4.1. Momentum balance of the colloid

In this section, we seek the velocity profile of a recirculation flow generated in a magnetic colloid situated in a closed rectangular channel and subject to non-uniform rotating magnetic field. The problem geometry is shown on the right of figure 1 and in figure 2. Recall that the channel dimensions along the x, y and z directions are denoted by l, $\; h$ and b, respectively. Let us consider a volume of a dilute magnetic colloid under magnetic field H rotating in the $xy$-plane with a non-zero field gradient (the absolute value of the magnetic field can in general vary along x, y and z). However, the field variation along the channel dimensions is considered to be relatively small, such that the field conserves its circular polarization in each point of the channel – the condition verified in our experiments (§ 2). As shown in a previous work (Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020), the rotating magnetic field creates micron-sized needle-like aggregates composed of several magnetic nanoparticles. The volume fraction $\varPhi $ of the aggregates is defined as the total volume of aggregates divided by the suspension volume. We will suppose that all the aggregates have the same size and a shape of a prolate ellipsoid of revolution characterized by the length-to-diameter ratio (aspect ratio) $r = L/D \gg 1$. The aggregates rotate synchronously with the field and exhibit a phase lag $\theta < {\rm \pi}/4$ between their orientation and the magnetic field vector. This is checked in the limit of low Mason numbers, $Ma < 1$. The torque balance on the synchronously rotating aggregate provides the following relationship between $\theta $ and $\; Ma$ established in the linear magnetization limit (Sandre et al. Reference Sandre, Browaeys, Perzynski, Bacri, Cabuil and Rosensweig1999; Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020):

(4.1a)\begin{gather}\textrm{sin}(2\theta ) = Ma \approx \frac{{8\beta {\eta _0}\omega (2 + \chi )}}{{3{\mu _0}H_0^2{\chi ^2}}},\end{gather}

(4.1b)\begin{gather}\beta = \frac{{{r^2}}}{{\textrm{ln}(2r) - 1/2}},\end{gather}

where ${\eta _0}$ is the suspending liquid (water) viscosity; $\chi = 22 \pm 5$ is the aggregate magnetic susceptibility (table 1) obtained experimentally in the previous work (Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020); ${\mu _0} = 4{\rm \pi} \times {10^{ - 7}}\;\textrm{H}\;{\textrm{m}^{ - 1}}$ is the magnetic permeability of vacuum; and $\beta $ is a dimensionless rotational friction coefficient evaluated in high aspect ratio limit $r \gg 1$ (Brenner Reference Brenner1974). In our experimental conditions, we are in a very-low-Mason-number limit $\; Ma \ll 1$, namely $Ma \le 1.3 \times {10^{ - 2}}$, resulting in an extremely small phase lag $\theta \approx Ma/2 \le 6.5 \times {10^{ - 3}}$.

According to our observations, the rotating aggregates translate along the magnetic field gradient and are mostly accumulated behind the middle plane $y\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel> \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }h/2$ of the channel, where the magnetic field is stronger (figure 2). In this § 4.1, we consider the motion of the whole colloid, which is characterized by the macroscopic quantities averaged over the whole volume of the colloid. Special attention must be paid to possible scale separation problems. As revealed in § 4.2, the average aggregate size is $L \times D = 30 \times 6$ μm, so that the size ratios of the aggregate size to the channel size are $L/h \approx 0.03{-}0.055$ and $D/b \approx 0.026$. In the studies of the fibre suspension rheology, it was established that the finite size of the fibres no longer affects the suspension viscosity or normal stresses at the ratios $L/h\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }0.2{-}0.33$ (Zirnsak, Hur & Boger Reference Zirnsak, Hur and Boger1994; Snook Reference Snook2015). Extrapolating these results to the present system, we expect that the colloid can still be considered as a continuous medium. The experimental time scale of the transient response related to the redistribution of aggregate concentration is much larger than the rotation period of aggregates. In this context, it is reasonable to consider the momentum balance equation averaged over the rotation period. Neglecting gravity, this equation takes the following general form valid for a suspension of elongated particles under magnetic field (Pokrovskiy Reference Pokrovskiy1978; Bashtovoi, Berkovsky & Vislovich Reference Bashtovoi, Berkovsky and Vislovich1988; López-López et al. Reference López-López, Kuzhir, Durán and Bossis2010):

(4.2a)\begin{gather}\rho \left( {\frac{{\partial \boldsymbol{v}}}{{\partial t}} + (\boldsymbol{v}\boldsymbol{\cdot }\boldsymbol{\nabla })\boldsymbol{v}} \right) = \textrm{div}\,\boldsymbol{\sigma },\end{gather}

(4.2b)\begin{gather}{\sigma _{ik}} ={-} P{\delta _{ik}} + 2{\eta _0}{\gamma _{ik}} + \sigma _{ik}^s + \sigma _{ik}^a + \sigma _{ik}^M,\end{gather}

(4.2c)\begin{gather}\sigma _{ik}^M ={-} {\textstyle{1 \over 2}}{\mu _0}{H^2}{\delta _{ik}} + {H_i}{B_k},\end{gather}

(4.2d)\begin{gather}\sigma _{ik}^a = {\textstyle{1 \over 2}}{\varepsilon _{ikl}}{K_l},\end{gather}

(4.2e)\begin{gather}{K_l} = {\mu _0}{(\boldsymbol{M} \times \boldsymbol{H})_l},\end{gather}

where $\rho $ is the colloid density; $\boldsymbol{v}$ and P are respectively the colloid velocity and pressure at a given point in the channel; ${\sigma _{ik}}$ are the components of the stress tensor $\boldsymbol{\sigma }$, with $\sigma _{ik}^s$ and $\sigma _{ik}^a$ being respectively symmetric and antisymmetric parts of the particle contribution to the viscous stress tensor, and $\sigma _{ik}^M$ being the Maxwell stress tensor; ${\gamma _{ik}} = (1/2)(\partial {v_i}/\partial {x_k} + \partial {v_k}/\partial {x_i})$ is the rate-of-strain tensor; ${\delta _{ik}}$ and ${\varepsilon _{ikl}}$ are Delta-Kronecker and Levi-Civita symbols, respectively; $\; {H_i}$ and $\; {B_k}$ are the i- or k-components of the magnetic field intensity vector $\boldsymbol{H}$ and magnetic flux density vector $\boldsymbol{B}$, respectively; $H = |\boldsymbol{H} |$; and ${K_l}$ is the l-component of the volume density $\boldsymbol{K}$ of a magnetic torque experienced by the magnetic colloid of a magnetization $\boldsymbol{M}$. The characteristic value of the particle stress $\sigma _{ik}^s$ is of the order of ${\sigma ^s}\sim \varPhi {r^2}{\eta _0}\dot{\gamma }$ with $\dot{\gamma }$ being a characteristic shear rate. In the dilute limit, $\varPhi {r^2} \ll 1$ valid for our experiments, the particle stress $\sigma _{ik}^s$ becomes negligible as compared with the solvent contribution, $2{\eta _0}{\gamma _{ik}}$, to the viscous stress, and is further omitted. With this in mind, and using the continuity equation, $\boldsymbol{\nabla }\boldsymbol{\cdot }\boldsymbol{v} = 0$, along with the magnetostatics Maxwell equations, $\boldsymbol{\nabla }\boldsymbol{\cdot }\boldsymbol{B} = 0$, $\boldsymbol{\nabla } \times \; \boldsymbol{H} = \boldsymbol{0}$, $\boldsymbol{B} = {\mu _0}(\boldsymbol{H} + \boldsymbol{M})$, the momentum balance equation (4.2a) becomes similar to that derived for the colloids of spherical magnetic particles (Bashtovoi et al. Reference Bashtovoi, Berkovsky and Vislovich1988):

(4.3a)\begin{gather}\rho \left( {\frac{{\partial \boldsymbol{v}}}{{\partial t}} + (\boldsymbol{v}\boldsymbol{\cdot }\boldsymbol{\nabla })\boldsymbol{v}} \right) ={-} \boldsymbol{\nabla }P + {\eta _0}{\nabla ^2}\boldsymbol{v} + {\boldsymbol{F}_m},\end{gather}

(4.3b)\begin{gather}{\boldsymbol{F}_m} = {\mu _0}(\boldsymbol{M}\boldsymbol{\cdot }\boldsymbol{\nabla })\boldsymbol{H} + {\textstyle{1 \over 2}}(\boldsymbol{\nabla } \times \boldsymbol{K}),\end{gather}

except for the viscosity ${\eta _0}$ which is replaced by the viscosity of the whole colloid by Bashtovoi et al. (Reference Bashtovoi, Berkovsky and Vislovich1988), the difference coming from the dilute limit considered in the present work. In the last equation, ${\boldsymbol{F}_m}$ stands for the volume density of the magnetic force, whose expression can be simplified for our experimental conditions.

