.e., the Fermi sea with all doubly occupied sites projected out. At half-filling a Mott-Hubbard metal-insulator transition occurs at interaction strength U c = W , with the Mott gap given by ∆ = U -U c for U ≥ U c [22]. This metal-insulator transition is also captured by correlated variational wave functions [34,35]. Interaction quenches We now consider the following nonequilibrium situation. For times t ≤ 0 the system is prepared in the ground state for interaction parameter U = 0 or ∞, i.e., |Ψ(0) = |ψ 0 ψ 0 |ψ 0 metallic state, or(7a) |Ψ(0) = |ψ ∞ ψ ∞ |ψ ∞ Mott insulator. | cond-mat.str-el |
The phase fluctuations are small, δφ 2 ∼ 1/g 0 , so we keep only quadratic terms to the action (44). The resulting Gaussian path integral over φ ± can be readily done. This procedure is equivalent to the summation of all one-loop diagrams of the conventional perturbation theory, i.e. to the "random-phase approximation" (RPA). We restrict ourselves to the most interesting low voltage/temperature limit, max{eV, kT } ≪ g 0 E C . In this limit, we evaluate the interaction correction to the CGF with the logarithmic accuracy. | cond-mat.mes-hall |
Plotting the square of the DW length versus the dot length reveals a linear variation. The cross-over is therefore Landau-like, i.e. of second order. Such transitions are associated with a breaking of symmetry, which in the present case is whether the top surface vortex shifts towards +x or -x. The results are qualitatively similar for other thicknesses. FIG. | cond-mat.mtrl-sci |
2(c) and 3(b) for the clockwise and counterclockwise QH edge states, respectively. By contrast, if an EC is grounded in the opposite side of the sample in the circulation path, the current is not drained to ground before reaching the drain D, resulting in the recovery of the conductance as illustrated in Figs. 2(b) and 3(c). Up to this point, we assume that conducting carriers are always emitted from the source S, i.e., a positive (negative) V SD for hole (electron) carriers. But, if hole (electron) carriers are emitted from both D and the grounded EC for a negative (positive) V SD , the conductance reduction occurs as the same amount of holes (electrons) enter into and leave from D simultaneously with vanishing current I. Thus, the results in Figs. | cond-mat.mes-hall |
In the adiabatic limit (ω 0 /t 0 ) ≪ 1, the motion of the particle is affected by quasi-static lattice deformations, whereas in the opposite, anti-adiabatic limit (ω 0 /t 0 ) ≫ 1 the lattice deformation is presumed to adjust instantaneously to the position of the carrier. The dimensionless EP coupling constant g 2 normally appears in (small polaron) strong-coupling perturbation theory, where it describes the polaronic mass enhancement m * /m = e g 2 (for homogeneous systems, g i = g). There is another natural measure of the strength of the electron-phonon interaction, the familiar polaronic level shift E p . At strong EP coupling, E p gives the leadingorder energy shift of the band dispersion. 18 In general, there is no simple relation between g 2 and E p . If the EP coupling is local and the phonon mode is dispersionless, however, then g 2 = E p /ω 0 , and E p is usually identified with the polaron binding energy. | cond-mat.str-el |
The contribution of weak localization is reduced at high magnetic field and the remaining correction to σ is usually attributed to Coulomb interaction. The correction to the conductivity due to Coulomb interaction in a two-dimensional metal reads [21,31] δσ ee = - e 2 2π 2 h • g 2D • ln h k B T τ e(4) where g 2D is a constant of the order of unity. Figure 7 shows that experiment agrees well with the theory using g 2D = 0.8. Unlike weak localization, which originates from a change in the diffusion constant [21], this Coulomb interaction effect is a correction to the density of states. These results indicate that the correction to the conductivity due to Coulomb interaction cannot be neglected in the transition between weak localization and strong localization in graphene, as it is treated in existing theoretical works [2][3][4][5][6][7][32][33][34]. CONCLUSION We have reported on the crossover from weak localization to strong localization in disordered graphene, the disorder being created with ozone. | cond-mat.mes-hall |
5(a)) easily. Acknowledgments We are grateful to R. Shindou, L. Balents, and X.-L. Qi for fruitful discussions. This research is supported in part by Grant-in-Aids from the Ministry of Education, Culture, Sports, Science and Technology of Japan. APPENDIX A: GENERAL DESCRIPTION OF THE GAP-CLOSING POINTS IN m-k SPACE In this Appendix we assume that the bands are nondegenerate almost everywhere in k space. It allows existence of isolated k points with band degeneracy; meanwhile, the degeneracy in an extended region in k space, such as Kramers degeneracy in systems with I-and time-reversal symmetry is excluded. The vectors A α (k), B α (k) are defined in the three-dimensional k space. | cond-mat.mes-hall |
The behaviour of the charge current is shown in Fig. 5 for k = 0 (strictly two-dimensional system) and the aspect ratio N sm /N = 10/20. The orientation of each arrow in Fig. 5 represents the direction of the current flow and the length of the arrow gives the magnitude of the local charge current flowing between neighboring atoms. Figure 5 demonstrates that there is strong penetration of transport electrons into the switching magnet, and it is the spin precession of these electrons that results in a spin-transfer torque (spin current absorption) which is as large as in the CPP geometry. Finally, we need to explain why the decay of the CPIP spin current in a half-metallic ferromagnet is slower than in the CPP geometry. | cond-mat.mtrl-sci |
Expressed differently, the system cannot choose a unique wave vector of an ordered structure which minimizes the free energy. Instead, it is frustrated between different structures with different critical wave vectors Q c 's and equally low free energy. For instance, low energy spin fluctuations are expected to be present in a very large region of momentum space which is the signature of frustrated itinerant magnetism. This is in contrast to non-frustrated systems where the fluctuations are confined to the immediate vicinity of a unique incipient ordering vector. This scenario for frustrated itinerant magnetism was recently investigated in detail for LiV 2 O 4 by present authors. 4 . | cond-mat.str-el |
For the fourth-order enhancement factor, Eq. ( 45), we again note that s and q are non-singular near the nucleus, while lim r→0 s 2 (r) = Z 2 /[3π 2 n(0)] 2/3 . (53) With a density of the form of Eq. ( 26), Eq. ( 48) thus gives the near-nucleus behavior of the fourth-order potential as v (4) θ (r) = c 0 16[9π 4 n(r)] 2/3 - 16Z 2 3r 2 (5b 2 + 3c 21 ) + 32Z 3 9r (18a 4 + 17b 2 + 18c 21 ) + nonsingular terms. (54) The singularities in 1/r 2 and 1/r can be removed by requiring that the numerators of the first two terms of Eq. | cond-mat.mtrl-sci |
Since η is typically smaller than η so (i.e. 2αk F > η) note that the static value of σ S,z xy is always negative. It is known however that the dc limit is problematic within the Kubo formula when using the conventional spin-current operator, namely because it leads to the incorrect physics for ε so ≫ η as a result of neglecting the contribution of the vertex corrections. This is not necessarily the case for finite frequencies and relatively low impurity densities (as discussed below), which is in fact the regime that we are mostly interested in here. In the opposite limit, ε so ≪ η, Kubo formula gives a result which coincides qualitatively with the expected result [18]. We first consider this case. | cond-mat.mes-hall |
( 29) are (i) commutators of an elementary diagram with itself at the same place or (ii) commutators of two elementary diagrams acting on disconnected places. Obviously such contributions vanish identically. As a direct consequence, only order p = 0 of Eq. ( 29) contribute to the amplitude of elementary diagrams in Ĥeff . In other words, for any elementary processĤeff = µ=0 β=1 ∂ µ ln Ẑ . (39) It is then straightforward, using Eq. | cond-mat.str-el |
4 (a). The arrows denote the changes caused by the laser irradiation. Two features are worth noting; (1) the laser-irradiation induces clear IMT only below a critical temperature (around 75K) and ( 2) the resulting metallic state is identical to that reached in the warming run at the respective temperatures [see also the insets of Fig. 4]. This transition is persistent as long as the temperature is held constant and can be reversed by heating above 120K as previously reported [10]. In contrast, the laser illumination effect for y = 0.40 has not been observed in all temperature range, including the insulating state. | cond-mat.str-el |
transport of the system. The differential conductance G is plotted as a function of the perpendicular magnetic field B and gate voltage V G . Six Coulomb peaks are visible. The positions of these peaks reflect the chemical potentials of ground-state transitions. These positions can be roughly understood using the CI model. According to this model, the potentials include a fixed Coulomb repulsion energy due to the electrostatic interaction of the electronic charge and a term due to the single particle excitation spectrum. | cond-mat.mes-hall |