First, it is shown in Appendix C that in our experimental conditions (including linear magnetization limit and the ratio ${L_H}/{L_\varPhi } \gg 1$ between the length scales of the magnetic and concentration field variations), the magnetic force density (4.3b) reduces to

(4.4)\begin{gather}{\boldsymbol{F}_m} \approx {\textstyle{1 \over 2}}\varPhi \varGamma {\mu _0}\boldsymbol{\nabla }({H^2}) + {\textstyle{1 \over 2}}(\boldsymbol{\nabla } \times \boldsymbol{K}),\end{gather}

where $\varGamma $ is given by (C6b) in function of the phase lag $\theta $ and the magnetic susceptibility $\chi $ of aggregates.

Second, we suppose that a characteristic shear rate of the induced shear flow is much smaller than the rotational frequency of the aggregates, $\dot{\gamma } \ll \omega $. This hypothesis, valid in the considered dilute regime, $\varPhi {r^2} \ll 1$, will be justified a posteriori once the velocity profile is calculated and the aggregate concentration determined (§ 4.5). With such a condition, the shear contribution to the hydrodynamic torque ${\boldsymbol{T}_h}$ experienced by an aggregate can be neglected and the torque balance in the inertialess limit will give us

(4.5)\begin{gather}\boldsymbol{K} \approx{-} n{\boldsymbol{T}_h} = 2{\eta _0}\beta \varPhi \boldsymbol{\omega },\end{gather}

with $\boldsymbol{\omega }$ being a vector magnitude of the field angular frequency, $n = \varPhi /{V_a}$ being the number density of aggregates, each having a volume ${V_a}$ and $\beta $ is given by (4.1b).

Third, in the low-Reynolds-number limit, valid for our experiments, the convective term $(\boldsymbol{v}\boldsymbol{\cdot }\boldsymbol{\nabla })\boldsymbol{v}$ can be neglected in the left-hand side of (4.3a). Finally, according to our observations, we consider only the steady-state regime at $t\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel> \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }100\;\textrm{s}$, meaning that the aggregate volume fraction $\varPhi $ and the fluid velocity $\boldsymbol{v}$ do not evolve with time. This last assumption will be revised in § 4.4 based on the results of the aggregate speed measurements. Applying the above conditions, (4.3) takes the following form:

(4.6)\begin{gather}- \boldsymbol{\nabla }P + {\eta _0}{\nabla ^2}\boldsymbol{v} + {\textstyle{1 \over 2}}\varPhi \varGamma {\mu _0}\boldsymbol{\nabla }({H^2}) + {\eta _0}\beta (\boldsymbol{\nabla }\varPhi \times \boldsymbol{\omega }) = {\bf 0}.\end{gather}

Tracking back to our experimental geometry, we can provide further simplifications. First, the angular speed $\boldsymbol{\omega }$ has the only non-zero $z$-component: ${\omega _z} ={-} \omega $ corresponding to the clockwise rotation in the $xy$-plane. Second, the channel length is much larger than two other dimensions, $l \gg b,h$ (figure 1), and considering the flow field far from the left and right borders of the channel, we can impose the single non-zero $x$-component of the velocity, ${v_x} = v(\kern0.7pt y,z)$ respecting the continuity equation, $\boldsymbol{\nabla }\boldsymbol{\cdot }\; \boldsymbol{v} = 0$. Third, in the experimental configuration of electromagnets, the $\boldsymbol{\nabla }({H^2})$ term has the only non-zero component along the y axis. Recall that the momentum balance equation is averaged over the aggregate rotation period, so that $\boldsymbol{\nabla }({H^2})$ term is given by the right-hand side of (2.3b).

These approximations allow us to rewrite (4.6) in the following component form:

(4.7a)\begin{gather}\frac{{\partial P}}{{\partial x}} = {\eta _0}\left( {\frac{{{\partial^2}v}}{{\partial {y^2}}} + \frac{{{\partial^2}v}}{{\partial {z^2}}}} \right) - \beta {\eta _0}\omega \frac{{\partial \varPhi }}{{\partial y}},\end{gather}

(4.7b)\begin{gather}\frac{{\partial P}}{{\partial y}} = \frac{1}{2}\varPhi \varGamma {\mu _0}\frac{{\partial {H^2}}}{{\partial y}} + \beta {\eta _0}\omega \frac{{\partial \varPhi }}{{\partial x}},\end{gather}

(4.7c)\begin{gather}\frac{{\partial P}}{{\partial z}} = 0.\end{gather}

The image processing of our experimental snapshots (§ 3, Appendix B) shows that, once averaged over time, the aggregate concentration does not depend on x and on z coordinates. Thus, the last term in right-hand side of (4.7c) can be omitted. Since the term $\partial {H^2}/\partial y$ is almost independent of $x,y,z$ in the present experimental conditions (cf. (2.3b)), integration of (4.7b), (4.7c) gives $P = \varPhi \varGamma {\mu _0}H_0^2y/(4{L_H}) + {\mathcal{G}}(x)$, where ${\mathcal{G}}(x)$ is an unknown function of x. With this in mind, the left-hand side of (4.7a), $\partial P/\partial x = \textrm{d}{\mathcal{G}}(x)/\textrm{d}\kern 0.06em x$ can only depend on x, while the right-hand side can only depend on y and z. With such a condition, (4.7a) can only hold if both its sides are independent of x, y and z, or rather $\partial P/\partial x = C$, with C being some unknown constant. Let us now introduce the following scaling factors for several physical magnitudes: $[v] = \beta \omega {\varPhi _0}h$ for the velocity, $[\varPhi ] = {\varPhi _0}$ – for the aggregate volume fraction and $[y] = h$, $[z] = b\;$ for the space coordinates, with ${\varPhi _0}$ being the average aggregate volume fraction in the suspension (before the aggregates migrate to the back region of the channel). The respective scaled quantities (hereinafter denoted by the tilde symbol) are obtained by dividing their dimensional counterparts by the scaling factors. Thus, (4.7a) can be rewritten in the following dimensionless form:

(4.8a)\begin{gather}\frac{{{\partial ^2}\tilde{v}}}{{\partial {{\tilde{y}}^2}}} + {\gamma ^2}\frac{{{\partial ^2}\tilde{v}}}{{\partial {{\tilde{z}}^2}}} = 2{C_1} + \frac{{\partial \tilde{\varPhi }}}{{\partial \tilde{y}}},\end{gather}

(4.8b)\begin{gather}\int_0^1 {\tilde{\varPhi }(\kern0.7pt \tilde{y})\,\textrm{d}\tilde{y}} = 1,\end{gather}

where $\gamma = h/b$ and ${C_1} = Ch/(2\beta {\eta _0}\omega {\varPhi _0})$ is a dimensionless unknown constant having a meaning of the dimensionless pressure gradient. Equation (4.8b) is nothing but the particle conservation condition. Equation (4.8a) is subjected to the non-slip boundary condition and a condition of zero flow rate across the channel that should be respected for the considered closed channel:

(4.9a)\begin{gather}\tilde{v}(0,\tilde{z}) = \tilde{v}(1,\tilde{z}) = \tilde{v}(\tilde{y},0) = \tilde{v}(\tilde{y},1) = 0,\end{gather}

(4.9b)\begin{gather}\int_0^1 {\textrm{d}\tilde{z}} \int_0^1 {\tilde{v}(\tilde{y},\tilde{z})\,\textrm{d}\tilde{y}} = 0.\end{gather}

Analytical solution of the boundary value problem (4.8a), (4.9a), (4.9b) is obtained by the Fourier series expansion method similar to that originally used by Boussinesq in 1868 for calculations of the velocity profile in a rectangular duct – see for instance Cornish (Reference Cornish1928). Alternatively, noticing that (4.8a) has a mathematical structure of Poisson equation, the solution can be obtained by the Green function method. The final expression for the velocity profile in terms of infinite series reads

(4.10a)\begin{gather}\tilde{v}(\tilde{y},\tilde{z}) = \sum\limits_{n = 1}^\infty {{D_n}{G_n}(\tilde{z})\,\textrm{sin}(n{\rm \pi} \tilde{y})} ,\end{gather}

(4.10b)\begin{gather}{G_n}(\tilde{z}) = \frac{{\textrm{sinh}(n{\rm \pi} \tilde{z}/\gamma ) + \textrm{sinh}(n{\rm \pi} (1 - \tilde{z})/\gamma )}}{{\textrm{sinh}(n{\rm \pi} /\gamma )}} - 1,\end{gather}

(4.10c)\begin{gather}{D_n} = \frac{{4[1 - {{( - 1)}^n}]}}{{{{(n{\rm \pi} )}^3}}}{C_1} - \frac{{2{K_n}}}{{n{\rm \pi} }},\end{gather}

(4.10d)\begin{gather}{K_n} = \int_0^1 {\textrm{cos}(n{\rm \pi} \tilde{y})\tilde{\varPhi }(\kern0.7pt \tilde{y})\,\textrm{d}\tilde{y}} ,\end{gather}

(4.10e)\begin{gather}{C_1} = \frac{1}{4}\frac{{\sum\limits_{n = 1}^\infty {{K_n}{M_n}[1 - {{( - 1)}^n}]/{{(n{\rm \pi} )}^2}} }}{{\sum\limits_{n = 1}^\infty {{M_n}[1 - {{( - 1)}^n}]/{{(n{\rm \pi} )}^4}} }},\end{gather}

(4.10f)\begin{gather}{M_n} = 2\gamma \frac{{\textrm{tanh}[n{\rm \pi} /(2\gamma )]}}{{n{\rm \pi} }} - 1,\end{gather}

where analytical expressions for the coefficients ${K_n}$ (4.10d) are provided in Appendix D based on the concentration profile $\tilde{\varPhi }(\kern0.7pt \tilde{y})$ determined experimentally in § 4.2 and theoretically in § 4.3.