The localized states along the zigzag edges are formed by unpaired electrons belonging to the natural non-bonded orbitals, which participate in formation of HOMO and LUMO orbitals. The α-and β-spin states of HOMO and LUMO orbitals are spatially separated, i.e. localized at opposite zigzag edges. The (HOMO-1) and (LUMO+1) orbitals usually correspond to the surface states, redistributed over the entire graphene structure. The surface states are important for conductivity of graphene in a transverse electric field, because the charge transfer between the spatially separated HOMO and LUMO orbitals may occurs through participation of the surface states. The electron density distribution for the edge states and the surface states is presented in Fig. | cond-mat.mtrl-sci |
Apparently, the variance Eq. ( 26) is affected by the wiggles at the top of the curves T Q(z -z 0 ) much stronger for z 0 = ±0.5 than for z 0 = ±1, which makes the evaluation of χ for small z 0 ineffective. VI. CONCLUSION It is interesting to point out that, while the classical limit of the Chalker-Coddington model based on potential Eq. ( 2) reduces to the bond percolation, similar form of potential Eq. ( 8) leads to the site percolation. | cond-mat.mes-hall |
However, this system appears to have strong structural disorder since V 5+ and Cu 2+ do not order perfectly in this fashion. Ingredients for other such honeycomb spin systems could be, for instance, Cu 2+ ions within CuO 5 units arranged on the honeycomb lattice. In such a trigonal bipyramidal crystal field environment of the oxygens, the copper ion would then have a single hole, which is located in the d 3z 2 -r 2 orbital. If the resulting S = 1/2 moments have significant next neighbor interactions, they might also be candidates to explore the physics discussed here. Our study is also relevant to bilayer triangular antiferromagnets where the triangular layers have an AB stacking, with antiferromagnetic exchange couplings present between neighboring sites within each layer (J 2 ) as well as between neighboring sites across the two layers (J 1 ). In this case, each layer acts as one sublattice of the honeycomb antiferromagnet. | cond-mat.str-el |
Since k ≥ 2, heat flows in the opposite direction of charge. As in the case of the MR state, the electron operator is no longer unique. 34,39 There are 2k + 1 electron operators with left-and right-scaling dimensions ∆ R = (k + 2)/4, ∆ L = k/4, so that their total scaling dimension is (k + 1)/2 and their conformal spin is ∆ R -∆ L = 1/2, as required for a fermion. These operators are given by 39 Ψ † e,RR m = χ m j=k/2 e i k+2 2 φ ,(16) with m = -k/2, -k/2 + 1, ..., k/2. The χ m j=k/2 are related to the spin j = k/2 primary fields of SU (2) k and are built entirely from fields in the neutral sector of the edge theory. The χ m j=k/2 have scaling dimension k/4 and can be constructed by operating with the SU(2) k current operator J + ∼ ψ 1 e iφσ multiple times on a "bare outer edge" electron operator. | cond-mat.str-el |
One possibility to account for this drop in K(T ) is the presence of DM interaction arising from the fact that in K 2 CuP 2 O 7 there is no center of inversion symmetryrelating two neighboring copper atoms along the chain. In the presence of DM interaction, application of a magnetic field parallel to the DM vector opens a gap ∆ in the magnetic excitation spectra. Since in a first approximation ∆ increases with B to the power 2/3, it shall not be visible in the B = 0 specific heat data or the low field susceptibility shown in Fig. 2. However, further experiments are needed to confirm or discard this explanation. In NMR valuable information on the dynamic of lowenergetic spin excitations can be gained from the analysis of the temperature dependence of nuclear spin-lattice relaxation rates. | cond-mat.str-el |
C. Reliability of exchange calculations The HL and HM methods used in this paper generally underestimate the electron correlations in the double dot, and results obtained using these approaches are more appropriately thought of as order-of-magnitude estimates. However, these methods do capture the basic physics of exchange coupling in DQD systems and calculations performed under these approximations can serve as guidelines for initial development of silicon-based DQD structures. In this section we discuss the expected regime of applicability of these methods. We first clarify the relationship between the valley splitting ∆ and the other energy scales in the problem as we work within a single-valley approximation. To have reliable control over the initialization process, when the electrons tunnel into the quantum dot from an outside reservoir, it is necessary for ∆ to exceed the thermal broadening of the reservoir Fermi level: ∆ ≫ k B T . If Coulomb interaction can couple different valleys, it could pose additional requirements on ∆. | cond-mat.mes-hall |
The specific-heat measurements were performed using a continuous relaxation-time technique. 11,12 For the X-band ESR measurements a Bruker ELEXSYS E500 spectrometer operating at a frequency of 9.4 GHz was used. The ESR experiments in the frequency range 36 -250 GHz were performed at temperatures down to 1.3 K using a tunable-frequency home-made ESR spectrometer (similar to that described in Ref. 13) equipped with a 16 T high-homogeneity superconducting magnet. In our experiments we used single crystals with a typical size of 4x3x1 mm 3 , synthesized by aqueous reaction of stoichiometric amounts of ammonium bifluoride, pyrazine, and copper(II) hexafluorophosphate hydrate (using the same procedure as reported in Ref. 9,14) III. | cond-mat.str-el |
This is in fact the same ground state as for the regular two-leg ladder with J rung ≫ J leg in the limit J rung → ∞. It also can be characterized by a string order parameter, e.g., 33 O string = lim n→∞ j+n k=j -4σ z j S z j = lim n→∞ exp iπ j+n k=j σ z k + S z k . (5) The excitation spectrum in this limit again has a gap. III. WEAK KONDO COUPLING The question we want to address is what happens for weak Kondo couplings |J ′ | ≪ J. The Kondo necklace model can be viewed as a particular limit of an asymmetric two-leg ladder model, in which the coupling along the first leg J is much larger than the rung coupling J ′ , which in turn is large compared to the exchange J 2 along the second legJ ≫ J ′ ≫ J 2 . | cond-mat.str-el |
We note that for the specific system the average filling fraction is somewhat different from ν = 2/3, due to the higher density near the edge. If the hole-droplet picture is correct, one can possibly (but not necessarily for energetic reasons) find that the global ground state of the semi-realistic model has a total angular momentum M = 280, if 20 electrons are distributed in N orb ≥ N I = 26 orbitals. We study the global ground state of a system of 20 electrons with various d and N orb (which serves as an additional hard edge confinement). We plot the results in Fig. 2. When the distance of the background charge d increases, the confining potential becomes weaker and the total angular momentum of the global ground state increases and goes through steps at M = 270, 280, 288, and 294. | cond-mat.mes-hall |
(6.5) where C x , C y , K 0 , K 1 , and K 2 are the stiffness in Eq. (6.4). When ∆ S = 0, at the tree level and in the longwavelength regime, both the gauge-like "coupling constant" Q S and K 0,1,2 scale to infinite, but the ratio Q 2 S / K 0,1,2 scales to 0 as a function of b -1 . This implies that the gauge-like coupling is irrelevant. Quantum fluctuations may change the tree-level scaling behavior as we include loop corrections. However, for large enough K 0,1,2 or small enough Q S , the irrelevancy of the gaugelike coupling will not be changed. | cond-mat.str-el |
Even though the absolute intensity of the D-band decreases slightly, the ratio of the D-to G-bands, a commonly used qualitative measure of defect density, 53,54 increases linearly with time as seen in Fig. 6b. This increase is evidence for the disruption of the sidewall due to oxidation. At 7 minutes, the intensity of the RBMs is abruptly extinguished, and the shape of the G-band dramatically changes character. The lower energy modes of the Gband abruptly disappear, and the lineshape acquires an asymmetric Breit-Wigner-Fano profile characteristic of nanotube ensembles and amorphous carbon. 51 At this point, based on the extinction of the breathing modes, we conclude that the nanotubes in resonance with the laser have been entirely destroyed by the ozone. | cond-mat.mtrl-sci |
In the vertical case, the dot is connected to the leads with a relatively large surface, therefore there are many different conduction channels coupling to a given dot state. For each dot state i and lead α one can construct a simple linear combination of the modes, c † ξ,iασ , that hybridizes with d iσ , and assumed to be independent for i = ±. The Hamiltonians are (cf. Eq. ( 5)-( 6) of paper I), H vert cond = ξ,i,α,σ ξ α c † ξiασ c ξiασ ,(4) H vert hyb = α,i,ξ,σ t αi (c † ξiασ d iσ + h.c.) . (5) Eqs. | cond-mat.str-el |
The observation of the clear spectral structure shows that our sample has a K iso = 3 % K aniso = -0.75 % Q = 2.9 MHz 75 As Spin-Echo Intensity ( arb. unit ) 7.0 6.5 6.0 5.5 H ( T ) K iso = -1.9 % K aniso = -0.33 % Q = 1.48 MHz 139 La (a)(b) Exp.Exp. Sim.Sim. 122 K 100 K high quality and crystallizes well in each polycrystalline grain. The spectra of both 75 As and 139 La nuclei show large shifts at low temperatures, and the directions of the shifts are opposite, as observed in Fig. 2. | cond-mat.str-el |
(31) The general trend of the data in Fig. 2 is consistent with there being an analogous reduction of the bosonic bath spectral density that requires the replacement of K 0 byK0 = K 0 /A Λ,s(32) when extrapolating NRG results to the continuum limit Λ = 1. However, we have not obtained a closed-form expression for A Λ,s . Table I lists values λ c0 (Λ → 1) extrapolated from the data plotted in Fig. 2. For s ≥ 0.4, these values are in good agreement with Eq. | cond-mat.str-el |
This holds for Au and Ag in Figs. 7(a) and (b). If W c ≤ ∆-0 , we observe pinning. For a close-packed monolayer the plane capacitor model explains the pinning (Eq. ( 18)) observed for Mg and Ca in Fig. 7(b). | cond-mat.mtrl-sci |
3(a). In the case of thermal transport, the influence of temperature on I 0 th is much more drastic: the maximum in I 0 th seen at T = 0.25J is quickly washed out at temperatures T > h c1 . By contrast, for spin transport and in I 0 s , the interesting features are hidden at all temperatures under the large field-induced decrease of the regular part as ω approaches the gap [Fig. 4(c)], whereas in the case of thermal transport, the low-frequency behavior dominates the field dependence of I 0 th at T < h c1 [Fig. 4(d)]. For T > 0.25J, the finite-size effects of the integrated weights become very small, which is shown in the insets of Fig. | cond-mat.str-el |
Although the phonon-mediated ee coupling (whose magnitude is given by g 2 ee ∼ 10 -2 ) is in general weaker than the Coulomb coupling (given by the interaction parameter [5] r s ∼ 0.7), e-ph interaction actually contributes more significantly to the effective velocity renormalization than Coulomb interaction, because the real part of the self-energy due to e-ph in- teraction exhibit sharp changes near the phonon energies which are not present in the case of Coulomb interaction. This results in a larger value of the energy derivative of Re Σ near the phonon energy and therefore larger value of v * /v. In Eq. ( 8), logarithmic singularities occur in Re Σk (ω) at ±ω 0 where Im Σk (ω) goes through a finite step jump, yielding also logarithmic singularities in the derivatives of Re Σk (ω) with respect to ω and k. Fig. 3 shows the calculated renormalized energy spectrum for electron densities n = 10 13 cm -2 , the sharp kink shows the logarithmic singularity at ω = ω 0 . Such a kink only occurs in the conduction band if the phonon energy is within the Fermi sea ω 0 < ε F ; if the phonon energy lies outside of the Fermi sea ω 0 > ε F the conduction band will be smooth and the logarithmic singularity could occur in the valence band. | cond-mat.mtrl-sci |
21 In their paper, this narrowing effect was indirectly observed by the decreasing of the oscillation period of the A-B ring, while in this paper such effect is directly shown by the spatial distributions of the local currents. When the Anderson disorder strength is getting larger, the local currents spread to the bulk and broaden the edge channels again(see as fig. 3(c)). However, only when the disorder strength exceeds the critical value W c , the spread local-current flow can reach the lower edge channels with different chirality, the effective backscattering (as shown in the local current flow vector located near the lower edge in the region 0 < X < 40a in fig . 3(d)) can take place, leading to the reduction of the conductance between the two terminals. These pictures explain why the traditional quantized plateau is robust under weak disorder and how it is destroyed in the strong disorder limit. | cond-mat.mes-hall |
23 However, the surfaces of real cantilevers (and in more general material systems; even for the most thoroughly polished surfaces) have random surface roughness on different lateral length scales. Recently, it was shown that the response of the curvature of cantilevers to changes in their surface stress depends significantly on the surface morphology. 23 This dependence was attributed to the transverse coupling between the out-of-plane and in-plane components of the surface-induced stress. Moreover, roughness corrections were introduced, which are highly important for experiments measuring the surface stress on nominally planar surfaces. 23 However, calculations of the roughness effects on the cantilever sensitivity as a function of characteristic parameters of random rough surfaces are still missing. This will be the topic of the present paper. | cond-mat.mes-hall |
3c compares the full 2DM model to the measured broad peaks with the background capacitance subtracted. The measured data are shifted horizontally to align peak C to the tallest peak predicted by the model. This is consistent with peak C lying near zero effective Rydbergs, the energy above which the electrons are unbound. No other free parameters were employed. We see that the model generally agrees well with the measurement, although some features in the data are not accounted for, such as the small peak near -7 Ry* and the relative sharpness of peak A. As indicated by the arrows, the model predicts that the peaks due to the third and four electrons ( 3,4 ) will be unresolved. | cond-mat.mes-hall |
This variation is independent of the sample morphology. Instead, it depends on the synthesis process or even on different runs with the same process. For a pure Cu 2 O with an ideal lattice structure, the intrinsic magnetic property is diamagnetic since the d shell of Cu +1 is full, i.e., the electron configuration of Cu +1 is 3d 10 , and neither Cu +1 nor O 2-is a magnetic ion. In this sense, we believe that the susceptibility of the diamagnetic octahedron, χ OH = -0.95×10 -6 emu.g -1 .Oe -1 , is close to its intrinsic value since the contribution from the magnetic impurities is smaller by one orders of magnitude as revealed by the impurity analysis later, see table 1. χ OH =-9.5×10 -6 Nanosphere 148. 30 a Value calculated using the bulk value of magnetic moments for the impurities b Value calculated using the atomic value. Magnetic measurements and analysis The ZFC and FC M(T) measurements of the PM nanospheres are shown in figure 3(a). | cond-mat.mtrl-sci |
On the basis of both first-principles calculations [19] and experiments [20] it has been suggested that if an Ag-Cu alloy, rather than pure Ag, is used as a catalyst, the selectivity toward ethylene oxide will be improved. In several recent works the mechanism behind the enhanced selectivity due to alloying with copper has been addressed. [19,21,22] The composition and structure of the catalyst surface, though, has not been thoroughly investigated. Given the discussion presented above, it is clear that the mechanism of this catalytic reaction can be strongly influenced by the atomistic details of the surface. In particular, given the conditions of operation of the Ag-Cu catalyst in realistic applications, the formation of surface oxides cannot be ruled out. It has been established that silver can form several surface oxide structures [23,24,25] and from the phase diagram of the Cu-O system, considering the temperature and oxygen pressure used in practical applications, but neglecting the effect of ethylene, the formation of CuO is expected, therefore suggesting that copper might also oxidate. | cond-mat.mtrl-sci |
12 @1 =C G 8H M ? M 9 # 4 M 4 9 N # $ ' $" 2331 @1 3C G 2H # N O N 4 # ! M # $ ( % 01 1 @1 C G 1 H # $ % - A ! ( 5 , 6 ' ) ! & 0/ @1 /C G 1 H # $ % I ' ! $ ( & O $E 0 0 2$ G 11H 9 ? | cond-mat.mtrl-sci |
Polycrystalline Ce(Ni 1-x Cu x ) 2 Al 3 compounds with various x were prepared by arc-melting stoichiometric amounts of constituent elements in a high purity argon atmosphere. The purities of the starting elements are 99.9%, 99.997%, 99.999%, and 99.999% for Ce, Ni, Cu and Al, respectively. No annealing treatment was carried out. Power x-ray diffraction spectra confirmed the formation of the hexagonal PrNi 2 Al 3 -type structure of the samples with x < 0.5, with tiny impurity peak appearing at around 2θ∼50 • as reported previously [6,9]. For x > 0.5, a trace of second phase, which was determined to be CeCuAl 3 [10], appears gradually. The second phase in the sample of x = 0.6, i.e., the upper substitution limit in this work, was estimated to be less than 3%. | cond-mat.str-el |
The displacement energy is of the order of the order of ~40 eV for graphite and diamond [65,66]. These values refer to bulk displacement energies while those near the surface can be significantly lower due to lower coordination number (lower binding energy) and lower displacement radius. If the ion energy is lower than the threshold for atomic displacements, it can still induce bondbreaking of pre-existing surface configurations [67], thus inducing chemical reactions which may also involve the incoming particle [68]. The energy threshold for bond breaking in C-C system is of the order of several eV [69], thus, having more probability to occur than the displacement event. In addition, molecular dynamics simulations show that surface processes dominate when the energy of impinging atoms is <30 eV, while subsurface processes govern the structure formation for higher energies [67]. Sputtering-based methods such as IBS or MS can be considered energetic growth processes themselves due to the energy of the emitted particles (mostly neutrals) in the hyperthermal range (from a few eV up to tens of eV) [46] that may activate or enhance surface processes. | cond-mat.mtrl-sci |