The average dimensionless velocity across the channel thickness b is obtained by integration of (4.10a) over $\tilde{z}$:

(4.11)\begin{gather}\langle \tilde{v}\rangle (\kern0.7pt \tilde{y}) = \int_0^1 {\tilde{v}(\tilde{y},\tilde{z})\,\textrm{d}\tilde{z}} = \sum\limits_{n = 1}^\infty {{D_n}{M_n}\,\textrm{sin}(n{\rm \pi} \tilde{y})} ,\end{gather}

with expressions for ${D_n}$ and ${M_n}$ provided in (4.10c) and (4.10f), respectively.

Having obtained exact solution for the velocity profile, let us first analyse some limiting cases. First, in the case of homogeneous aggregate volume fraction $\tilde{\varPhi }(\kern0.7pt \tilde{y}) = 1$ or $\varPhi = \textrm{const}\textrm{.} = {\varPhi _0}$, we obtain zero velocity everywhere in the channel. This result directly follows from (4.6), in which the last term on the right-hand side vanishes, and applying curl operator to the other three terms, one obtains $\boldsymbol{\nabla } \times ({\nabla ^2}\boldsymbol{v}) = {\bf 0}$ – a linear equation with a unique solution $\boldsymbol{v} = {\bf 0}$ satisfying the boundary conditions (4.9). This points to the necessity of a heterogeneous concentration profile ($\boldsymbol{\nabla }\varPhi \ne {\bf 0}$) for generation of recirculation flows, in agreement with the basic claims of the present paper (cf. § 1), as also suggested in the literature on ferrofluid spin-up (Shliomis Reference Shliomis2021). Second, in a thick channel limit respecting the strong inequality $h \ll b \ll l$, the velocity is almost independent of the $\tilde{z}$ coordinate (except for the regions in a close proximity to the bottom and top channel walls, $\tilde{z} = 0\;\textrm{or}\;1$), and (4.10a), (4.11) reduce to

(4.12a)\begin{gather}\tilde{v}(\kern0.7pt \tilde{y}) = \langle \tilde{v}\rangle (\kern0.7pt \tilde{y}) \approx {I_1}(\kern0.7pt \tilde{y}) + 3(2{I_2} - 1){\tilde{y}^2} + 2(1 - 3{I_2})\tilde{y},\quad \textrm{at}\;\gamma = h/b \ll 1,\end{gather}

(4.12b)\begin{gather}{I_1}(\kern0.7pt \tilde{y}) = \int_0^{\tilde{y}} {\tilde{\varPhi }(\kern0.7pt \tilde{y})\,\textrm{d}\tilde{y}} ,\quad {I_2} = \int_0^1 {{I_1}(\kern0.7pt \tilde{y})\,\textrm{d}\tilde{y}} .\end{gather}

This approximate solution can be obtained by direct integration of (4.8a) neglecting the second term on the left-hand side.

Recall that dimensional flow velocity can be obtained by multiplying the dimensionless velocity ((4.10a) or (4.11)) by the scaling factor $[v]$:

(4.13a)\begin{gather}v(\kern0.7pt y,z) = \beta \omega {\varPhi _0}h\tilde{v}(\tilde{y},\tilde{z}),\end{gather}

(4.13b)\begin{gather}\langle v\rangle (\kern0.7pt y) = \beta \omega {\varPhi _0}h\langle \tilde{v}\rangle (\kern0.7pt \tilde{y}),\end{gather}

with $\tilde{y} = y/h$ and $\tilde{z} = z/b$.

Note that a similar scaling behaviour has been obtained by Bente et al. (Reference Bente, Bakenecker, von Gladiss, Bachmann, Cbers, Buzug and Faivre2021) for the velocity of the magnetic particle swarm under a coupled action of a rotating magnetic field and gravity. However, as mentioned in § 1, the results are obtained by considering zero tangential stress on the surface between the magnetic swarm and an ambient fluid, so that existence of such net surface seems to be necessary (at least from theoretical perspective) to generate the swarm motion. Such a surface is absent in our present study.

To get numerical values of the velocity, we need to define the factor $\beta $ depending on the aggregate aspect ratio r (4.1b), the average aggregate volume fraction ${\varPhi _0}$, as well as the aggregate concentration profile $\tilde{\varPhi }(\kern0.7pt \tilde{y})$ intervening into (4.10d). In the present study, the first two quantities ($\beta $ and ${\varPhi _0}$) are drawn from the experiments, while the concentration profile is both measured (§ 4.2) and evaluated by hydrodynamic diffusion approach (§ 4.3).

4.2. Aggregate size and concentration profile: experiments

The experimental distribution of aggregate lengths L, obtained through image processing as detailed in § 2, is shown in figure 3(a) for the following set of experimental parameters: f = 5 Hz, H ₀ = 6.4 kA m⁻¹, h = 1000 μm, ${\varphi _p} = 1.6 \times {10^{ - 3}}$. The distributions of aggregate diameters D and β parameters (4.1b) were also measured. The average value and standard deviation for these three quantities is equal to $L = 30 \pm 6\;{\rm \mu}\textrm{m}$, $D = 6 \pm 1\;{\rm \mu}\textrm{m}$ and $\beta = 14 \pm 2$. Surprisingly, these quantities remained unchanged (within the experimental errors) in the range of our experimental parameters f = 5–15 Hz, H ₀ = 3–9.5 kA m⁻¹, h = 550–1000 μm, ${\varphi _p} = (1.6{-}3.2) \times {10^{ - 3}}$. This is in contradiction to the theoretical models (Melle & Martin Reference Melle and Martin2003; Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020) predicting hydrodynamic rupture of aggregates leading to a decrease of their aspect ratio according to the scaling law $r \propto {\omega ^{ - 1/2}}$. However, in our previous experiments with homogeneous rotating magnetic field, we have already observed a very small variation of the aggregate length with the field frequency (cf. figure 6 of Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020). Such an effect can be tentatively explained by thermodynamic arguments, but first, we need to clearly define different volume fractions involved in the further analysis.

Figure 3.

(a) Experimental histograms of the aggregate length distribution. (b) Experimental and calculated profiles of the aggregate volume fraction across the channel y-dimension. Both graphs correspond to the following set of experimental parameters: f = 5 Hz, H ₀ = 6.4 kA m⁻¹, h = 1000 μm, ${\varphi _p} = 1.6 \times {10^{ - 3}}.$

The volume fraction of individual nanoparticles (without adsorbed sodium citrate layer) is denoted by ${\varphi _p}$ and is a well-defined quantity in our colloids (${\varphi _p} = (1.6{-}3.2) \times {10^{ - 3}}$). As mentioned in § 2, the individual nanoparticles are assembled into primary aggregates due to weakly attractive colloidal interactions in the absence of the magnetic field. However, only a small fraction of individual nanoparticles forms these aggregates, as revealed by hydrodynamic size distribution measurements. The volume fraction of primary aggregates ${\varphi _0}$ is defined as the volume occupied by the primary aggregates divided by the colloid volume. This quantity cannot be easily measured in experiments. It is supposed to vary linearly with the volume fraction ${\varphi _p}$ of individual nanoparticles. When the magnetic field is applied, the primary aggregates are self-assembled into the elongated secondary aggregates. Again, only a fraction of the primary aggregates forms the secondary ones. The remaining number of primary aggregates is characterized by the volume fraction $\varphi $, which decreases with time.

Tracking back to the explanation of the aggregate length independence of f, H ₀, h, ${\varphi _p}$, we notice that there exist two limitations for the aggregate maximal volume. On one hand, in the absence of the aggregate growth by absorption of neighbouring nanoparticles by the field-induced aggregates, the maximal aggregate volume is limited by hydrodynamic forces coming from either flow fields of neighbouring aggregates or short-ranged hydrodynamic interactions at collisions with neighbouring aggregates. This volume is expected to scale as ${V_{max1}} \propto {D^3}{[{\mu _0}{H^2}/({\eta _0}\omega )]^{1/2}}$ (Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020). On the other hand, in the absence of coalescence/fragmentation, the aggregate growth stops when the supersaturation $\varDelta = \varphi - \varphi ^{\prime}$ of the colloid (excess concentration of primary aggregates $\varphi $ with respect to the concentration at the aggregation threshold $\; \varphi ^{\prime}$) goes to zero. This corresponds to the maximal aggregate volume (Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020): ${V_{max2}} = {\varDelta _0}/(\varphi ^{\prime\prime}n)$, where $\varphi ^{\prime\prime} \approx \textrm{const}\textrm{.} \approx 0.3$ is the nanoparticle concentration inside the aggregates, n is the number density of the secondary aggregates and ${\varDelta _0 = {\varphi _0} - \varphi ^{\prime}}$ is the initial supersaturation of primary aggregates. The ${V_{max2}}$ value could be nearly independent of the parameters $\omega $, H and ${\varphi _p}$ provided that the initial supersaturation ${\varDelta _0}$ is independent of $\omega $ (Raboisson-Michel et al. Reference Raboisson-Michel, Queiros Campos, Schaub, Zubarev, Verger-Dubois and Kuzhir2020), and assuming that the number density n is linear with ${\varphi _0}$ (and ${\varphi _p}$), and at small aggregation threshold, $\varphi ^{\prime} \ll {\varphi _0}$, the ratio ${\varDelta _0}/n$ is nearly independent of ${\varphi _0}$ (${\varphi _p}$) and H. In the present situation, when both aggregate growth, coalescence and fragmentation are observed, the final aggregate volume is expected to be situated between ${V_{max1}}$ and ${V_{max2}}$. It is possible that in our experimental conditions, the maximal aggregate volume is closer to the ${V_{max2}}$ value, which is independent of $\omega $, H, ${\varphi _p}$ and h parameters. To get a grounded confirmation of this scenario, we need to reconsider the kinetics of field-induced phase assembly under rotating fields accounting for growth, coalescence and fragmentation events.

The experimental average value of the aggregate volume fraction is evaluated to be ${\varPhi _0} = (4.0 \pm 1.0) \times {10^{ - 4}}$ for ${\varphi _p} = 1.6 \times {10^{ - 3}}$ and ${\varPhi _0} = (6.5 \pm 2.3) \times {10^{ - 4}}$ for ${\varphi _p = 3.2 \times 10^{ - 3}}$ independently of the field frequency $\omega $, amplitude H and channel height h. High statistical errors of ${\varPhi _0}$ do not allow precise determination of the effect of nanoparticle concentration ${\varphi _p}$ on the concentration of the secondary aggregates ${\varPhi _0}$. However, we can discern a global tendency of increasing ${\varPhi _0}$ with increasing ${\varphi _p}$. Qualitatively, this behaviour is consistent with the above evaluation of the maximal volume ${V_{max2}}$ and of the initial supersaturation ${\varDelta _0} = {\varphi _0} - \varphi ^{\prime} \approx {\varphi _0}$ resulting in ${\varPhi _0} = n{V_{max2}} \approx {\varphi _0}/\varphi ^{\prime\prime}$. The ensemble of the experimental values determined in this section are summarized in table 1 along with their standard deviations.

The experimental aggregate concentration profile $\varPhi (\kern0.7pt y)$ across the channel was obtained through image processing (§ 2, Appendix B) and it was normalized by the average value ${\varPhi _0}$, namely $\tilde{\varPhi }(\kern0.7pt \tilde{y}) = \varPhi (\kern0.7pt y = \tilde{y}h)/{\varPhi _0}$. The experimental profile $\tilde{\varPhi }(\kern0.7pt \tilde{y})$ is represented by dots in figure 3(b), while the error bars correspond to the standard deviation between six measurement series. As observed on the snapshot of figure 2, the concentration profile is non-uniform: the aggregate volume fraction remains near zero in the front part of the channel at $\tilde{y}\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }0.6$; it shows an abrupt increase at $0.6\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y}\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }0.9$ followed by a decrease near the back wall at $0.9\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y} \le 1$. This last decrease is related to geometrical constraints. In fact, when spinning under a rotating field, the aggregates cannot penetrate the back wall, so that the distance between their geometrical centres and the back wall cannot be smaller than the aggregate half-length. This creates a kind of depletion layer near the back wall with a smaller aggregate volume fraction. For calculations of the velocity profile, we fitted the experimental concentration profile with a Gaussian function (red solid line in figure 3b), as follows:

(4.14a)\begin{gather}\tilde{\varPhi }(\kern0.7pt \tilde{y}) = A\,\textrm{exp}\left( { - \frac{{{{(\tilde{y} - {{\tilde{y}}_0})}^2}}}{{{{(2\delta )}^2}}}} \right),\end{gather}

(4.14b)\begin{gather}A = {\left[ {\delta \sqrt {\rm \pi}\left( {\textrm{erf}\left( {\frac{{1 - {{\tilde{y}}_0}}}{{2\delta }}} \right) + \textrm{erf}\left( {\frac{{{{\tilde{y}}_0}}}{{2\delta }}} \right)} \right)} \right]^{ - 1}},\end{gather}

where $\textrm{erf(} \ldots )$ stands for the error function; the value of the normalization constant A is chosen to respect the particle conservation condition (4.8b); and the fitting parameters ${\tilde{y}_0} = 0.9$ and $\delta = 0.07$ (reported in table 1) represent the position of the Gaussian peak and the distribution width, respectively. Note finally that the Gaussian fit (4.14) allows evaluation of the experimental length-scale of the concentration variation: ${L_\varPhi } = \sqrt 2 \delta h \approx 100\;{\rm \mu}\textrm{m}$ for the channel width $h = 1000\;{\rm \mu}\textrm{m}$.

4.3. Aggregate concentration profile: hydrodynamic diffusion model

Let us now try to evaluate the concentration profile independently from experimental data. From the first glance, this profile comes from the balance between the magnetic force pushing the aggregates toward the back wall and hydrodynamic interaction repulsing the aggregates from the back wall that have been extensively studied by different authors (see for example Hsu & Ganatos Reference Hsu and Ganatos1994; Mitchell & Spagnolie Reference Mitchell and Spagnolie2015; Zheng et al. Reference Zheng, Xu, Wang and Han2022). The aggregate position y can be evaluated by balancing both these forces. Using the full hydrodynamic resistance matrix of an ellipsoid translating and spinning along a solid wall (calculated by boundary integral method by Hsu & Ganatos (Reference Hsu and Ganatos1989, Reference Hsu and Ganatos1994)), we evaluate that the extremity of the aggregates (when they are oriented normal to the back wall) is separated from the back wall by a small distance ${h_w} \ll L$, recalling that $L \approx 30$ μm is the average aggregate length. This means that we should find all the aggregates confined near the back wall within a thin layer of a thickness ${\sim} L$ taking 3 % of the channel width. This is not consistent with the measured concentration profile which spreads over a 400 μm thick layer taking approximately 40 % of the channel width ($0.6\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y} \le 1$, cf. figure 3b). Such a discrepancy cannot be explained by the polydispersity of the aggregate length because the number of aggregates of the length $L > 100$ μm is only approximately 3 % (figure 3a). However, it is often observed in experiments that the geometric centre of an aggregate gets spinning around a neighbouring larger aggregate, as already mentioned in § 3. This overall motion of the aggregates can be imagined as a combination of a regular leftward translation along the back wall and a stochastic reciprocal translation in the x and y directions, as a result of hydrodynamic interactions with neighbouring aggregates. The stochastic part of this motion is responsible for spreading of the aggregate positions across the channel, leading to spreading of the concentration profile. It is thus convenient to mimic this process by hydrodynamic diffusion, as often suggested for colliding spherical particles in concentrated suspensions (see for instance Nicolai et al. (Reference Nicolai, Herzhaft, Hinch, Oger and Guazzelli1995)). The hydrodynamic diffusion has also been suggested in the mathematical model of magnetically assisted delivery of thrombolytics (Clements Reference Clements2016). However, the nature of this diffusion was not specified, and the diffusivity was not evaluated, which may have biased conclusions on the aggregate angular speed effect on generated macroscopic flow.