For this reason, we instead performed 139 La NMR and 121/123 Sb NQR experiments on LaOs 4 Sb 12 because the characteristic motion of R ions is believed to be similar for the ROs 4 Sb 12 (R= La, Pr, Sm, Nd) family. Because the 139 La nucleus (I=7/2) has an electric quadrupolar moment Q, the nuclear spin-lattice relaxation rate 1/T 1 at the La site can directly probe the unusual dynamics of the La ions through quadrupolar coupling of the La nucleus to the electric field gradient (EFG). In this paper, we report results of 139 La NMR and 121/123 Sb NQR measurements on LaOs 4 Sb 12 . We found that 1/T 1 T at the La site exhibits a different temperature dependence from that at the Sb site. By subtracting the conduction electron contribution from the observed 1/T 1 T at the La site, we obtained additional relaxation at the La site 139 (1/T 1 T ) add . 139 (1/T 1 T ) add is nearly con- stant at high temperatures and decreases rapidly below about 50 K. We show that this temperature dependence can be explained quantitatively by a recent theoretical calculation based on the Raman process of anharmonic phonons due to the rattling motion. | cond-mat.str-el |
Since the groups of -(CH 2 ) n CH 3 only has little effect upon the properties of the polymers [5], these groups have been removed from the backbone of a polymer and are, therefore, excluded in all calculations. Table 1 shows the first singlet and triplet excitation energies of the polymers in gas phase calculated with the adiabatic TDDFT. The experimental results are also listed for comparison. Usually a polymer is of infinite chain length. In practical calculations, we only choose several repeating monomeric units. The number of "molecular" rings included in our calculations for each polymer is given in the parentheses in Tables 1 and3. | cond-mat.mtrl-sci |
In preparation of our main themes -interaction and thermodynamics of spinons -we introduce alternative quantum numbers for fermion momenta and for spinon momenta and spins, along with rules that translate the fermion composition of any XX eigenstate into the corresponding spinon composition (Sec. 2). For the energy of an arbitrary XX eigenstate, the mapping converts its dependence on the fermion quantum numbers into its dependence on the spinon quantum numbers. The resulting expression is akin to a coordinate Bethe ansatz (CBA) for the spinon momenta and spins (Sec. 3). From an entirely different perspective, the XX chain is interpreted as a set of interacting spinon orbitals with internal degrees of freedom exhibiting features akin to electrons in partially filled electronic shells (Sec. | cond-mat.str-el |
(16) Note again that for propagating waves, k x is real; while for evanescent solutions, k x is complex. Similar to the case of the monolayer graphene, by calculating the group velocity v x (k) = dǫ (k) / dk x , we can prove that k 1,3 are right movers (transmission waves) while k 2,4 are left movers (reflective waves). On the other hand, to select appropriate evanescent states, since κ lx (l = 1, 2, 3, 4) are complex numbers, we should consider the asymptotic behavior of these states at ±∞. Finally, the general solutions for bilayer graphene can be expressed as 10 (incident electron coming from k 1 ) Ψ I (r) = ψ + -1 (k 1 ) + r 1 ψ + -1 (k 2 ) + r 2 ψ + -1 (k 4 ) + r 3 ψ + +1 (κ 2 ) + r 4 ψ + +1 (κ 4 ) ,(17a) Ψ III (r) = 4 l=1 f l ψ - -1 (q l ) + g l ψ - +1 (τ l ) ,(17b) Ψ V (r) = t 1 ψ + -1 (k 1 ) + t 2 ψ + -1 (k 3 ) + t 3 ψ + +1 (κ 1 ) + t 4 ψ + +1 (κ 3 ) . (17c) Here q l and τ l are the corresponding wavevectors for propagating states (E - -1 ) and evanescent states (E - +1 ) inside the barrier. Then by solving Eq. | cond-mat.mes-hall |
The two cases differ only in the boundary conditions applied on the top surface. In the first case (Fig. 4(a)), a pressure P 0 is imposed on both the liquid and the solid, corresponding to an average hydrostatic pressure on the system. 2 The fluid flow is free, thus allowing feeding from the upper boundary. We refer to this case as "ideal feeding" because it simulates a sample in contact with a liquid reservoir (a feeder) at pressure P 0 . In the second case, we also impose a pressure P 0 on the upper surface of the solid grains but the fluid flow is set to zero. | cond-mat.mtrl-sci |
Note, that the temperature dependence in Fig. 3 is also similar to that previously reported for non-magnetic semiconductors [6]. We now turn to microscopic mechanisms responsible for photocurrent generation. In case of Drude absorption, photocurrents stem from spin-dependent asymmetry of the optical transitions accompanied by scattering and/or from energy relaxation [6]. Here we focus on the energy relaxation of electron gas heated by THz yielding a polarization independent photocurrent. While the first mechanism depends on the radiation polarization the latter one is polarization independent. | cond-mat.mes-hall |
I. INTRODUCTION Understanding the ground state properties of frustrated low-dimensional spin systems is a central issue in current condensed matter physics. These systems have competing interactions which stabilize different states with distinct symmetries. 1 An interesting phenomenon in spin systems is the formation of a spin-liquid with no long-range magnetic order (LRO) due to suppression by geometric/magnetic frustration. A simple example is the frustrated S = 1/2 square lattice (FSL) model. In this model which is also known as J 1 -J 2 model, the spin Hamiltonian is H = (k B J 1 ) <ij> S i • S j + (k B J 2 ) <ik> S i • S k ,(1) where the first sum is over nearest-neighbor spin pairs and the second is over next-nearest-neighbor spin pairs. | cond-mat.str-el |
However, since such a relaxation is mediated magnetically, it can be expected to be much slower than isospin relaxation. A third way of coupling the +-and -subspaces is not related to relaxation. As can be seen from the master equation, Eq. ( 3), a detuning of the energy levels ∆ǫ = ǫ u -ǫ d gives rise to additional terms. It turns out that for small detuning, ∆ǫ ∼ Γ, this results in precession of the isospin about the axis n = (0, 0, 1),d dt I prec = ∆ǫ n × I. (17) Since three spatial directions (n L , n R and n) appear in the master equation, the symmetry of the fluxdependence about φ = π is lost. | cond-mat.mes-hall |
NUMERICAL RESULTS In the following we present numerical results on the equilibrium spectral function and linear conductance, when the quantum dot is in the Kondo regime. We will distinguish between two different situations; symmetric (ε d = -U/2) and asymmetric (ε d = -U/2) Anderson models. The origin of such a distinction stems from the way in which ferromagnetic leads act on the quantum dot. More specifically, in the asymmetric Anderson model ferromagnetism of the leads gives rise to a spin splitting of the Kondo resonance in the parallel configuration, while in the symmetric model no such a splitting appears (assuming that the dot is coupled with the same strength to the left and right leads). 6,7,10 In other words, an effective exchange field, due to coupling to magnetic leads, acts on the dot in the former case, while such a field vanishes in the latter case. The effective field is directly related to the difference in the coupling strengths of the dot and ferromagnetic leads for the two spin orientations. | cond-mat.mes-hall |
26 Pbnm. Our model for the β-phase was therefore that of GdFeO 3 , 28 which exhibits the most common tilting distortion of a perovskite, corresponding to the Glazer tilt system a -a -b + . The refinement profile and the derived atom parameters are given in Figure 3 and Table 1, respectively. Nevertheless, the evolution of the BiFeO 3 structure can be followed quantitatively. The lattice parameters of the β-BiFeO temperature is approached, while the shorter Fe-O bond lengths slowly increase. In the Pbnm phase the FeO 6 octahedron is much more regular, with bond lengths being almost equal but occurring as three degenerate pairs, due to Fe being on an inversion centre. | cond-mat.mtrl-sci |
associated with the m-th excitation of the electron) it is the m-th of the participating energy levels which becomes the brightest in the center of the avoided crossings. 2) The order of the bright energy levels as they reappear outside the avoided crossing depends only on the position of the specific dot within the stack in which the hole becomes trapped. On the positive F side of the avoided crossing, lower energies corresponds to the hole localization in the lower dots [F > 0 pushes the hole down -see Fig. 1(a)]. The avoided crossings that are due to the hole transfer -with the spectacular modulation of the recombination probabilities -are only deformed by the confinement variation. The order of the subsequent avoided crossings in energy remains unchanged by the dot variation since their character is defined by the state of the electron and the electron energy splitting is huge when the dots are close enough for the hole to tunnel. | cond-mat.mes-hall |