In the present case, the associated translational diffusivity can be evaluated through the mean free path approach applied to two-dimensional motion of aggregates (Pitaevskii & Lifshitz Reference Pitaevskii and Lifshitz2012): $\mathfrak{D} = (1/2)U\lambda $, where U is the average velocity of stochastic motion and $\lambda $ is the mean free path. In our case, the trajectory of the aggregate centre changes each time when the target aggregate is ‘trapped’ into the vortex flow field of a neighbouring aggregate and starts spinning around the latter. Thus, the mean free path $\lambda $ is a distance the target aggregate travels from the moment it is ‘left free’ from one aggregate and ‘trapped’ by another aggregate. Our observations show that in the back part of the channel, the large aggregates are situated close to each other and the small aggregates spin around the large ones (of a length $L$) with a linear velocity $U \approx \omega L/2$ traveling a distance $\lambda \sim L$ between two large aggregates. This results to the following scaling for the diffusivity:

(4.15)\begin{gather}\mathfrak{D} = \frac{{\omega {L^2}}}{{4{C_2}}},\end{gather}

where the average aggregate length can be chosen for the characteristic length L and ${C_2}$ is an unknown constant supposed to be of the order of unity and intentionally introduced into the denominator for further utility.

We need now to write down the mass conservation and momentum equations for the solid phase (field-induced aggregates) of the colloid using the general framework of Drew & Lahey (Reference Drew and Lahey1993), Nott & Brady (Reference Nott and Brady1994) and Morris & Boulay (Reference Morris and Boulay1999) applied to the inertialess limit and incorporating hydrodynamic diffusion:

(4.16a)\begin{gather}\frac{{\partial \varPhi }}{{\partial t}} + \boldsymbol{\nabla }\boldsymbol{\cdot }\boldsymbol{j} = 0,\end{gather}

(4.16b)\begin{gather}\boldsymbol{j} = \varPhi \boldsymbol{u} - \mathfrak{D}\boldsymbol{\nabla }\varPhi ,\end{gather}

(4.16c)\begin{gather}{\bf 0} = \boldsymbol{\nabla }\boldsymbol{\cdot }{\boldsymbol{\sigma }^s} + n{\boldsymbol{F}_h} + {\boldsymbol{F}_m},\end{gather}

where $\boldsymbol{j}$ is the volume flux density of aggregates, $\boldsymbol{u}$ is the average migration velocity of the aggregates, n is their number fraction, ${\boldsymbol{\sigma }^s}$ is the symmetric part to the particle contribution to the viscous stress tensor, ${\boldsymbol{F}_h}$ is the hydrodynamic drag force that the suspending liquid exerts on the aggregates, and ${\boldsymbol{F}_m}$ is the magnetic body force averaged over the whole suspension volume and provided by (4.4). The term $\boldsymbol{\nabla }\boldsymbol{\cdot }{\boldsymbol{\sigma }^s}$ has already been shown negligible with respect to ${\eta _0}{\nabla ^2}\boldsymbol{v}$, and since ${\eta _0}{\nabla ^2}\boldsymbol{v}$ is of the same order of magnitude as ${\boldsymbol{F}_m}$ (cf. (4.3a)), the term $\boldsymbol{\nabla }\boldsymbol{\cdot }{\boldsymbol{\sigma }^s}$ can also be omitted in (4.16b). As already stated, the hydrodynamic interactions between aggregates and the back wall cannot solely provide a correct estimation of the concentration profile. Moreover, tracking of aggregate angular motion with a high-speed camera shows that the aggregate angular speed does not change (within experimental precision) during the rotation period and is always equal to the field angular frequency, even for the aggregates whose end passes very close to the channel wall. This is possibly a consequence of the low Mason number $Ma \ll 1$ (table 1) depreciating hydrodynamic interactions even in close proximity to the wall. So, we can suppose that the wall interaction is at least not decisive in the current problem. In this context, as a first approximation, we can use an expression for the hydrodynamic drag force for the aggregates in an unbounded liquid neglecting wall interactions: ${\boldsymbol{F}_h} ={-} \boldsymbol{\xi }\boldsymbol{\cdot }(\boldsymbol{u} - \boldsymbol{v})$, where $\boldsymbol{\xi }$ is the tensor of hydrodynamic friction coefficients of aggregates and $\boldsymbol{v}$ is the suspension velocity defined by the momentum equation (4.3a). Equations can be further simplified if we recall that the concentration profile is steady state and all the physical magnitudes are independent of x. This leads us to ${j_x} = 0$, $\textrm{d}{j_y}/\textrm{d}y = 0$ in (4.16a), and with non-penetration condition at the front and back walls, we get zero aggregate flux across the channel:

(4.17)\begin{gather}{j_y} = \varPhi {u_y} - \mathfrak{D}\frac{{\textrm{d}\varPhi }}{{\textrm{d}y}} = 0.\end{gather}

Combining (4.4) and (4.16c), we get the expression for the aggregate migration speed $\boldsymbol{u}$ and its y component ${u_y}$:

(4.18a)\begin{gather}\boldsymbol{u} = \boldsymbol{v} + \boldsymbol{b}\boldsymbol{\cdot }{\boldsymbol{F}_m},\end{gather}

(4.18b)\begin{gather}{u_y} = {v_y} - \frac{{{b_{yx}}}}{n}\beta {\eta _0}\omega \frac{{\textrm{d}\varPhi }}{{\textrm{d}y}} + {b_{yy}}\frac{1}{2}{V_a}\varGamma {\mu _0}\frac{{\partial {H^2}}}{{\partial y}},\end{gather}

where ${V_a} = {\rm \pi}{D^2}L/6$ is the aggregate volume, $\boldsymbol{b}$ is the hydrodynamic mobility tensor with the components ${b_{yx}} = ({b_\parallel } - {b_ \bot })\,\textrm{cos}(\omega t - \theta )\,\textrm{sin}(\omega t - \theta )$, $\; {b_{yy}} = {b_\parallel }\,\textrm{co}{\textrm{s}^2}(\omega t - \theta ) + {b_ \bot }\,\textrm{si}{\textrm{n}^2}(\omega t - \theta )$ expressed through the mobilities ${b_\parallel }$, ${b_ \bot }$ parallel and perpendicular to the aggregate major axis. The last magnitudes are evaluated in high aspect ratio limit $r \gg 1$ (Brenner Reference Brenner1974):

(4.19a–c)\begin{gather}{b_{{\parallel} , \bot }} = \frac{2}{{{\eta _0}L{\xi _{{\parallel} , \bot }}}},\quad {\xi _\parallel } = \frac{{4{\rm \pi} }}{{\textrm{ln}(2r) - 1/2}},\quad {\xi _ \bot } = \frac{{8{\rm \pi} }}{{\textrm{ln}(2r) + 1/2}}.\end{gather}

Equation (4.18) still represents instantaneous value of the migration speed, provided that the expression (4.4) is valid not only for time-averaged value of ${\boldsymbol{F}_m}$ but also for its instantaneous value with the field gradient $\partial {H^2}/\partial y$ given by (2.3a). However, the first term in (4.18a) vanishes because the y component ${v_y}$ of the instantaneous velocity field of the whole colloid is zero thanks to the inertialess limit and specific boundary conditions ((4.9a), (4.9b)). In addition, averaging of (4.18b) over the rotation period of the aggregates cancels the second term on the right-hand side and finally gives

(4.20a)\begin{gather}{u_y} = {\rm \pi}\varGamma \psi \frac{{{\mu _0}H_0^2{D^2}}}{{48{\eta _0}{L_H}}},\end{gather}

(4.20b)\begin{gather}\psi = \frac{{2 + \textrm{cos(}2\theta )}}{{{\xi _\parallel }}} + \frac{{2 - \textrm{cos(}2\theta )}}{{{\xi _ \bot }}},\end{gather}

where $\varGamma $ is given by (C6b) and the length scale ${L_H} \approx 27\;\textrm{mm}$ of the magnetic field variation was defined in § 2 (cf. (2.2)). Substituting (4.20a) into (4.17) and integrating, we get the following expression for the concentration profile:

(4.21a)\begin{gather}\tilde{\varPhi }(\kern0.7pt \tilde{y}) = {C_3}\,\textrm{exp}(\alpha \tilde{y}),\end{gather}

(4.21b)\begin{gather}\alpha = \frac{{{u_y}h}}{\mathfrak{D}} = {C_2}{\rm \pi} \varGamma \psi \frac{{{\mu _0}H_0^2h}}{{12{\eta _0}\omega {r^2}{L_H}}},\end{gather}

where ${C_3}$ is an integration constant.

We can see that (4.21) simply reflects Boltzmann statistical distribution corresponding to the balance between diffusive and magnetophoretic fluxes. As a consequence, it gives a monotonous increase of the aggregate concentration across the channel with the maximal concentration on the channel back wall. In reality, the geometric centres of the aggregates cannot approach the back channel wall to a distance closer than their half-length. Thus, as already stated in § 4.2, we expect a depletion layer poor of aggregates near the back wall. However, due to polydispersity of the aggregate length, abrupt changes of the aggregate concentration across the depletion layer are not expected. The characteristic thickness of this layer ${L_d}$ is possibly defined by the length of the largest aggregates and can be taken as an adjustable parameter. A similar depletion layer (not allowing the aggregates to cross the wall) should in principle exist near the front wall. For the sake of simplicity, we can suppose that the concentration profile linearly decreases to zero within both depletion layers and is given by (4.21a) outside these layers:

(4.22a)\begin{gather}\tilde{\varPhi }(\kern0.7pt \tilde{y}) = {C_3} \times \left\{ {\begin{array}{@{}ll@{}} \dfrac{{\textrm{exp}(\alpha \kappa )}}{\kappa }\tilde{y},& \textrm{at}\;0 \le \tilde{y} < \kappa ,\\ \textrm{exp}(\alpha \tilde{y}),& \textrm{at}\;\kappa \le \tilde{y} < 1 - \kappa, \\ \dfrac{{\textrm{exp}(\alpha (1 - \kappa ))}}{\kappa }(1 - \tilde{y}),& \textrm{at}\;1 - \kappa < \tilde{y} \le 1, \end{array}} \right.\end{gather}

(4.22b)\begin{gather}{C_3} = {\left[ {\textrm{exp}(\alpha \kappa )\left( {\frac{\kappa }{2} - \frac{1}{\alpha }} \right) + \textrm{exp}(\alpha (1 - \kappa ))\left( {\frac{\kappa }{2} + \frac{1}{\alpha }} \right)} \right]^{ - 1}},\end{gather}

where $\kappa = {L_d}/h$ and we have used the aggregate conservation condition (4.8b) to get the expression for the constant ${C_3}$. The upper, middle and lower expressions of (4.22a) stand for the front depletion layer, the main middle layer of the channel and the back depletion layer, respectively.

The experimental concentration profile was fitted by (4.22a). The fitted curve is presented by a blue solid line in figure 3(b). We observe a rather good agreement between experiments and theory at reasonable values of the adjustable parameters: ${C_2} = 0.81$ (${C_2} = O(1)$ was expected) and ${L_d} = 100\;{\rm \mu}\textrm{m}$, resulting in $\kappa = 0.1$, which corresponds to the tail of the aggregate length distribution (figure 3a). The sharp peak at $\tilde{y} = 1 - \kappa = 0.9$ of the theoretical concentration profile comes from the depletion layer approximation imposing an abrupt change of the exponentially growing concentration outside the layer to linearly decreasing concentration inside the layer. However, such a sharp profile is not shocking and seems to be well compared to the experimental shape. Furthermore, (4.21a), (4.21b) allow us to define the expression for the length scale of the concentration variation:

(4.23)\begin{gather}{L_\varPhi } = \frac{h}{\alpha } = \frac{{12{\eta _0}\omega {r^2}}}{{{C_2}{\rm \pi} \varGamma \psi {\mu _0}H_0^2}}{L_H} \approx \frac{{36}}{{7{C_2}}}{L_H}Ma\sim {L_H}Ma,\end{gather}

where the Masson number is defined by (4.1a), and the pre-factor $36/(7{C_2}) = O(1)$ is a result of the approximations $r \gg 1$, $\chi \gg 1$ and $\theta ,Ma \ll 1$. Note that the later strong inequality allows using the values of $\varGamma $ and $\psi $ parameters ((C6b) and (4.20b)) at $\theta = 0$, namely

(4.24)\begin{gather}\varGamma \approx \chi ,\psi \approx \frac{3}{{{\xi _\parallel }}} + \frac{1}{{{\xi _ \bot }}}\quad \textrm{at}\;Ma,\;\theta \ll 1.\end{gather}

At the given magnetic field frequency and amplitude, f = 5 Hz, H ₀ = 6.4 kA m⁻¹, the Mason number is $Ma \approx 1.1 \times {10^{ - 3}}$, so with ${L_H} \approx 27\;\textrm{mm}$, we get ${L_\varPhi } \approx 190\;{\rm \mu}\textrm{m}$, which is consistent with the measured concentration profile (figure 3b) and with experimentally evaluated length scale ${L_\varPhi } \approx 100\;{\rm \mu}\textrm{m}$ (§ 4.2). Anyway, both experimental and theoretical evaluation of the concentration variation length scale clearly show that it is much lower than the field variation length scale, ${L_\varPhi } \ll {L_H}$, which is the consequence of the low-Mason-number limit $Ma \ll 1$, as inferred from (4.23). Recall that such scale separation is important for simplification of the magnetic force expression to (4.4) (cf. Appendix C).

4.4. Velocity profiles: experiments and comparison with the theory

Let us now focus on the velocity profiles of the recirculation flow in the closed microchannel, recalling that experimental and theoretical profiles describe the effective velocities of the whole colloid without a distinguishing of aggregate motion from the suspending fluid motion. An example of the velocity vector field obtained experimentally from the PIV analysis for a moment of time $t = 60\;\textrm{s}$ elapsed from the moment of the magnetic field application is shown in figure 4. It is clear from this figure that the velocity field is in general strongly perturbed by the presence of rotating aggregates. One clearly observes clockwise flow vortexes within the area swapped by the aggregate clockwise rotation. Interference between the vortexes created by neighbouring aggregates gives a rather complicated flow pattern. However, one clearly distinguishes a regular rightward flow in the channel's front layer free of aggregates. As already discussed in § 3, this rightward flow arises to compensate the leftward one induced by the aggregate translation along the back wall.

Figure 4.

Velocity vector field deduced from the PIV analysis taken at a moment of time $\; t = 60\;\textrm{s}$ and for the following set of experimental parameters: f = 5 Hz, H = 6.4 kA m⁻¹, h = 1000 μm, ${\varphi _p} = 1.6 \times {10^{ - 3}}$. The coloured arrows stand for the velocity vectors whose x and y components enter (green arrows) the $10 \times 5$ px frame⁻¹ window (cf. Appendix A) or lie outside this window (brown arrows). White thin rods stand for the position of aggregates, within which the instantaneous velocity profile cannot be established.

To smooth the effects of the flow irregularities introduced by rotating aggregates, the longitudinal ($x$-) component of the velocity profiles was averaged over either the observation window length l or observation period T, with the respective average magnitudes ${\langle {v_x}(\kern0.7pt y,t)\rangle _l}$ or ${\langle {v_x}(x,y)\rangle _T}$, defined in Appendix A, with the time t being counted from the beginning of the steady-state regime, thus 100 s after the moment of the magnetic field application. The experimental colourmaps of the ${\langle {v_x}(\kern0.7pt y,t)\rangle _l}$- and ${\langle {v_x}(x,y)\rangle _T}$-dependencies are shown in figure 5(a,b), respectively. First, we clearly observe a recirculation flow in both these figures: the back layer (upper half of the colourmaps) exhibits the leftward flow with negative longitudinal speeds and the front layer (lower half of the colourmaps) shows a rightward flow with positive speeds. Second, a few white-coloured vertical bands sometimes appear in the spatiotemporal colourmap ${\langle {v_x}(\kern0.7pt y,t)\rangle _l}$ (figure 5a). These bands correspond to near zero velocity and point out a go-and-stop (or rather fast and slow) sequence of the recirculation flow. This sequence visually corresponds to the fluctuation of the aggregate amount (or rather instantaneous volume fraction ${\varPhi _0}(t)$) in the observation window: smaller ${\varPhi _0}$ results in slower recirculation, higher ${\varPhi _0}$ results in stronger flows, in agreement with (4.13). However, excluding these fluctuations from consideration, we can claim that the steady-state velocity profile globally holds for the whole analysed period of 200 s starting from the beginning of the steady-state regime until the end of the experiment. Third, the spatial colourmap ${\langle {v_x}(x,y)\rangle _T}$ in figure 5(b) shows that the velocity field averaged over time is globally independent of the longitudinal coordinate with some fluctuations likely appearing as a result of a relatively small number of aggregates per observation window increasing the dispersion of the statistical averaging. An independence of the velocity profile on x is likely ensured thanks to the fact that the observation window is centred with respect to the channel centre and it is a few times shorter than the channel full length, such that the fringing effects from the left and the right borders of the channel are not perceived.

Figure 5.

(a) Spatiotemporal ${\langle {v_x}(\kern0.7pt y,t)\rangle _l}$ and (b) 2-D spatial ${\langle {v_x}(x,y)\rangle _T}$ experimental velocity maps averaged over one of six movies for the following set of experimental parameters: f = 5 Hz, H = 6.4 kA m⁻¹, h = 1000 μm, ${\varphi _p} = 1.6 \times {10^{ - 3}}$. The red and blue colours stand respectively for the rightward and leftward flows.