Finally, the actual values of the classical fields (and consequently the order parameter p) are obtained by iterating the above procedure (i.e. calculated from the diagonalized Hamiltonian) until their initial and final values converge. As the main result of this self-consistent procedure, we obtained the increasing order parameter p as a function of the interladder interaction V for a realistic value of J = 0.4t [8] and for the three interesting hole dopings, see Fig. 3. First, we see that for finite values of V the CDW phase with the peculiar odd period is stable. Moreover, it is stable for realistic values of parameter V ∼ 0.5t. | cond-mat.str-el |
With periodic temporal and spatial boundary conditions, the partition function is periodic in θ under the 2π shift, and the system is invariant under the time reversal symmetry at θ = 0 and θ = π. However, with open boundary conditions, the partition function is no longer periodic in θ, and time reversal symmetry is generally broken, but only on the boundary, even when θ = (2n + 1)π. Ref. 2 gives the following physical interpretation. Time reversal invariant topological insulators have a bulk energy gap, but have gapless excitations with an odd number of Dirac cones on the surface. When the surface is coated with a thin magnetic film, time reversal symmetry is broken, and an energy gap also opens up at the surface. In this case, the low energy theory is completely determined by the surface term in Eq. | cond-mat.mes-hall |
(1) Here, α and σ = ρ -1 denote the Peltier and conductivity tensor, respectively. To detect tiny transverse temperature gradients, an AuFe-Chromel thermocouple has been attached directly to the wires that pick up transverse voltages. The difference between the isothermal and the adiabatic Nernst signal is given by the product of the thermopower and κ xy /κ and is shown in Fig. 3 (lower panel) at 14 T. It becomes clear that the magnitude of the thermopower that might add to the transverse voltages due to a transverse temperature gradient is much smaller than the measured e N . In 14 T, κ xy /κ xx is of the order of 1% so that its product with the thermopower S(T, 14T) results in a contribution of less than 10% of the measured e N (T, B). Thus, we can safely discuss the Nernst signal as being essentially isothermal. | cond-mat.str-el |
This change in the spectral line shape is accompanied by a slight reduction in the integrated XLD spectrum 44 I xld (ω)dω/ I iso (ω)dω (1) corresponding to a small change of the total quadrupole moment of 0.05(2). This might be interpreted as that the Fe 3d shell gains between 3 and 7% net z 2 -r 2 character at base temperature. The room temperature spectra of hydronium iron jarosite were identical to the spectra obtained on potassium iron jarosite within the experimental resolution as measured at the SRS (the red, broken line in the main panel of Fig. 2 for comparison). IV. LIGAND-FIELD MULTIPLET CALCULATIONS Due to the strong electrostatic interactions between the 2p core hole and the 3d levels in the final state, the 2p → 3d absorption spectrum is not simply proportional to the density of unoccupied 3d levels 45 as a function of energy. | cond-mat.str-el |
RESULTS AND DISCUSSION Fig. 3 demonstrates dependencies of the longitudinal resistance of two dimensional electrons on the magnetic field in sample N2. Two upper curves present dependencies obtained at different temperatures T=2.16K (dotted curve) and T=4.2K (solid curve) at zero dc bias. At small magnetic fields B <0.1T the magnetoresistance demon- strates the classical independence on the magnetic field [63]. At B >0.1T the electron spectrum is quantized and at temperature T =0.3K the resistance demonstrates quantum oscillations (not shown). An arrow marks the magnetic field B =0.1T above which the electron spectrum is modulated due to the quantization of the electron motion in magnetic fields. | cond-mat.mes-hall |
The picture is as follows: If the maxima of the modulation potential is at the center of the sample, the incompressible strip is formed at a higher magnetic field value, whereas, the edge incompressible strips become narrower at the lower field side. Hence, due to the larger incompressible strip at the bulk of the sample the plateau is shifted to the higher field, in contrast, due to the narrower (compared to the unmodulated system) edge strips the plateau is cut off at higher fields. Since, the edge incompressible strip becomes narrower than the extend of the wave function. For the odd parity, the edge incompressible strips become wider, therefore, the plateau extends to the lower B fields. The enhancement at the high field edge results from the two maximum in the proximity of the center. For a better visualization of the incompressible strip distribution we suggest reader to look at Fig. | cond-mat.mes-hall |
Consequently, we concluded that the magnetic structure is described by the ordering wavevector (1/3 1/3 0), the presence of 1/2-integer reflections occurring as a result of the charge ordering which decorates the lattice with differing magnetic moment on Fe 2+ and Fe 3+ sites with a periodicity of (1/3 1/3 1/2). Representational analysis with the (1/3 1/3 0) wavevector again yielded two allowed irreducible representations. For spins pointing along the c-axis, as suggested by the magnetization measurements, these representations correspond to ferromagnetic (FM) or antiferromagnetic (AFM) alignment of the two spins of the primitive basis (see Fig. 3). The AFM case can be ruled out immediately as the magnetic structure, in- cluding symmetry equivalent wavevectors, does not yield intensity at the (1/3 1/3 0) position. The FM coupling between spins in the basis, the only remaining symmetry allowed possibility, results in a ferrimagnetic structure as shown in Fig. | cond-mat.str-el |
We use 2000 k points in the first Brillouin zone, make the harmonic expansion up to l max =10 in atomic spheres, and set R mt ×K max to 7.5. The radii of the atomic spheres of Cr and others are chosen so that as high accuracy as possible is obtained. The volumes are optimized in terms of total energy, and the internal position parameters with a force standard of 3 mRy/a.u. The simplest antiferromagnetic structures are constructed by doubling the unit cells along the [100] and [110] directions. There are six Cr atoms in each of these doubled cells. We let three of the six Cr spins orient up and the other three down. | cond-mat.mtrl-sci |
T 4 -4 (4) = T 4 4 (4) * = - √ 70s + 20t r -12i √ 3t i 36 , T 4 -3 (4) = -T 4 3 (4) * = (1 -i) 2 √ 35s -7 √ 2t r + 3i √ 6t i 36 , T 4 -2 (4) = T 4 2 (4) * = i √ 10s - √ 7t r 9 , T 4 -1 (4) = -T 4 1 (4) * = (1 + i) 2 √ 5s - √ 14t r + 3i √ 42t i 36 , T 4 0 (4) = - 7s + 2 √ 70t r 18 . (22) The symmetrized second-rank tensor in the cubic frame is the same as for spinel. The effect of the angle α on the experimental spectrum can be considerable, as is illustrated in figure 3. From site symmetry to crystal symmetry We consider in this section another type of problem. We assume that we have calculated a symmetrized tensor T ℓ for a certain site A. We want to know the value of the same tensor for all the sites equivalent to A. | cond-mat.str-el |
[7], that the large values of Young's modulus of stiff coatings, such as diamond-like carbon films, can still account for their excellent lubricating properties, making use of Eqs. ( 5), ( 6) and ( 9). Combining Eqs. ( 5) and ( 6) with Eq. ( 9), we obtain for the condition for the n th order asperities to be multistable K E < L ′ n a q 0 h 0 4(1 -ν 2 ) ( πH 1 -H ) 1/2 ζ (1-H) n ,(12a) if the interfaces between each level asperity less than or equal to n th order are in the strong pinning regime (defined in the previous two sections), or K E < L ′ n L n q 0 h 0 4(1 -ν 2 ) ( πH 1 -H ) 1/2 (ζ n ζ a ) (1-H) 2 ,(12b) if the interfaces between each level asperity less than or equal to n th order are in the weak pinning regime. Since the n=0 order asperities are the ones most likely to be multistable, let us apply Eqs. | cond-mat.mtrl-sci |
According to Eq. ( 12), the self-energy contributed from the the source and drain coupling is obtained as, ΣK lm:l ′ m ′ (E) = ΣC lm:l ′ m ′ (E) δ m,0 tanh( E -µ S 2k B T ) +δ m,M-1 tanh( E -µ D 2k B T )(16) with ΣC (E) = ΣR (E) -ΣA (E), the correlated component of the self-energy. However, for the Keldysh component of the impurity-induced self-energy Σ imp,K (E) the result is not given in a closed form and should be calculated self-consistently as in the case of the retarded one via Eqs. ( 8) and (14). C. Electron density and current The ensemble average of n lm = b † lm b lm gives local electron density of the device and, consequently, the electron density distribution in Eq. ( 2) becomes n el ( r) = lm n lm χ l ( ρ)ψ(z m ). | cond-mat.mes-hall |
Eq. ( 1)]. For each set of nanodevice parameters we have calculated the lowest singlet (S) and triplet (T ) energy levels for the initial and final potential energy profiles (cf. Figs. 1 and2). In the absence of magnetic field, the three triplet states (T 0 , T ± ) are degenerate, therefore, we are dealing with the triply degenerate level T . | cond-mat.mes-hall |