Finally, let us consider the velocity profiles ${\langle {v_x}(\kern0.7pt y)\rangle _{l,T}} \equiv \langle v\rangle (\kern0.7pt y)$ averaged over time and length of the observation window, as detailed in Appendix A. Such averaging allows smoothing longitudinal and temporal fluctuations, and focusing on the net effect of the rotating magnetic field on generated recirculation flows. The experimental averaged velocity profiles $\langle v\rangle (\kern0.7pt y)$ are compared with the calculated profiles in figure 6 using either Gaussian fit of the experimental concentration field ((4.14), figure 6a) or theoretical concentration field evaluated through hydrodynamic diffusion approach ((4.22), figure 6b). Analytical expressions for the coefficients ${K_n}$ (4.10d) intervening into the velocity profile (4.10a) are provided in Appendix D for both cases of the concentration profiles. Points correspond to the data averaged over six experimental runs. Solid lines correspond to the upper and the lower limits of the calculated profile which were obtained as ${v_{up}} = v(1 + \varepsilon )$ and ${v_{low}} = v(1 - \varepsilon )$, where $v = \langle v\rangle (\kern0.7pt y)$ is the velocity profile calculated using (4.11), (4.13b) using the average experimental values of $\beta ,\;{\varPhi _0},\;h$, and $\varepsilon \approx {({(\Delta \beta /\beta )^2} + {(\Delta {\varPhi _0}/{\varPhi _0})^2} + {(\Delta h/h)^2})^{1/2}}$ is the relative velocity error evaluated using experimental standard deviations $\Delta \beta ,\;\Delta {\varPhi _0},\;\Delta h$ of $\beta ,\;{\varPhi _0},\;h$ (cf. table 1).

Figure 6.

Experimental (symbols with error bars) and theoretical (solid lines) velocity profiles ${\langle {v_x}(\kern0.7pt \tilde{y})\rangle _{l,T}} \equiv \langle v\rangle (\kern0.7pt \tilde{y})$ averaged over time and length of the observation window and presented for the following set of experimental parameters: f = 5 Hz, H = 6.4 kA m⁻¹, h = 1000 μm, ${\varphi _p} = 1.6 \times {10^{ - 3}}$. The calculated velocity profiles correspond to either (a) Gaussian fit of the concentration profile (4.14) or (b) the concentration profile evaluated by the hydrodynamic diffusion approach (4.22). The shaded regions in panels (a) and (b) encompass the region of the model validity.

As inferred from this figure, both experimental and theoretical velocity profiles reflect recirculation flows. In the close vicinity to the back wall, $0.9\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y} < 1$, a distinguishable rightward flow is observed in experiments and confirmed by the model. This flow is likely associated to the propulsion of the fluid layer to the right along the wall by the upper end of the aggregates spinning in the clockwise direction. In the region still situated in the back part of the channel but further from the wall, $0.5\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y}\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }0.9$, a leftward intense flow is observed and could be assigned to the propulsion of the liquid layer to the left by lower ends of the aggregates. The flux generated by the upper ends of the aggregates is smaller because of the proximity with the wall imposing non-slip condition. The front layer of the channel, $0 < \tilde{y}\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }0.5$ exhibits an intense rightward flow. As pointed out before, this layer is almost free of aggregates (cf. figure 3b) and the rightward flow arises to compensate the leftward one in the back layer. By integration of the experimental velocity profile, we have checked that the zero total flux condition is satisfied at a maximal error of 1 % (with respect to the rightward or leftward flux). Note that experimental data mostly fit into the confidence interval of the theoretical model (shaded region between the two solid lines in figure 6a,b) except for the middle plane layer in figure 6(a) and an upper layer in figure 6(b). Recall that calculation of the velocity profile is free of adjustable parameters once the concentration profile is defined. This could be an argument for the validity of basic assumptions used in our model. If we compare theoretical predictions of the velocity profile, the concentration field predicted by the hydrodynamic diffusion approach overestimates the rightward flux near the back wall at $0.9\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel< \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }\tilde{y} < 1$ (figure 6b) possibly because of a sharp concentration peak observed at $\tilde{y} = 0.9$ (blue line in figure 3b). Such a sharp peak is absent in the Gaussian fitted concentration profile (red line in figure 3b) and the rightward flux near the back wall is quite well reproduced at the price of a higher leftward flux observed in figure 6(a).

4.5. Parametric study of the recirculation intensity

It is now important to inspect the effect of different experimental parameters ( f, H ₀, h, ${\varphi _p}$) on the recirculation flow. Note that in our experimental system, the effect of the magnetic field gradient is fully represented by H ₀ and ${L_H}$ parameters (2.2), with the length scale ${L_H} \approx 27$ mm of the magnetic field variation being fixed by the size of the Helmholtz coils. For the sake of comparison, it is instructive to define an integral parameter (not depending on the position y within the channel) characterizing the intensity of the recirculation. We choose for this purpose the absolute velocity averaged across the channel:

(4.25)\begin{gather}q = \frac{1}{h}\int_0^h {\left|{\langle v\rangle (\kern0.7pt y)} \right|\,\textrm{d}y} = \beta \omega {\varPhi _0}h\int_0^1 {\left|{\langle \tilde{v}\rangle (\kern0.7pt \tilde{y})} \right|\,\textrm{d}\tilde{y}} ,\end{gather}

where the middle part of (4.25) applies to the experimental data and the right-hand part to the theoretical velocity profile. The integral of the theoretical velocity profile over the channel width is evaluated numerically. As in the case of the velocity profile $\langle v\rangle (\kern0.7pt y)$, the theoretical evaluation of the magnitude q is subjected to errors related to experimental parameters $\beta ,\;{\varPhi _0},\;h$. Thus, we define a medium experimental value q using average values of $\beta ,\;{\varPhi _0},\;h$ and the confidential interval ${q_{low}} \le q \le {q_{up}}$, with ${q_{up}} = q(1 + \varepsilon )$ and ${q_{low}} = q(1 - \varepsilon )$, with the relative error $\varepsilon $ being defined in § 4.4.

Before analysing parametric behaviours of the average absolute velocity q, let us establish the scaling law for this magnitude based on (4.7a). The two terms on the right-hand side of that equation scale as ${\eta _0}|{{\nabla^2}v} |\sim {\eta _0}v/L_v^2$ and $\beta {\eta _0}\omega |{\partial \varPhi /\partial y} |\sim \beta {\eta _0}\omega {\varPhi _0}/{L_\varPhi }$, which, making use of (4.23) for ${L_\varPhi }$ and (4.1b) for $\beta $, finally gives

(4.26)\begin{gather}q\sim v\sim \frac{{\beta \omega {\varPhi _0}L_v^2}}{{{L_\Phi }}} = {\mathcal{F}}(\omega ,{H_0})\frac{{{\varPhi _0}\varGamma \psi {\mu _0}H_0^2}}{{{\eta _0}{L_H}({b^{ - 2}} + {h^{ - 2}})}},\end{gather}

where $L_v^2 \approx 2/({b^{ - 2}} + {h^{ - 2}})$ is the square of the velocity variation length scale taken equal to that for a channel with an elliptic cross-section. We see that using the expression (4.23) for the length scale ${L_\varPhi }$ of the concentration variation rules out the field angular frequency $\omega $ from (4.26). However, the frequency is not ruled out from the exact expression (4.25) for q, because it intervenes in the parameter $\alpha $ (4.21b) which affects the velocity in a nonlinear way. Because of similar reasons, the magnetic field dependence of q is not mandatorily quadratic. To stress the fact that the scaling law (4.26) does not reflect precisely the $\omega $ and H ₀ behaviours, we intentionally add a pre-factor ${\mathcal{F}}(\omega ,{H_0})$ to the right-hand side of (4.26).

The scaling (4.26) also allows checking the low shear rate hypothesis supposed in the present model (§ 4.1). Using the middle part of (4.26), the shear rate in the $xy$-plane scales as $\dot{\gamma } = |{\textrm{d}v/\textrm{d}y} |\sim \beta {\varPhi _0}\omega L_v^2/({L_\varPhi }h)$. Using experimental value ${L_\varPhi } \approx 100\;{\rm \mu}\textrm{m}$ (§ 4.2) and other values presented in table 1, we get $\dot{\gamma } \approx (6 \times {10^{ - 3}} - 1.5 \times {10^{ - 2}})\omega $. Thus, the supposed limit $\dot{\gamma } \ll \omega $ holds in our experiments.

The average absolute velocity q is plotted against the field frequency f in figure 7(a,b) for different magnetic field amplitudes H ₀ but for fixed magnetic nanoparticle concentration ${\varphi _p} = 1.6 \times {10^{ - 3}}$ and fixed channel width h = 1000 μm. The medium values $q(f)$ are plotted as solid lines in figure 7(a), while confidential intervals are presented as shaded regions in figure 7(b) respecting the same colourmap as that of the experimental points. It is quite difficult to discern an unambiguous tendency provided that the experimental error bars are very large, and the theoretical confidential intervals overlap to some extent. Globally, we could presume an increasing dependency of the recirculation intensity on the magnetic field amplitude. This behaviour can be qualitatively explained by sharpening of the concentration profile with increasing magnetic field (and consequently magnetic field gradient). This leads to coarsening of the length scale ${L_\varPhi }$, which leads to stronger recirculation in accordance with the scaling law (4.26). The effect of the field frequency could be somewhat more complicated. First, in the limit of nearly constant aggregate length revealed in experiments (§ 4.2), the parameter $\beta \sim {(L/D)^2}$ is not affected by the frequency, so the multiplier before the integral on the right-hand part of (4.25) is linear with the frequency. Second, the hydrodynamic diffusion is more intense at higher angular frequency of aggregates (4.15), and the aggregate concentration profile becomes more homogeneous with increasing frequency. This should increase the length scale ${L_\varPhi }$ and decrease the intensity of the recirculation, according to the middle term of (4.26). Third, with the increasing frequency, the phase lag $\theta $ between the field and aggregate orientations increases. This decreases the aggregate magnetization according to (C6a) and the parameter $\varGamma $ (C6b). This also decreases the average hydrodynamic mobility expressed through the parameter $\psi $ (4.20b). The product $\varGamma \psi $ in (4.26) leads therefore to a decrease of q with $\omega $. However, as already stated in § 4.3, in the present limit of very low Mason numbers, $Ma \ll 1$, the variation of $\varGamma \psi $ with $\theta $ is negligible (of the order of ${\theta ^2}$), so that this third effect can be ruled out for our experimental conditions. The competition between both first effects defines the final shape of the $q(f)$ dependencies at the fixed magnetic field. The model predicts a plateau for intermediate frequencies, at least at H ₀ = 6.4 and 9.5 kA m⁻¹, which is consistent with the scaling law (4.26). The experiments seem to follow a similar tendency, even though it is hardly distinguished because of the error bars overlap. In what concerns agreement between theory and experiments, most of the experimental points enter (at least by their error bars) to the confidence interval of the model. However, larger frequency and magnet field intervals (not accessible in the current experimental set-up) should be studied in the future.

Figure 7.

Experimental and theoretical dependencies of average absolute velocity on the frequency of the rotating magnetic field of different amplitudes H ₀ and for fixed nanoparticle concentration ${\varphi _p} = 1.6 \times {10^{ - 3}}$ and fixed channel width h = 1000 μm. Points correspond to experiments. Medium calculated values of q are plotted as solid lines in panel (a). The confidence intervals of the model are represented by shaded regions in panel (b). These regions follow the same colourmap as the experimental symbols.

In what concerns the effect of the particle concentration and the channel width, a few collected data (with 5 runs for each set of parameters) are summarized in table 2. Globally, the experiments suggest an increase of the average absolute velocity with the volume fraction of nanoparticles in agreement with the model: doubling the nanoparticle concentration ${\varphi _p}$ increases the volume fraction ${\varPhi _0}$ of the secondary field-induced aggregates, as inferred from experiments (§ 4.2) – see also the second column of table 2 for experimental ${\varPhi _0}$ values. This leads to an increase of recirculation intensity in proportion to ${\varPhi _0}$, as suggested by (4.26). It should be stressed however that the increase of ${\varPhi _0}$ and q with ${\varphi _p}$ is weaker than linear. The recirculation intensity q also seems to show a weaker than linear increase with the channel height h, both in experiments and in theory, which is qualitatively captured by the scaling law (4.26). Indeed, when approaching the thin channel limit, $b/h \ll 1$, the viscous dissipation is mostly governed by the channel thickness b, while h is ruled out from the length scale ${L_v}$ and from the scaling for q. In any event, the experimental and theoretical confidential intervals of the data for different h values overlap. Nevertheless, the conclusions on ${\varphi _p}$ and h effects on recirculation intensity should be taken with care because of overlapping error bars.

Table 2.

Experimental and theoretical values of the average absolute velocity for two different ${\varphi _p}$ and h values. For all the data in this table, H ₀ = 9.5 kA m⁻¹ and f = 5 Hz.

Based on these comparisons, we can claim that the model reproduces at least semi-quantitatively the major experimental trends. From the application perspective, it could be interesting to project these results to the clot dissolution application. First, we need to stress that the segment of the vascular network between the inlet to the blocked vessel and the blood clot is the most difficult and the most important pathway for the transport of the thrombolytic drug. For this reason, the speed of a few μm/s of a flow generated in an initially blocked vessel is believed to be high enough for efficient drug delivery (Clements Reference Clements2016). We note that the average recirculation speeds $q = 5{-}8$ μm s⁻¹ achieved in our experiments are approximately an order of magnitude higher than an effective speed (${\sim} 0.8\;{\rm \mu}\textrm{m}\;{\textrm{s}^{ - 1}}$) of the diffusive transport of the streptokinase thrombolytic drug through the blocked blood vessels (Clements Reference Clements2016). Of course, the comparison should be carried out for equivalent channel sizes and fluid viscosities. In the absence of experimental data, the extrapolation can be done using our theoretical model. To achieve the same recirculation speed of approximately $q = 5$ μm s⁻¹ in a blood vessel of $h \times b = 1 \times 1$ mm size by changing the aqueous solvent (${\eta _0}\sim {10^{ - 3}}\;\textrm{Pa}\;\textrm{s}$) with human blood (${\eta _0}\sim 7 \times {10^{ - 2}}\;\textrm{Pa}\;\textrm{s}$), one needs to apply the magnetic field amplitude of approximately ${H_0}\sim 50\;\textrm{kA}\;{\textrm{m}^{ - 1}}$ at the same other conditions ($\,f = 5\;\textrm{Hz}$, ${\varphi _p} = 1.6 \times {10^{ - 3}}$, $L \approx 30\;{\rm \mu}\textrm{m}$, $D \approx 6\;{\rm \mu}\textrm{m}$). The characteristic time ${\tau _p}$ of the drug ‘pumping’ within the occluded vessel is ${\tau _p}\sim {L_c}/q$, where ${L_c}\sim 200{-}500\;{\rm \mu}\textrm{m}$ is a typical distance between the inlet of the occluded vessel and the clot (inferred from MRI images in Nishimura et al. Reference Nishimura, Rosidi, Iadecola and Schaffer2010, Nguyen et al. Reference Nguyen, Nishimura, Fetcho, Iadecola and Schaffer2011). Taking the ‘worst’ value ${L_c} = 500\; \;{\rm \mu}\textrm{m}$, we obtain the maximum time of the convective delivery of molecules towards the clot of the order of ${\tau _p}\sim 100\;\textrm{s}$. However, the characteristic delivery time of the thrombolytic drug (tissue plasminogen activator, t-PA) by pure diffusion from the inlet of the occluded vessel towards the clot is ${\tau _D}\sim L_c^2/{\mathfrak{D}_{t\textrm{ - }PA}}\sim 1\;\textrm{h}$, where ${\mathfrak{D}_{t\textrm{ - }PA}}\sim 6.7 \times {10^{ - 11}}\;{\textrm{m}^2}\;{\textrm{s}^{ - 1}}$ is the t-PA diffusivity. This time appears to be higher than the convective delivery time ${\tau _p}$. This (possibly very optimistic) estimation provides some physical grounds for explanation for accelerated lysis of a blood clot formed in a rabbit jugular vein using magnetic colloids and rotating magnetic fields (Creighton Reference Creighton2012). More detailed studies coupling the hydrodynamic effects with the blood clot dissolution are required for a deeper understanding of this technique.

Another important feature of the observed recirculation flow is the duration of the steady-state regime, which starts at the time $t\mathrm{\ \mathbin{\lower.3ex\hbox{$\buildrel> \over {\smash{\scriptstyle\sim}\vphantom{_x}}$}}\ }100\;\textrm{s}$ from the moment of magnetic field application. As already mentioned, because of the heating of electromagnets, we were unable to determine with confidence the end of the steady state in experiments; we know that it lasts for at least 200 s from its beginning. However, we observe that the aggregates migrating to the left along the back wall do not come back but are accumulated near the left end of the channel once they arrive there. We can anticipate the end of the recirculation flow when most of the aggregates have moved to the left end, which is expected to occur at the time scale $\tau \sim l/{v_a} \sim 130\;\textrm{s}$, where ${v_a} \approx 15\;{\rm \mu}\textrm{m}\;{\textrm{s}^{ - 1}}$ is the average aggregate speed evaluated by image processing. This time is likely underestimated because ‘disappearance’ of some aggregates can be in principle compensated by the appearance and growth of other aggregates. Indeed, previous study has revealed that, in the absence of macroscopic flows, the aggregates stop to appear when the supersaturation is still high, and this is likely because of strong repulsive dipolar interactions between them significantly hampering the evolution to the equilibrium state (Ezzaier et al. Reference Ezzaier, Alves Marins, Razvin, Abbas, Ben Haj Amara, Zubarev and Kuzhir2017). Removing the aggregates from the major part of the channel can likely trigger a new nucleation/aggregate growth sequence. However, longer experimental observations with careful observation of the aggregate fate are required to confirm or reject this hypothesis.