The M(T) curves in Fig. 7a display a decrease of the magnetization and a lowering of the temperature onset (T C ) of ferromagnetic behavior as the thickness decreases. This trend has been reported earlier for (001) films of LCMO [3,13] and other R 1-x A x MnO 3 manganites [1,2] grown on different substrates. The same trend is displayed by (110)LCMO films (Fig. 7c). However, there are two fundamental differences: the decay of T C when reducing thickness is much less pronounced and the ferromagnetic transition is sharper for (110) films than for (001) ones. | cond-mat.mtrl-sci |
The localization radius corresponds to a binding energy of 110 meV which is ∼ 3nm for a light hole. The corresponding ρ(ε f ) ∼ 3 × 10 19 /cm 3 eV extracted is a very large number and normally above the value required for delocalization. One can see that the band pseudo-hopping expression can easily account for the dc conduction. We see that by selecting ν, one can create an e -( T 0 T )1/2 or e -( T 0 T ) 1/4 behavior without variable range hopping. Thus, if some experiments give one or the other law, it only means that the effective density of states is different. This has nothing to do with VRH. | cond-mat.mtrl-sci |
In Fig. 7 we have plotted QMC results for different system size showing again that these are well converged for N ≥ 64. Thus we conclude that the striking difference between GFT and QMC is not a mere finite size effect. The breakdown of magnetization in GFT occurs at a critical temperature T crit /J = 0.5 whereas no such breakdown exists in QMC. However the exposed maximum of the magnetization in QMC lies near the breakdown point. The differences between QMC and GFT in the temperature range T /J ≈ 0.3 . | cond-mat.str-el |
This generalizes the block structure of U(1) invariant matrices in Eq. 35 (where Tnn is denoted Tn ) to tensors of arbitrary rank k. The canonical decomposition in Eq. 85 is important, in that it allows us to identify the degrees of freedom of tensor T that are not determined by the symmetry. Expressing tensor T in terms of the tensors Tn 1 n 2 •••n k with N in = N out ensures that we store T in the most compact possible way. Notice that the canonical form of Eq. 85 is a particular case of the canonical form presented in Eq. | cond-mat.str-el |
The situation for Frenkel excitons is far more complex. We could at first think that Frenkel excitons have two lengths: the atomic wave function extension for the electron-hole pair "relative motion" and the Bohr radius for Coulomb interaction. Actually, Frenkel excitons have not one but two relevant Coulomb interactions: one for direct Coulomb processes between sites; the other for indirect processes, this last interaction being the one responsible for the excitation transfer between sites. In addition, as the electrons and holes on which the Frenkel excitons are constructed are not free, the proper effective mass to use in a "Bohr radius" is a priori unclear. Finally, the tight-binding approximation underlying Frenkel exciton corresponds to reduce the atomic wave function extension to zero, so that the length associated to the electron-hole "relative motion" de facto disappears. Due to these intrinsic difficulties and the expected interplay between these various "physical lengths", the many-body physics of Frenkel excitons is going to be far more subtle than the one of Wannier excitons, controlled by one length only. | cond-mat.mes-hall |
Computational models of diffusion-limited solidification have typically been divided into two broad camps -'front-tracking' models, in which one keeps track of the solidification interface explicitly (e.g. [2,3,4]), and 'phase-field' models, in which the solidification front is numerically smoothed and not explicitly tracked (e.g. [5,6]). To date, both techniques have had considerable success in modeling simple dendritic growth, but neither has been able to satisfactorily model complex morphological structures in the presence of strong faceting [7,8,9,10], owing to the appearance of dynamical and numerical instabilities. A third computation technique -local cellular automata (LCA) -has recently been applied to the problem of modeling faceted crystal growth with excellent initial success, and LCA models have produced realistic-looking snow crystal growth simulations in both 2D [11,12] and 3D [13]. To date, however these local lattice models have incorporated largely ad hoc growth rules (albeit physically motivated to some extent), so the connection between the resulting simulations and real crystal growth remains tenuous. | cond-mat.mtrl-sci |
In the same fashion the temperature rises rapidly on the Frenkel-like branch (Path 0 → 7 on the Fig. 3 and Fig. 4,(b))with a down-scan. The temperature goes down and the detuning between FE and WE reaches its minimum and an abrupt heating occurs. For the up-scan the sample is heated enough (close to the upper cross-point on Fig. 4,(b)) to maintain large detuning between the FE and WE but drops down rapidly when the FE and WE detuning reaches its maximum δ -= ∆ (See Fig. | cond-mat.mtrl-sci |
In (4), ψ σk is the fermionic annihilation field for an electron with energy-momentum vector k and spin σ, S q is a bosonic field describing spin-fluctuations near a momentum Q and g is the coupling constant measuring the strength of the interaction between fermionic and bosonic excitations. The last term in (4) stands for higher order terms in S. These are shown to be irrelevant for d > 2 and marginal for d = 2 and can therefore be neglected [22]. In the absence of the interaction, fermionic and bosonic excitations are described by the bare electron Green's function G 0 (k) and the bare spin susceptibility χ 0 (q) respectively G 0 (k) = z 0 iω -v F |k -k F | , χ 0 (q) = χ 0 ξ -2 0 + |q -Q| 2 + Ω 2 /v 2 s . (5) z 0 is the quasiparticle renormalization factor given by the Fermi liquid theory, and the electron dispersion is linearized with a Fermi velocity v F and Fermi momentum k F . The electron's chemical potential can incorporate effects of the condensed part of the bosonic field. The bare spin susceptibility is the usual Ornstein-Zernicke form where ξ 0 is the bare correlation length of spins and χ 0 ξ 2 0 is the static susceptibility. | cond-mat.str-el |
so-called either vortices or anti-vortices) are 1D, Bloch points are 0D [4,5]. Each class may serve as a boundary to the class of immediately-greater dimensionality: DWs are found at domain boundaries, Bloch lines inside domain walls to separate areas with opposite winding [3,5], and Bloch points separate two parts of a vortex with opposite polarities [4,6]. Beyond magnetism, the notions of DWs and vortices are shared by all states of matter ordered with a unidirectional order parameter, i.e. characterized by a vector field n with |n| = 1. Liquid crystals in the nematic state have a uniaxial order parameter. A strict analogue of magnetic materials is the common case of slabs with anchoring conditions at both surfaces: upon application of a magnetic or electric field perpendicular to the easy axis of anchoring a breaking of symmetry occurs known as the Fredericks transition [7], transforming the order parameter in a unidirectional one. | cond-mat.mtrl-sci |
Following the method of Keldysh and Onishchenko [2], Andryushin et al. [3] studied the behavior of the free MFEG by summing over all orders of Green's function contributions, they found an exact expression for the correlation energy of a MFEG (which dominates the interacting energy in the extreme many-flavor limit). This paper describes the derivation of a more versatile formalism, based on a path integral, which gives an exact expression for the total energy of the MFEG; the theory could apply with as few as six flavors where the exchange energy assumed small by Andryushin et al. [3] would be significant. As well as studying the uniform case, the previously unstudied density response of a MFEG not constrained to be uniform is investigated. The screening length-scales of the MFEG are shown to be short relative to the inverse Fermi momentum, suggesting that a local density approximation (LDA) might be a good approximation, motivating a gradient approximation. | cond-mat.str-el |
A plot of ∆V was already shown in Fig. 3 and a plot of the local density of states (i.e. the diagonal part of the spectral operator) was given in Fig. 1. Fig. 6 shows a plot of the evanescent Bloch solutions of the periodic Hamiltonian with potential V 0 , evaluated at the Fermi level, for device (c). | cond-mat.mes-hall |
We will construct an effective Hamiltonian for one tetrahedron unit and demonstrate that the T (-) 2 -orbital electrons only are sufficient to describe the low-energy one-particle excitations of Hamiltonian (1) in 3 T 1 phase. This construction can be regarded as a procedure of a real-space renormalization group. 32 Based on the results obtained in this section, we will proceed to the next procedure of the renormalization group in Sec. V. A. One-particle excitations Let us now investigate one-particle excitations in the 3 T 1 phase in detail. First, we examine which molecular orbitals in a tetrahedron play a dominant role in the oneparticle excitations upon changing electron number n d = 6 → 7 and n d = 6 → 5. To this end, we define matrix elements A n d ΓΓ ′ by A n d ΓΓ ′ ≡ gn d gn d +1 g n d +1 |d † Γ↑ |g n d g n d |d Γ ′ ↑ |g n d +1 . | cond-mat.str-el |
Usually current proceed by sequential tunneling as it is the case in quantum cascade lasers. In the pioneer work of Kazarinov, the current is expressed in a density matrix model where the resonance curve is found lorentzian with an homogeneous broadening given by the average value of elastic scattering matrix elements. As a consequence of the averaging the electrons tunnel between subbands conserving their in-plane wavevectors. More recently a refined model that include previously averaged-out second-order mechanisms was developed 4 . Second order scattering is known to yield gain without a net population inversion 5 through scattering assisted optical transitions, but it also affects more generally resonant tunneling, by allowing transitions between subband states of different wavevectors 4 . It is found that resonant tunneling occurs with conservation of the energy rather than the wavevector, contrarily to the first order case 6 . | cond-mat.mes-hall |
We use an atomistic approach for the stochastic dynamics of a local moment ferromagnet with the inclusion of spin transfer torque. Such a model is more appropriate for systems like the dilute magnetic semiconductor GaMnAs. Our use of simple approximations for the temperature dependence of the magnetic anisotropy, demagnetization field, and damping allow us to focus in the interplay between thermal fluctuations and spin transfer torque. We find that within this model, spin currents can change the size of the magnetization. We give an expression for this "spin-current longitudinal susceptibility", and propose an experimental scheme to measure this effect. We construct a Landau-Lifshitz-Bloch + Slonczewski (LLBS) equation to describe both longitudinal fluctuations and spin transfer torques. | cond-mat.mes-hall |
Thermodynamic measurements in fields up to 30T revealed a complex phase diagram associated with the triplet states, consisting of two distinct ordered states for fields oriented away from the c axis, attributed to the competition between single ion anisotropy and interdimer exchange in the frustrated lattice [18]. Here we probe the behavior of single crystals of Ba 3 Mn 2 O 8 via magnetization, heat capacity and magnetocaloric effect measurements in fields large enough to close the gap to the S z = 2 quintuplet states. Extraction magnetization measurements were performed in pulsed magnetic fields up to 60T in a 3 He refrigerator for fields applied perpendicular to the c axis [19]. Data were obtained by integrating the field derivative of the magnetization and are plotted for increasing fields in Figure 1(c). The data were cross calibrated with low field SQUID measurements to attain absolute values of the magnetization. Critical fields, evident as discontinuities in the slope of the magnetization, were determined from peaks in the second derivative of magnetization with respect to field. | cond-mat.str-el |
The obtained solutions for the displacement and strain fields in the (D) zones are trivially continued to the voided zone (V). These admissible continuations are non-physical except on the void boundary. However, the above symmetry considerations make clear that, modulo an appropriate translation, these continuations also provide the physical deformation fields in (C). Mutatis mutandis, the obtained expressions for the fields and the moduli are valid only up to f = f = π/8, due to the rescaling of a. This concentration acts as a "mechanically driven percolation" threshold, at half the close-packing value. It corresponds to the configuration where the summits (D) in Fig. | cond-mat.mtrl-sci |
The latter can be detected by producing a spin polarization in the normal metal (De Jong and Beenakker, 1995). Analogously, the former can be detected by producing a valley polarization in graphene, as we will discuss in Sec. V.C. The electron and hole in Fig. 8 are both from the conduction band. This intraband Andreev reflection applies if ε < E F . | cond-mat.mes-hall |
We tested this procedure on the moiré of the known R0 graphene phase. The simulated image, Fig. 5c 5b andc). The simulated image Fig. 5f also produces the observed laterals shifts in the minima of the fine-scale modulation (the "holes") by half the row separation when going along a ! 3 Ir direction. | cond-mat.mtrl-sci |
Namely, we consider the inner and outer edge channels at ν = 2 as two chiral boson fields and introduce the Luttinger-type Hamiltonian 3,19 to describe the equilibrium state. Second, we introduce the density-density interaction, which is known to be irrelevant in the low-energy limit. 18 This fact has no influence on the physics that we discuss below, because we focus on the processes at finite energy and length scale, which take place inside the MZI. A. Fields and Hamiltonian We assume that at filling factor ν = 2 there are two edge channels at each edge of the quantum Hall system and two chiral fermions associated with them and denoted as: ψ αj (x), α = 1, 2 and j = U, D. Here the subscript 1 corresponds to the fermion on outer channel, and 2 to the fermion on inner channel (see Fig. 4), while the index j stands for upper and lower arms of the interferometer. | cond-mat.mes-hall |
In this case the stable ground state is a spiral phase, and now if we attempt to move too close to Néel collinearity the real solution terminates. Such terminations of CCM solutions are very common and are very well documented. 21 In all such cases a termination point always arises due to the solution of the CCM equations becoming complex at this point, beyond which there exist two branches of entirely unphysical complex conjugate solutions. 21 In the region where the solution reflecting the true physical solution is real there actually also exists another (unstable) real solution. However, only the (shown) upper branch of these two solutions reflects the true (stable) physicsl ground state, whereas the lower branch does not. The physical branch is usually easily identified in practice as the one which becomes exact in some known (e.g., perturbative) limit. | cond-mat.str-el |
3(b)]. As clearly seen in Fig. 3(c), the HWHM of the spatial profile X 1/2 decreases with increasing magnetic field and saturates for B > 5 mT at the resolution limit of ∆ = 7.5 µm resulting from the finite sizes of pump and probe spots. Obviously, also the spin gradient ∇ s ∝ ∂θ/∂x in Fig. 3 depends on the in-plane magnetic field. We now concentrate on the spin gradient at the point x = ∆. | cond-mat.mtrl-sci |
It is therefore unnecessary to account for additional exponentially small contributions in the scattering rates W (p 1 , p ′ 1 ) and W (p ′ 1 , p 1 ), so that one can safely replace f 1 ≃ 1 and f 1 ′ ≃ 1 in Eqs. (B2) and (B3). The Fokker-Planck approximation exploits the fact that collisions typically induce small momentum changes of order O(T /v F ). For the following, it is convenient to introduce the momentum exchanges q 1 = p ′ 1 -p 1 . With this notation, W (p ′ 1 , p 1 ) describes the transition rate for the process in which a hole scatters with momentum transfer q 1 , from the initial state p 1 , and can thus be rewritten as W (p ′ 1 , p 1 ) = W q1 (p 1 ). Following the same prescription, the transition rate for the inverse process becomes W (p 1 , p ′ 1 ) = W -q1 (p 1 + q 1 ). | cond-mat.mes-hall |
The fundamental and superlattice peaks of the L1 0 phase are observed clearly in all the samples. We also note that with the increasing thickness, the intensity of the (200) reflection of L1 0 increases and for 100 nm thick film it becomes equal to the intensity of the (002) peak. The M-H curves for in-plane magnetization M(H ) and the morphology of these films are shown in Fig. 2(b) and Fig. 2(c,d,e) respectively. We observe an interconnected maze-like pattern of CoPt layer in the 20 nm thick film. | cond-mat.mtrl-sci |
Moreover, adsorption of benzene gives a work function lowering that is of a similar size for all surfaces (at a fixed molecule-surface distance). This leads to S = 0.9 at d = 3.6 Å, and S = 0.8 at d = 3.0 Å, which is not extremely far from the Schottky-Mott limit S = 1. The absolute size of the work function shift depends on the molecule-surface distance with the number at d = 3.0 Å being roughly twice as large as that at d = 3.6 Å. The sign of the work function shift, its relatively weak dependence on the metal, and its sensitivity to the molecule-metal distance all point to an interpretation in terms of the pillow effect. The effect is determined by the Pauli repulsion between the molecular and surface electrons, which decreases the surface dipole and therefore the work function. 34,37,115 Pauli repulsion critically depends on the overlap between the molecular and surface wave functions and therefore on the distance between the molecule and the surface. | cond-mat.mtrl-sci |
This indicates that Na 4 Ir 3 O 8 is a weak Mott insulator, which can readily undergo bandwidth and doping-controlled transitions to the nearby metallic phase. The weak Mott insulator scenario for Na 4 Ir 3 O 8 was first proposed in Ref. 6, in which the bandwidth-tuned transition to a metal was studied in detail. In this theory, the electrons are factorized into fermionic spinons, which carry spin but no charge, and bosonic rotors, which carry charge but no spin 19 . When a critical interaction strength is reached, the rotors become gapped, and the system becomes an insulator. On the other hand, the spinons remain gapless even in the insulator, forming a neutral spinon Fermi sea, responsible for the finite density of excitations seen in the specific heat. | cond-mat.str-el |
Room temperature ferromagnetism has been predicted for Mo doped In 2 O 3 films 13 and observed in Ni, Fe, and Co doped samples 14 as well as undoped In 2 O 3 11 . It has been shown that the electrical and magnetic properties of Cr:In 2 O 3 films are both sensitive to the oxygen vacancy defect concentration, and that the ferromagnetic interaction depends on carrier density 8 . While it has been suggested that ferromagnetism in Cr:In 2 O 3 films is carrier mediated 15 , the precise relationship between the spin transport properties of the charge carriers and the net ferromagnetic moment remains unclear. In this Letter we demonstrate that the charge carriers in undoped In 2 O 3 films have a significant spin polarization at helium temperatures. Furthermore, measurements on Cr doped In 2 O 3 samples yield quantitatively similar results to measurements on undoped samples, suggesting that transition metal dopants may not play any significant role in the development of ferromagnetic order. We prepared ceramic samples of In 2 O 3 (with a base purity of 99.99%) and In 2 O 3 doped with 2 at% Cr using a standard solid state process 16 . | cond-mat.mtrl-sci |
The steep decrease in the ordered moment below x = 0.6 calls for a different interpretation since Nd 3+ is normally immune from Kondo-type moment suppression. We believe that this effect actually arises from a competition between exchange and QP interactions, as was previously argued in the case of PrB 6 to explain the rather low value of the ordered Pr moment. 16 Here, however, the situation is more complex since two different types of QP moments, O 0 2 and O xy , associated with easy magnetic axes along z or in the xy plane, respectively, are possibly involved. This can explain why the moment reduction sets in just at the concentration where the Nd-like AF component starts to prevail. V. CONCLUSION The present NPD study provides an overview of magnetic ordering phenomena in the Ce x Nd 1-x B 6 series, from which four different regimes can be identified. At both ends of the composition range, x < 0.4 and x > 0.85 (not considered here), the compounds retain the same general behavior as in pure PrB 6 and CeB 6 , respectively, and only the values of their transition temperatures and order parameters are changed. | cond-mat.str-el |
A more serious disadvantage is the use of incoherent transitions to prepare a biexcitation. Owing to this, it is hard if possible at all to generate entangled pairs on demand, a functionality that is required in most quantum algorithms [23]. In this article, we address the rich potential of the newly proposed Josephson LED for quantum manipulation purposes. We show how to operate the device for ondemand production of entangled photon pairs. We demonstrate that Josephson LEDs may be used as a two-qubit quantum gates. Moreover, we show how to entangle the spin of a particle in one of the quantum dots with the polarization of an emitted photon. | cond-mat.mes-hall |
. . , affect the form of the distribution p(ln g) in 3D. Since the deviation from Gaussian distribution cannot be obtained from the DMPK equation, the generalization of the last has been proposed by Muttalib and co-workers. [5,6] The derivation the DMPK equation [2] is based on the parametrization of the transfer matrix T. For spin-less electrons and the time reversal symmetry of the system, T = u 0 0 u * √ 1 + λ √ λ √ λ √ 1 + λ v 0 0 v * , (2) where λ is a diagonal matrix with diagonal elements λ i = (1 + cosh 2x i )/2, and u, v are unitary matrices. [7] For sufficiently long systems, it is assumed that the elements of matrices u, v and λ are statistically independent. | cond-mat.mes-hall |
Here t represents the hopping amplitude of the site i to nearest neighbors i + δ, t ′ is the diagonal hopping amplitude to next nearest neighbors i + δ ′ , J is the exchange constant of the spin-spin interaction, c i,σ is the fermionic operator with excluded double occupancy, S i is the 1 2 -spin operator at site i, and n i is the site i number operator. We introduce the EPI dimensionless coupling constant, λ = g 2 ω 0 /4t, with the value λ = 1 dividing the weak and strong coupling regimes of the Holstein model in the adiabatic limit. Below we set = 1, t = 1, J = 0.3, t ′ = -0.25, ω = 0.15 and OC is in units of 2πe 2 . The one-hole ground state of the t-J model on 4 × 4 lattice is sixfold degenerate. This degeneracy between (±π/2, ±π/2), (0, π) and (π, 0) is partially removed by t ′ [19] providing a four-fold degenerate ground state at momentum (±π/2, ±π/2). Hence, one can naively expect that the OC should be sensitive to the value of t ′ . | cond-mat.str-el |
The spectrum measured above T N at temperature T = 355 K matches well to the one measured at T = 300 K. On the other hand, the pre-edge RIXS spectrum shows a noticeable change when crossing T N . Above T N , the P1 peak becomes broad, resulting in a decreased spectral intensity. In optical spectroscopy, evidence has accumulated for the sensitivity of MH excitations to local magnetic correlations, that is, to the alignments of the nearest neighbor (NN) spins. For example, LaMnO 3 is an A-type antiferromagnetic insulator, with a ferromagnetic coupling along the b-axis and an antiferromagnetic coupling along the c-axis. Upon magnetic ordering, the spectral intensity of the lowest energy excitation around 2 eV, which is associated with the MH excitation, increases for the b-axis but decreases for the c-axis. 41 Similar sensitivity of MH excitations to local magnetic correlations was reported in a recent RIXS study on a series of doped manganite systems. | cond-mat.str-el |
6. In addition to conductance, we studied thermoelectric effects in partially equilibrated wires, limiting ourselves to the most interesting regime L ≫ l 1 . The equilibration of the electron system has a dramatic effect on the thermopower and thermal conductance of the wire. As the length of the wire increases, the thermopower increases dramatically, from exponentially small values at L ≪ l eq to S ∼ T /eµ at L ≫ l eq , see Eq. (74). Conversely, the thermal conductance of the wire decreases due to the equilibration of the electron system in the wire from the Wiedemann-Franz value K = 2π 2 T /3h to zero, Eq. | cond-mat.mes-hall |
The O 2p bandwidth is much larger for MgO but this has little effect on the defect state, except per-haps contributing a small additional broadening. In spite of this large band gap variation, and the change in the volume (hence the near neighbor distance) the position of the unoccupied hole state changes rather little, staying 2±0.5 eV above the highest occupied state. The variation that does occur is nonmonotonic and may be due to the conduction bands that fall in energy through the series. For BaO, narrowest band gap member, the N hole band merges into the bottom of the conduction Ba d bands. Because the intra-atomic Coulomb interaction strength U is at least a few eV, these systems are in the regime U/W >> 1 (being W the N 2p hole bandwidth) so hopping of the holes will be inhibited, giving insulating behavior. This is the limit in which the LSDA+U method works well. | cond-mat.str-el |
Energy spectrum of three fermions A semi-analytic solution is available also for three interacting particles in a harmonic trap [52]. Figure 6 compares the exact energies of N = 3 unpolarized fermions (solid lines) with those obtained by the CI method (red [light gray] and blue [dark gray] bullets, respectively, for g < 0 and g > 0). The exact energies are obtained following the method reported in reference [52] (there are a few discrepancies with respect to The total angular momentum is M = 0 and calculations are performed for Nmax = 3 to 10. The interaction strength g is determined to give an energy for the two-particle system of E = 1.500 (a) and E = 2.500 (b). Solid lines are power-law fits to the data and the dashed lines show the exact values given by (9). our derivation + ). | cond-mat.mes-hall |
I. INTRODUCTION The last decade one of the most active research areas in solid state and materials science is the field of magnetoelectronics also known as spintronics. 1,2,3 The aim is to replace conventional electronics by new devices where the spin of the electrons and not the charge transfer plays the key role. A central problem in this field is the injection of spin-polarized current from a metal into a semiconductor. 4 In principle it is possible to achieve 100% spin-polarized injected current if the magnetic lead is a half-metallic material. These alloys are hybrids between normal metals and semiconductors. | cond-mat.mtrl-sci |
Coarse-grained cells where e 1 is larger than a threshold, show yielding (nonlinear and erratic σ 1 -e 1 ) and appreciable plastic flow, χ = 0. We have verified that these coarse-grained cells showing plastic deformation are indeed the NAZs reported above. We now focus on one coarse-grained cell, and study the time development of σ 1 , e 1 and χ as the transformation proceeds (Fig. 8). We find that at earlier times, the strains are small and the stress-strain response is elastic. Beyond a yield stress σ 1c , the stress-strain relation is nonlinear, giving rise to non-affine deformations χ = 0. | cond-mat.mtrl-sci |